Feynman diagram

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https://en.wikipedia.org/wiki/Feynman_diagram

In this Feynman diagram, an electron (e⁻) and a positron (e⁺) annihilate, producing a photon (γ, represented by the blue sine wave) that becomes a quark–antiquark pair (quark q, antiquark q̄), after which the antiquark radiates a gluon (g, represented by the green helix).

Feynman diagram

In theoretical physics, a Feynman diagram is a pictorial representation of the mathematical expressions describing the behavior and interaction of subatomic particles. The scheme is named after American physicist Richard Feynman, who introduced the diagrams in 1948.

The calculation of probability amplitudes in theoretical particle physics requires the use of large, complicated integrals over a large number of variables. Feynman diagrams instead represent these integrals graphically.

Feynman diagrams give a simple visualization of what would otherwise be an arcane and abstract formula. According to David Kaiser, "Since the middle of the 20th century, theoretical physicists have increasingly turned to this tool to help them undertake critical calculations. Feynman diagrams have revolutionized nearly every aspect of theoretical physics."

While the diagrams apply primarily to quantum field theory, they can be used in other areas of physics, such as solid-state theory. Frank Wilczek wrote that the calculations that won him the 2004 Nobel Prize in Physics "would have been literally unthinkable without Feynman diagrams, as would [Wilczek's] calculations that established a route to production and observation of the Higgs particle."

A Feynman diagram is a graphical representation of a perturbative contribution to the transition amplitude or correlation function of a quantum mechanical or statistical field theory. Within the canonical formulation of quantum field theory, a Feynman diagram represents a term in the Wick's expansion of the perturbative S-matrix. Alternatively, the path integral formulation of quantum field theory represents the transition amplitude as a weighted sum of all possible histories of the system from the initial to the final state, in terms of either particles or fields. The transition amplitude is then given as the matrix element of the S-matrix between the initial and final states of the quantum system.

Feynman used Ernst Stueckelberg's interpretation of the positron as if it were an electron moving backward in time. Thus, antiparticles are represented as moving backward along the time axis in Feynman diagrams.

Motivation and history

Richard Feynman in 1984

In this diagram, a kaon, made of an up quark and strange antiquark, decays both weakly and strongly into three pions, with intermediate steps involving a W boson and a gluon, represented by the blue sine wave and green spiral, respectively.

When calculating scattering cross-sections in particle physics, the interaction between particles can be described by starting from a free field that describes the incoming and outgoing particles, and including an interaction Hamiltonian to describe how the particles deflect one another. The amplitude for scattering is the sum of each possible interaction history over all possible intermediate particle states. The number of times the interaction Hamiltonian acts is the order of the perturbation expansion, and the time-dependent perturbation theory for fields is known as the Dyson series. When the intermediate states at intermediate times are energy eigenstates (collections of particles with a definite momentum) the series is called old-fashioned perturbation theory (or time-dependent/time-ordered perturbation theory).

The Dyson series can be alternatively rewritten as a sum over Feynman diagrams, where at each vertex both the energy and momentum are conserved, but where the length of the energy-momentum four-vector is not necessarily equal to the mass, i.e. the intermediate particles are so-called off-shell. The Feynman diagrams are much easier to keep track of than "old-fashioned" terms, because the old-fashioned way treats the particle and antiparticle contributions as separate. Each Feynman diagram is the sum of exponentially many old-fashioned terms, because each internal line can separately represent either a particle or an antiparticle. In a non-relativistic theory, there are no antiparticles and there is no doubling, so each Feynman diagram includes only one term.

Feynman gave a prescription for calculating the amplitude (the Feynman rules, below) for any given diagram from a field theory Lagrangian. Each internal line corresponds to a factor of the virtual particle's propagator; each vertex where lines meet gives a factor derived from an interaction term in the Lagrangian, and incoming and outgoing lines carry an energy, momentum, and spin.

In addition to their value as a mathematical tool, Feynman diagrams provide deep physical insight into the nature of particle interactions. Particles interact in every way available; in fact, intermediate virtual particles are allowed to propagate faster than light. The probability of each final state is then obtained by summing over all such possibilities. This is closely tied to the functional integral formulation of quantum mechanics, also invented by Feynman—see path integral formulation.

The naïve application of such calculations often produces diagrams whose amplitudes are infinite, because the short-distance particle interactions require a careful limiting procedure, to include particle self-interactions. The technique of renormalization, suggested by Ernst Stueckelberg and Hans Bethe and implemented by Dyson, Feynman, Schwinger, and Tomonaga compensates for this effect and eliminates the troublesome infinities. After renormalization, calculations using Feynman diagrams match experimental results with very high accuracy.

Feynman diagram and path integral methods are also used in statistical mechanics and can even be applied to classical mechanics.

Alternate names

Murray Gell-Mann always referred to Feynman diagrams as Stueckelberg diagrams, after Swiss physicist Ernst Stueckelberg, who devised a similar notation many years earlier. Stueckelberg was motivated by the need for a manifestly covariant formalism for quantum field theory, but did not provide as automated a way to handle symmetry factors and loops, although he was first to find the correct physical interpretation in terms of forward and backward in time particle paths, all without the path-integral.

Historically, as a book-keeping device of covariant perturbation theory, the graphs were called Feynman–Dyson diagrams or Dyson graphs, because the path integral was unfamiliar when they were introduced, and Freeman Dyson's derivation from old-fashioned perturbation theory borrowed from the perturbative expansions in statistical mechanics was easier to follow for physicists trained in earlier methods. Feynman had to lobby hard for the diagrams, which confused physicists trained in equations and graphs.

Representation of physical reality

In their presentations of fundamental interactions, written from the particle physics perspective, Gerard 't Hooft and Martinus Veltman gave good arguments for taking the original, non-regularized Feynman diagrams as the most succinct representation of the physics of quantum scattering of fundamental particles. Their motivations are consistent with the convictions of James Daniel Bjorken and Sidney Drell:

The Feynman graphs and rules of calculation summarize quantum field theory in a form in close contact with the experimental numbers one wants to understand. Although the statement of the theory in terms of graphs may imply perturbation theory, use of graphical methods in the many-body problem shows that this formalism is flexible enough to deal with phenomena of nonperturbative characters ... Some modification of the Feynman rules of calculation may well outlive the elaborate mathematical structure of local canonical quantum field theory ...

In quantum field theories, Feynman diagrams are obtained from a Lagrangian by Feynman rules.

Dimensional regularization is a method for regularizing integrals in the evaluation of Feynman diagrams; it assigns values to them that are meromorphic functions of an auxiliary complex parameter d, called the dimension. Dimensional regularization writes a Feynman integral as an integral depending on the spacetime dimension d and spacetime points.

Particle-path interpretation

A Feynman diagram is a representation of quantum field theory processes in terms of particle interactions. The particles are represented by the diagram lines. The lines can be squiggly or straight, with an arrow or without, depending on the type of particle. A point where lines connect to other lines is a vertex, and this is where the particles meet and interact. The interactions are: emit/absorb particles, deflect particles, or change particle type.

The three different types of lines are: internal lines, connecting vertices; incoming lines, extending from "the past" to a vertex, representing an initial state; and outgoing lines, extending from a vertex to "the future", representing the end state (the latter two are also known as external lines). Traditionally, the bottom of the diagram is the past and the top the future; alternatively, the past is to the left and the future to the right. When calculating correlation functions instead of scattering amplitudes, past and future are not relevant and all lines are internal. The particles then begin and end on small x's, which represent the positions of the operators whose correlation is calculated.

Feynman diagrams are a pictorial representation of a contribution to the total amplitude for a process that can happen in different ways. When a group of incoming particles scatter off each other, the process can be thought of as one where the particles travel over all possible paths, including paths that go backward in time.

Feynman diagrams are graphs that represent the interaction of particles rather than the physical position of the particle during a scattering process. They are not the same as spacetime diagrams and bubble chamber images even though they all describe particle scattering. Unlike a bubble chamber picture, only the sum of all relevant Feynman diagrams represent any given particle interaction; particles do not choose a particular diagram each time they interact. The law of summation is in accord with the principle of superposition—every diagram contributes to the total process's amplitude.

Description

General features of the scattering process A + B → C + D:
• internal lines (red) for intermediate particles and processes, which has a propagator factor ("prop"), external lines (orange) for incoming/outgoing particles to/from vertices (black),
• at each vertex there is 4-momentum conservation using delta functions, 4-momenta entering the vertex are positive while those leaving are negative, the factors at each vertex and internal line are multiplied in the amplitude integral,
• space x and time t axes are not always shown, directions of external lines correspond to passage of time.

A Feynman diagram represents a perturbative contribution to the amplitude of a quantum transition from some initial quantum state to some final quantum state.

For example, in the process of electron-positron annihilation the initial state is one electron and one positron, while the final state is two photons.

Conventionally, the initial state is at the left of the diagram and the final state at the right (although other layouts are also used).

The particles in the initial state are depicted by lines pointing in the direction of the initial state (e.g., to the left). The particles in the final state are represented by lines pointing in the direction of the final state (e.g., to the right).

QED involves two types of particles: matter particles such as electrons or positrons (called fermions) and exchange particles (called gauge bosons). They are represented in Feynman diagrams as follows:

• Electron in the initial state is represented by a solid line, with an arrow indicating the spin of the particle e.g. pointing toward the vertex (→•).
• Electron in the final state is represented by a line, with an arrow indicating the spin of the particle e.g. pointing away from the vertex: (•→).
• Positron in the initial state is represented by a solid line, with an arrow indicating the spin of the particle e.g. pointing away from the vertex: (←•).
• Positron in the final state is represented by a line, with an arrow indicating the spin of the particle e.g. pointing toward the vertex: (•←).
• Virtual Photon in the initial and the final states is represented by a wavy line (~• and •~).

In QED each vertex has three lines attached to it: one bosonic line, one fermionic line with arrow toward the vertex, and one fermionic line with arrow away from the vertex.

Vertices can be connected by a bosonic or fermionic propagator. A bosonic propagator is represented by a wavy line connecting two vertices (•~•). A fermionic propagator is represented by a solid line with an arrow connecting two vertices, (•←•).

The number of vertices gives the order of the term in the perturbation series expansion of the transition amplitude.

Electron–positron annihilation example

Feynman diagram of electron/positron annihilation

The electron–positron annihilation interaction:

e⁺ + e⁻ → 2γ

has a contribution from the second order Feynman diagram:

In the initial state (at the bottom; early time) there is one electron (e⁻) and one positron (e⁺) and in the final state (at the top; late time) there are two photons (γ).

Canonical quantization formulation

The probability amplitude for a transition of a quantum system (between asymptotically free states) from the initial state |i⟩ to the final state | f ⟩ is given by the matrix element

${\displaystyle S_{\mathrm{f}\mathrm{i}}=⟨\mathrm{f}|S|\mathrm{i}⟩,}$

where S is the S-matrix. In terms of the time-evolution operator U, it is simply

${\displaystyle S=\underset{t_2\to +\mathrm{\infty }}{lim}\left[\underset{t_1\to -\mathrm{\infty }}{lim}U(t_2,t_1)\right].}$

In the interaction picture, this expands to

${\displaystyle S=\mathcal{T}\exp \left(-i{\int }_{-\mathrm{\infty }}^{+\mathrm{\infty }}d\tau H_V(\tau )\right).}$

where H_V is the interaction Hamiltonian and ${\displaystyle \mathcal{T}}$ signifies the time-ordered product of operators. Dyson's formula expands the time-ordered matrix exponential into a perturbation series in the powers of the interaction Hamiltonian density,

${\displaystyle S=\sum _{n=0}^{\mathrm{\infty }}\frac{(-i{)}^n}{n!}\left(\prod _{j=1}^n\int d^4{x}_j\right)\mathcal{T}\left\{\prod _{j=1}^n{\mathcal{H}}_V\left(x_j\right)\right\}\equiv \sum _{n=0}^{\mathrm{\infty }}{S}^{(n)}.}$

Equivalently, with the interaction Lagrangian L_V, it is

${\displaystyle S=\sum _{n=0}^{\mathrm{\infty }}\frac{i^n}{n!}\left(\prod _{j=1}^n\int d^4{x}_j\right)\mathcal{T}\left\{\prod _{j=1}^n{\mathcal{L}}_V\left(x_j\right)\right\}\equiv \sum _{n=0}^{\mathrm{\infty }}{S}^{(n)}.}$

A Feynman diagram is a graphical representation of a single summand in the Wick's expansion of the time-ordered product in the nth-order term S⁽ⁿ⁾ of the Dyson series of the S-matrix,

${\displaystyle \mathcal{T}\prod _{j=1}^n{\mathcal{L}}_V\left(x_j\right)=\sum _{\text{A}}(\pm )\mathcal{N}\prod _{j=1}^n{\mathcal{L}}_V\left(x_j\right),}$

where N signifies the normal-ordered product of the operators and (±) takes care of the possible sign change when commuting the fermionic operators to bring them together for a contraction (a propagator) and A represents all possible contractions.

Feynman rules

The diagrams are drawn according to the Feynman rules, which depend upon the interaction Lagrangian. For the QED interaction Lagrangian

${\displaystyle L_v=-g\overline{\psi }{\gamma }^{\mu }\psi A_{\mu }}$

describing the interaction of a fermionic field ψ with a bosonic gauge field A_μ, the Feynman rules can be formulated in coordinate space as follows:

• Each integration coordinate x_j is represented by a point (sometimes called a vertex);
• A bosonic propagator is represented by a wiggly line connecting two points;
• A fermionic propagator is represented by a solid line connecting two points;
• A bosonic field ${\displaystyle A_{\mu }(x_i)}$ is represented by a wiggly line attached to the point x_i;
• A fermionic field ψ(x_i) is represented by a solid line attached to the point x_i with an arrow toward the point;
• An anti-fermionic field ψ(x_i) is represented by a solid line attached to the point x_i with an arrow away from the point;

Example: second order processes in QED

The second order perturbation term in the S-matrix is

${\displaystyle S^{(2)}=\frac{(ie{)}^2}{2!}\int d^{4}x{d}^4{x}^{\prime }T\overline{\psi }(x){\gamma }^{\mu }\psi (x)A_{\mu }(x)\overline{\psi }(x^{\prime }){\gamma }^{\nu }\psi (x^{\prime })A_{\nu }(x^{\prime }).}$

Scattering of fermions

The Feynman diagram of the term ${\displaystyle N\overline{\psi }(x)ie{\gamma }^{\mu }\psi (x)\overline{\psi }(x^{\prime })ie{\gamma }^{\nu }\psi (x^{\prime })A_{\mu }(x)A_{\nu }(x^{\prime })}$

The Wick's expansion of the integrand gives (among others) the following term

${\displaystyle N\overline{\psi }(x){\gamma }^{\mu }\psi (x)\overline{\psi }(x^{\prime }){\gamma }^{\nu }\psi (x^{\prime })\underset{\_}{A_{\mu }(x)A_{\nu }(x^{\prime })},}$

where

${\displaystyle \underset{\_}{A_{\mu }(x)A_{\nu }(x^{\prime })}=\int \frac{d^{4}k}{(2\pi {)}^4}\frac{-i{g}_{\mu \nu }}{k^2+i0}{e}^{-ik(x-x^{\prime })}}$

is the electromagnetic contraction (propagator) in the Feynman gauge. This term is represented by the Feynman diagram at the right. This diagram gives contributions to the following processes:

1. e⁻ e⁻ scattering (initial state at the right, final state at the left of the diagram);
2. e⁺ e⁺ scattering (initial state at the left, final state at the right of the diagram);
3. e⁻ e⁺ scattering (initial state at the bottom/top, final state at the top/bottom of the diagram).

Compton scattering and annihilation/generation of e⁻ e⁺ pairs

Another interesting term in the expansion is

${\displaystyle N\overline{\psi }(x){\gamma }^{\mu }\underset{\_}{\psi (x)\overline{\psi }(x^{\prime })}{\gamma }^{\nu }\psi (x^{\prime })A_{\mu }(x)A_{\nu }(x^{\prime }),}$

where

${\displaystyle \underset{\_}{\psi (x)\overline{\psi }(x^{\prime })}=\int \frac{d^{4}p}{(2\pi {)}^4}\frac{i}{\gamma p-m+i0}{e}^{-ip(x-x^{\prime })}}$

is the fermionic contraction (propagator).

Path integral formulation

In a path integral, the field Lagrangian, integrated over all possible field histories, defines the probability amplitude to go from one field configuration to another. In order to make sense, the field theory must have a well-defined ground state, and the integral must be performed a little bit rotated into imaginary time, i.e. a Wick rotation. The path integral formalism is completely equivalent to the canonical operator formalism above.

Scalar field Lagrangian

A simple example is the free relativistic scalar field in d dimensions, whose action integral is:

${\displaystyle S=\int \frac{1}{2}{\mathrm{\partial }}_{\mu }\varphi {\mathrm{\partial }}^{\mu }\varphi d^{d}x.}$

The probability amplitude for a process is:

${\displaystyle {\int }_A^B{e}^{iS}D\varphi ,}$

where A and B are space-like hypersurfaces that define the boundary conditions. The collection of all the φ(A) on the starting hypersurface give the field's initial value, analogous to the starting position for a point particle, and the field values φ(B) at each point of the final hypersurface defines the final field value, which is allowed to vary, giving a different amplitude to end up at different values. This is the field-to-field transition amplitude.

The path integral gives the expectation value of operators between the initial and final state:

${\displaystyle {\int }_A^B{e}^{iS}\varphi (x_1)\cdots \varphi (x_n)D\varphi =⟨A\left|\varphi (x_1)\cdots \varphi (x_n)\right|B⟩,}$

and in the limit that A and B recede to the infinite past and the infinite future, the only contribution that matters is from the ground state (this is only rigorously true if the path-integral is defined slightly rotated into imaginary time). The path integral can be thought of as analogous to a probability distribution, and it is convenient to define it so that multiplying by a constant does not change anything:

${\displaystyle \frac{{\displaystyle \int e^{iS}\varphi (x_1)\cdots \varphi (x_n)D\varphi }}{{\displaystyle \int e^{iS}D\varphi }}=⟨0\left|\varphi (x_1)\cdots \varphi (x_n)\right|0⟩.}$

The field's partition function is the normalization factor on the bottom, which coincides with the statistical mechanical partition function at zero temperature when rotated into imaginary time.

The initial-to-final amplitudes are ill-defined if one thinks of the continuum limit right from the beginning, because the fluctuations in the field can become unbounded. So the path-integral can be thought of as on a discrete square lattice, with lattice spacing a and the limit a → 0 should be taken carefully. If the final results do not depend on the shape of the lattice or the value of a, then the continuum limit exists.

On a lattice

On a lattice, (i), the field can be expanded in Fourier modes:

${\displaystyle \varphi (x)=\int \frac{dk}{(2\pi {)}^d}\varphi (k)e^{ik\cdot x}={\int }_k\varphi (k)e^{ikx}.}$

Here the integration domain is over k restricted to a cube of side length ⁠2π/a⁠, so that large values of k are not allowed. It is important to note that the k-measure contains the factors of 2π from Fourier transforms, this is the best standard convention for k-integrals in QFT. The lattice means that fluctuations at large k are not allowed to contribute right away, they only start to contribute in the limit a → 0. Sometimes, instead of a lattice, the field modes are just cut off at high values of k instead.

It is also convenient from time to time to consider the space-time volume to be finite, so that the k modes are also a lattice. This is not strictly as necessary as the space-lattice limit, because interactions in k are not localized, but it is convenient for keeping track of the factors in front of the k-integrals and the momentum-conserving delta functions that will arise.

On a lattice, (ii), the action needs to be discretized:

${\displaystyle S=\sum _{⟨x,y⟩}\frac{1}{2}(\varphi (x)-\varphi (y){)}^2,}$

where ⟨x,y⟩ is a pair of nearest lattice neighbors x and y. The discretization should be thought of as defining what the derivative ∂_μφ means.

In terms of the lattice Fourier modes, the action can be written:

${\displaystyle S={\int }_k((1-\cos (k_1))+(1-\cos (k_2))+\cdots +(1-\cos (k_d))){\varphi }_k^{\ast }{\varphi }^k.}$

For k near zero this is:

${\displaystyle S={\int }_k\frac{1}{2}{k}^2{\left|\varphi (k)\right|}^2.}$

Now we have the continuum Fourier transform of the original action. In finite volume, the quantity d^dk is not infinitesimal, but becomes the volume of a box made by neighboring Fourier modes, or (⁠2π/V⁠)^d
 .

The field φ is real-valued, so the Fourier transform obeys:

${\displaystyle \varphi (k{)}^{\ast }=\varphi (-k).}$

In terms of real and imaginary parts, the real part of φ(k) is an even function of k, while the imaginary part is odd. The Fourier transform avoids double-counting, so that it can be written:

${\displaystyle S={\int }_k\frac{1}{2}{k}^2\varphi (k)\varphi (-k)}$

over an integration domain that integrates over each pair (k,−k) exactly once.

For a complex scalar field with action

${\displaystyle S=\int \frac{1}{2}{\mathrm{\partial }}_{\mu }{\varphi }^{\ast }{\mathrm{\partial }}^{\mu }\varphi d^{d}x}$

the Fourier transform is unconstrained:

${\displaystyle S={\int }_k\frac{1}{2}{k}^2{\left|\varphi (k)\right|}^2}$

and the integral is over all k.

Integrating over all different values of φ(x) is equivalent to integrating over all Fourier modes, because taking a Fourier transform is a unitary linear transformation of field coordinates. When you change coordinates in a multidimensional integral by a linear transformation, the value of the new integral is given by the determinant of the transformation matrix. If

${\displaystyle y_i=A_{ij}{x}_j,}$

then

${\displaystyle det(A)\int d{x}_{1}d{x}_2\cdots d{x}_n=\int d{y}_{1}d{y}_2\cdots d{y}_n.}$

If A is a rotation, then

${\displaystyle A^{\mathrm{T}}A=I}$

so that det A = ±1, and the sign depends on whether the rotation includes a reflection or not.

The matrix that changes coordinates from φ(x) to φ(k) can be read off from the definition of a Fourier transform.

${\displaystyle A_{kx}=e^{ikx}}$

and the Fourier inversion theorem tells you the inverse:

${\displaystyle A_{kx}^{-1}=e^{-ikx}}$

which is the complex conjugate-transpose, up to factors of 2π. On a finite volume lattice, the determinant is nonzero and independent of the field values.

${\displaystyle detA=1}$

and the path integral is a separate factor at each value of k.

${\displaystyle \int \exp \left(\frac{i}{2}\sum _k{k}^2{\varphi }^{\ast }(k)\varphi (k)\right)D\varphi =\prod _k{\int }_{{\varphi }_k}{e}^{\frac{i}{2}{k}^2{\left|{\varphi }_k\right|}^2{d}^{d}k}}$

The factor d^dk is the infinitesimal volume of a discrete cell in k-space, in a square lattice box

${\displaystyle d^{d}k={\left(\frac{1}{L}\right)}^d,}$

where L is the side-length of the box. Each separate factor is an oscillatory Gaussian, and the width of the Gaussian diverges as the volume goes to infinity.

In imaginary time, the Euclidean action becomes positive definite, and can be interpreted as a probability distribution. The probability of a field having values φ_k is

${\displaystyle e^{{\int }_k-\frac{1}{2}{k}^2{\varphi }_k^{\ast }{\varphi }_k}=\prod _k{e}^{-k^2{\left|{\varphi }_k\right|}^2{d}^{d}k}.}$

The expectation value of the field is the statistical expectation value of the field when chosen according to the probability distribution:

${\displaystyle ⟨\varphi (x_1)\cdots \varphi (x_n)⟩=\frac{{\displaystyle \int e^{-S}\varphi (x_1)\cdots \varphi (x_n)D\varphi }}{{\displaystyle \int e^{-S}D\varphi }}}$

Since the probability of φ_k is a product, the value of φ_k at each separate value of k is independently Gaussian distributed. The variance of the Gaussian is ⁠1/k²d^dk⁠, which is formally infinite, but that just means that the fluctuations are unbounded in infinite volume. In any finite volume, the integral is replaced by a discrete sum, and the variance of the integral is ⁠V/k²⁠.

Monte Carlo

The path integral defines a probabilistic algorithm to generate a Euclidean scalar field configuration. Randomly pick the real and imaginary parts of each Fourier mode at wavenumber k to be a Gaussian random variable with variance ⁠1/k²⁠. This generates a configuration φ_C(k) at random, and the Fourier transform gives φ_C(x). For real scalar fields, the algorithm must generate only one of each pair φ(k), φ(−k), and make the second the complex conjugate of the first.

To find any correlation function, generate a field again and again by this procedure, and find the statistical average:

${\displaystyle ⟨\varphi (x_1)\cdots \varphi (x_n)⟩=\underset{|C|\to \mathrm{\infty }}{lim}\frac{{\displaystyle \sum _C{\varphi }_C(x_1)\cdots {\varphi }_C(x_n)}}{|C|}}$

where |C| is the number of configurations, and the sum is of the product of the field values on each configuration. The Euclidean correlation function is just the same as the correlation function in statistics or statistical mechanics. The quantum mechanical correlation functions are an analytic continuation of the Euclidean correlation functions.

For free fields with a quadratic action, the probability distribution is a high-dimensional Gaussian, and the statistical average is given by an explicit formula. But the Monte Carlo method also works well for bosonic interacting field theories where there is no closed form for the correlation functions.

Scalar propagator

Each mode is independently Gaussian distributed. The expectation of field modes is easy to calculate:

${\displaystyle ⟨{\varphi }_k{\varphi }_{k^{\prime }}⟩=0}$

for k ≠ k′, since then the two Gaussian random variables are independent and both have zero mean.

${\displaystyle ⟨{\varphi }_k{\varphi }_k⟩=\frac{V}{k^2}}$

in finite volume V, when the two k-values coincide, since this is the variance of the Gaussian. In the infinite volume limit,

${\displaystyle ⟨\varphi (k)\varphi (k^{\prime })⟩=\delta (k-k^{\prime })\frac{1}{k^2}}$

Strictly speaking, this is an approximation: the lattice propagator is:

${\displaystyle ⟨\varphi (k)\varphi (k^{\prime })⟩=\delta (k-k^{\prime })\frac{1}{2(d-\cos (k_1)+\cos (k_2)\cdots +\cos (k_d))}}$

But near k = 0, for field fluctuations long compared to the lattice spacing, the two forms coincide.

The delta functions contain factors of 2π, so that they cancel out the 2π factors in the measure for k integrals.

${\displaystyle \delta (k)=(2\pi {)}^d{\delta }_D(k_1){\delta }_D(k_2)\cdots {\delta }_D(k_d)}$

where δ_D(k) is the ordinary one-dimensional Dirac delta function. This convention for delta-functions is not universal—some authors keep the factors of 2π in the delta functions (and in the k-integration) explicit.

Equation of motion

The form of the propagator can be more easily found by using the equation of motion for the field. From the Lagrangian, the equation of motion is:

${\displaystyle {\mathrm{\partial }}_{\mu }{\mathrm{\partial }}^{\mu }\varphi =0}$

and in an expectation value, this says:

${\displaystyle {\mathrm{\partial }}_{\mu }{\mathrm{\partial }}^{\mu }⟨\varphi (x)\varphi (y)⟩=0}$

Where the derivatives act on x, and the identity is true everywhere except when x and y coincide, and the operator order matters. The form of the singularity can be understood from the canonical commutation relations to be a delta-function. Defining the (Euclidean) Feynman propagator Δ as the Fourier transform of the time-ordered two-point function (the one that comes from the path-integral):

${\displaystyle {\mathrm{\partial }}^2\mathrm{\Delta }(x)=i\delta (x)}$

So that:

${\displaystyle \mathrm{\Delta }(k)=\frac{i}{k^2}}$

If the equations of motion are linear, the propagator will always be the reciprocal of the quadratic-form matrix that defines the free Lagrangian, since this gives the equations of motion. This is also easy to see directly from the path integral. The factor of i disappears in the Euclidean theory.

Wick theorem

Main article: Wick's theorem

Because each field mode is an independent Gaussian, the expectation values for the product of many field modes obeys Wick's theorem:

${\displaystyle ⟨\varphi (k_1)\varphi (k_2)\cdots \varphi (k_n)⟩}$

is zero unless the field modes coincide in pairs. This means that it is zero for an odd number of φ, and for an even number of φ, it is equal to a contribution from each pair separately, with a delta function.

${\displaystyle ⟨\varphi (k_1)\cdots \varphi (k_{2n})⟩=\sum \prod _{i,j}\frac{\delta \left(k_i-k_j\right)}{k_i^2}}$

where the sum is over each partition of the field modes into pairs, and the product is over the pairs. For example,

${\displaystyle ⟨\varphi (k_1)\varphi (k_2)\varphi (k_3)\varphi (k_4)⟩=\frac{\delta (k_1-k_2)}{k_1^2}\frac{\delta (k_3-k_4)}{k_3^2}+\frac{\delta (k_1-k_3)}{k_3^2}\frac{\delta (k_2-k_4)}{k_2^2}+\frac{\delta (k_1-k_4)}{k_1^2}\frac{\delta (k_2-k_3)}{k_2^2}}$

An interpretation of Wick's theorem is that each field insertion can be thought of as a dangling line, and the expectation value is calculated by linking up the lines in pairs, putting a delta function factor that ensures that the momentum of each partner in the pair is equal, and dividing by the propagator.

Higher Gaussian moments — completing Wick's theorem

There is a subtle point left before Wick's theorem is proved—what if more than two of the ${\displaystyle \varphi }$s have the same momentum? If it's an odd number, the integral is zero; negative values cancel with the positive values. But if the number is even, the integral is positive. The previous demonstration assumed that the ${\displaystyle \varphi }$s would only match up in pairs.

But the theorem is correct even when arbitrarily many of the ${\displaystyle \varphi }$ are equal, and this is a notable property of Gaussian integration:

${\displaystyle I=\int e^{-a{x}^2/2}dx=\sqrt{\frac{2\pi }{a}}}$

${\displaystyle \frac{{\mathrm{\partial }}^n}{\mathrm{\partial }{a}^n}I=\int \frac{x^{2n}}{2^n}{e}^{-a{x}^2/2}dx=\frac{1\cdot 3\cdot 5\dots \cdot (2n-1)}{2\cdot 2\cdot 2\dots \cdot 2}\sqrt{2\pi }{a}^{-\frac{2n+1}{2}}}$

Dividing by I,

${\displaystyle ⟨x^{2n}⟩=\frac{{\displaystyle \int x^{2n}{e}^{-a{x}^2/2}}}{{\displaystyle \int e^{-a{x}^2/2}}}=1\cdot 3\cdot 5\dots \cdot (2n-1)\frac{1}{a^n}}$

${\displaystyle ⟨x^2⟩=\frac{1}{a}}$

If Wick's theorem were correct, the higher moments would be given by all possible pairings of a list of 2n different x:

${\displaystyle ⟨x_1{x}_2{x}_3\cdots x_{2n}⟩}$

where the x are all the same variable, the index is just to keep track of the number of ways to pair them. The first x can be paired with 2n − 1 others, leaving 2n − 2. The next unpaired x can be paired with 2n − 3 different x leaving 2n − 4, and so on. This means that Wick's theorem, uncorrected, says that the expectation value of x²ⁿ should be:

${\displaystyle ⟨x^{2n}⟩=(2n-1)\cdot (2n-3)\dots \cdot 5\cdot 3\cdot 1{⟨x^2⟩}^n}$

and this is in fact the correct answer. So Wick's theorem holds no matter how many of the momenta of the internal variables coincide.

Interaction

Interactions are represented by higher order contributions, since quadratic contributions are always Gaussian. The simplest interaction is the quartic self-interaction, with an action:

${\displaystyle S=\int {\mathrm{\partial }}^{\mu }\varphi {\mathrm{\partial }}_{\mu }\varphi +\frac{\lambda }{4!}{\varphi }^4.}$

The reason for the combinatorial factor 4! will be clear soon. Writing the action in terms of the lattice (or continuum) Fourier modes:

${\displaystyle S={\int }_k{k}^2{\left|\varphi (k)\right|}^2+\frac{\lambda }{4!}{\int }_{k_1{k}_2{k}_3{k}_4}\varphi (k_1)\varphi (k_2)\varphi (k_3)\varphi (k_4)\delta (k_1+k_2+k_3+k_4)=S_F+X.}$

Where S_F is the free action, whose correlation functions are given by Wick's theorem. The exponential of S in the path integral can be expanded in powers of λ, giving a series of corrections to the free action.

${\displaystyle e^{-S}=e^{-S_F}\left(1+X+\frac{1}{2!}XX+\frac{1}{3!}XXX+\cdots \right)}$

The path integral for the interacting action is then a power series of corrections to the free action. The term represented by X should be thought of as four half-lines, one for each factor of φ(k). The half-lines meet at a vertex, which contributes a delta-function that ensures that the sum of the momenta are all equal.

To compute a correlation function in the interacting theory, there is a contribution from the X terms now. For example, the path-integral for the four-field correlator:

${\displaystyle ⟨\varphi (k_1)\varphi (k_2)\varphi (k_3)\varphi (k_4)⟩=\frac{{\displaystyle \int e^{-S}\varphi (k_1)\varphi (k_2)\varphi (k_3)\varphi (k_4)D\varphi }}{Z}}$

which in the free field was only nonzero when the momenta k were equal in pairs, is now nonzero for all values of k. The momenta of the insertions φ(k_i) can now match up with the momenta of the Xs in the expansion. The insertions should also be thought of as half-lines, four in this case, which carry a momentum k, but one that is not integrated.

The lowest-order contribution comes from the first nontrivial term e^−S_FX in the Taylor expansion of the action. Wick's theorem requires that the momenta in the X half-lines, the φ(k) factors in X, should match up with the momenta of the external half-lines in pairs. The new contribution is equal to:

${\displaystyle \lambda \frac{1}{k_1^2}\frac{1}{k_2^2}\frac{1}{k_3^2}\frac{1}{k_4^2}.}$

The 4! inside X is canceled because there are exactly 4! ways to match the half-lines in X to the external half-lines. Each of these different ways of matching the half-lines together in pairs contributes exactly once, regardless of the values of k_1,2,3,4, by Wick's theorem.

Feynman diagrams

The expansion of the action in powers of X gives a series of terms with progressively higher number of Xs. The contribution from the term with exactly n Xs is called nth order.

The nth order terms has:

1. 4n internal half-lines, which are the factors of φ(k) from the Xs. These all end on a vertex, and are integrated over all possible k.
2. external half-lines, which are the come from the φ(k) insertions in the integral.

By Wick's theorem, each pair of half-lines must be paired together to make a line, and this line gives a factor of

${\displaystyle \frac{\delta (k_1+k_2)}{k_1^2}}$

which multiplies the contribution. This means that the two half-lines that make a line are forced to have equal and opposite momentum. The line itself should be labelled by an arrow, drawn parallel to the line, and labeled by the momentum in the line k. The half-line at the tail end of the arrow carries momentum k, while the half-line at the head-end carries momentum −k. If one of the two half-lines is external, this kills the integral over the internal k, since it forces the internal k to be equal to the external k. If both are internal, the integral over k remains.

The diagrams that are formed by linking the half-lines in the Xs with the external half-lines, representing insertions, are the Feynman diagrams of this theory. Each line carries a factor of ⁠1/k²⁠, the propagator, and either goes from vertex to vertex, or ends at an insertion. If it is internal, it is integrated over. At each vertex, the total incoming k is equal to the total outgoing k.

The number of ways of making a diagram by joining half-lines into lines almost completely cancels the factorial factors coming from the Taylor series of the exponential and the 4! at each vertex.

Loop order

A forest diagram is one where all the internal lines have momentum that is completely determined by the external lines and the condition that the incoming and outgoing momentum are equal at each vertex. The contribution of these diagrams is a product of propagators, without any integration. A tree diagram is a connected forest diagram.

An example of a tree diagram is the one where each of four external lines end on an X. Another is when three external lines end on an X, and the remaining half-line joins up with another X, and the remaining half-lines of this X run off to external lines. These are all also forest diagrams (as every tree is a forest); an example of a forest that is not a tree is when eight external lines end on two Xs.

It is easy to verify that in all these cases, the momenta on all the internal lines is determined by the external momenta and the condition of momentum conservation in each vertex.

A diagram that is not a forest diagram is called a loop diagram, and an example is one where two lines of an X are joined to external lines, while the remaining two lines are joined to each other. The two lines joined to each other can have any momentum at all, since they both enter and leave the same vertex. A more complicated example is one where two Xs are joined to each other by matching the legs one to the other. This diagram has no external lines at all.

The reason loop diagrams are called loop diagrams is because the number of k-integrals that are left undetermined by momentum conservation is equal to the number of independent closed loops in the diagram, where independent loops are counted as in homology theory. The homology is real-valued (actually R^d valued), the value associated with each line is the momentum. The boundary operator takes each line to the sum of the end-vertices with a positive sign at the head and a negative sign at the tail. The condition that the momentum is conserved is exactly the condition that the boundary of the k-valued weighted graph is zero.

A set of valid k-values can be arbitrarily redefined whenever there is a closed loop. A closed loop is a cyclical path of adjacent vertices that never revisits the same vertex. Such a cycle can be thought of as the boundary of a hypothetical 2-cell. The k-labellings of a graph that conserve momentum (i.e. which has zero boundary) up to redefinitions of k (i.e. up to boundaries of 2-cells) define the first homology of a graph. The number of independent momenta that are not determined is then equal to the number of independent homology loops. For many graphs, this is equal to the number of loops as counted in the most intuitive way.

Symmetry factors

The number of ways to form a given Feynman diagram by joining half-lines is large, and by Wick's theorem, each way of pairing up the half-lines contributes equally. Often, this completely cancels the factorials in the denominator of each term, but the cancellation is sometimes incomplete.

The uncancelled denominator is called the symmetry factor of the diagram. The contribution of each diagram to the correlation function must be divided by its symmetry factor.

For example, consider the Feynman diagram formed from two external lines joined to one X, and the remaining two half-lines in the X joined to each other. There are 4 × 3 ways to join the external half-lines to the X, and then there is only one way to join the two remaining lines to each other. The X comes divided by 4! = 4 × 3 × 2, but the number of ways to link up the X half lines to make the diagram is only 4 × 3, so the contribution of this diagram is divided by two.

For another example, consider the diagram formed by joining all the half-lines of one X to all the half-lines of another X. This diagram is called a vacuum bubble, because it does not link up to any external lines. There are 4! ways to form this diagram, but the denominator includes a 2! (from the expansion of the exponential, there are two Xs) and two factors of 4!. The contribution is multiplied by ⁠4!/2 × 4! × 4!⁠ = ⁠1/48⁠.

Another example is the Feynman diagram formed from two Xs where each X links up to two external lines, and the remaining two half-lines of each X are joined to each other. The number of ways to link an X to two external lines is 4 × 3, and either X could link up to either pair, giving an additional factor of 2. The remaining two half-lines in the two Xs can be linked to each other in two ways, so that the total number of ways to form the diagram is 4 × 3 × 4 × 3 × 2 × 2, while the denominator is 4! × 4! × 2!. The total symmetry factor is 2, and the contribution of this diagram is divided by 2.

The symmetry factor theorem gives the symmetry factor for a general diagram: the contribution of each Feynman diagram must be divided by the order of its group of automorphisms, the number of symmetries that it has.

An automorphism of a Feynman graph is a permutation M of the lines and a permutation N of the vertices with the following properties:

1. If a line l goes from vertex v to vertex v′, then M(l) goes from N(v) to N(v′). If the line is undirected, as it is for a real scalar field, then M(l) can go from N(v′) to N(v) too.
2. If a line l ends on an external line, M(l) ends on the same external line.
3. If there are different types of lines, M(l) should preserve the type.

This theorem has an interpretation in terms of particle-paths: when identical particles are present, the integral over all intermediate particles must not double-count states that differ only by interchanging identical particles.

Proof: To prove this theorem, label all the internal and external lines of a diagram with a unique name. Then form the diagram by linking a half-line to a name and then to the other half line.

Now count the number of ways to form the named diagram. Each permutation of the Xs gives a different pattern of linking names to half-lines, and this is a factor of n!. Each permutation of the half-lines in a single X gives a factor of 4!. So a named diagram can be formed in exactly as many ways as the denominator of the Feynman expansion.

But the number of unnamed diagrams is smaller than the number of named diagram by the order of the automorphism group of the graph.

Connected diagrams: linked-cluster theorem

Roughly speaking, a Feynman diagram is called connected if all vertices and propagator lines are linked by a sequence of vertices and propagators of the diagram itself. If one views it as an undirected graph it is connected. The remarkable relevance of such diagrams in QFTs is due to the fact that they are sufficient to determine the quantum partition function Z[J]. More precisely, connected Feynman diagrams determine

${\displaystyle iW[J]\equiv \ln Z[J].}$

To see this, one should recall that

${\displaystyle Z[J]\propto \sum _k{D}_k}$

with D_k constructed from some (arbitrary) Feynman diagram that can be thought to consist of several connected components C_i. If one encounters n_i (identical) copies of a component C_i within the Feynman diagram D_k one has to include a symmetry factor n_i!. However, in the end each contribution of a Feynman diagram D_k to the partition function has the generic form

${\displaystyle \prod _i\frac{C_i^{n_i}}{n_i!}}$

where i labels the (infinitely) many connected Feynman diagrams possible.

A scheme to successively create such contributions from the D_k to Z[J] is obtained by

${\displaystyle \left(\frac{1}{0!}+\frac{C_1}{1!}+\frac{C_1^2}{2!}+\cdots \right)\left(1+C_2+\frac{1}{2}{C}_2^2+\cdots \right)\cdots }$

and therefore yields

${\displaystyle Z[J]\propto \prod _i\sum _{n_i=0}^{\mathrm{\infty }}\frac{C_i^{n_i}}{n_i!}=\exp \sum _i{C}_i\propto \exp W[J].}$

To establish the normalization Z₀ = exp W[0] = 1 one simply calculates all connected vacuum diagrams, i.e., the diagrams without any sources J (sometimes referred to as external legs of a Feynman diagram).

The linked-cluster theorem was first proved to order four by Keith Brueckner in 1955, and for infinite orders by Jeffrey Goldstone in 1957.

Vacuum bubbles

An immediate consequence of the linked-cluster theorem is that all vacuum bubbles, diagrams without external lines, cancel when calculating correlation functions. A correlation function is given by a ratio of path-integrals:

${\displaystyle ⟨{\varphi }_1(x_1)\cdots {\varphi }_n(x_n)⟩=\frac{{\displaystyle \int e^{-S}{\varphi }_1(x_1)\cdots {\varphi }_n(x_n)D\varphi }}{{\displaystyle \int e^{-S}D\varphi }}.}$

The top is the sum over all Feynman diagrams, including disconnected diagrams that do not link up to external lines at all. In terms of the connected diagrams, the numerator includes the same contributions of vacuum bubbles as the denominator:

${\displaystyle \int e^{-S}{\varphi }_1(x_1)\cdots {\varphi }_n(x_n)D\varphi =\left(\sum E_i\right)\left(\exp \left(\sum _i{C}_i\right)\right).}$

Where the sum over E diagrams includes only those diagrams each of whose connected components end on at least one external line. The vacuum bubbles are the same whatever the external lines, and give an overall multiplicative factor. The denominator is the sum over all vacuum bubbles, and dividing gets rid of the second factor.

The vacuum bubbles then are only useful for determining Z itself, which from the definition of the path integral is equal to:

${\displaystyle Z=\int e^{-S}D\varphi =e^{-HT}=e^{-\rho V}}$

where ρ is the energy density in the vacuum. Each vacuum bubble contains a factor of δ(k) zeroing the total k at each vertex, and when there are no external lines, this contains a factor of δ(0), because the momentum conservation is over-enforced. In finite volume, this factor can be identified as the total volume of space time. Dividing by the volume, the remaining integral for the vacuum bubble has an interpretation: it is a contribution to the energy density of the vacuum.

Sources

Correlation functions are the sum of the connected Feynman diagrams, but the formalism treats the connected and disconnected diagrams differently. Internal lines end on vertices, while external lines go off to insertions. Introducing sources unifies the formalism, by making new vertices where one line can end.

Sources are external fields, fields that contribute to the action, but are not dynamical variables. A scalar field source is another scalar field h that contributes a term to the (Lorentz) Lagrangian:

${\displaystyle \int h(x)\varphi (x)d^{d}x=\int h(k)\varphi (k)d^{d}k}$

In the Feynman expansion, this contributes H terms with one half-line ending on a vertex. Lines in a Feynman diagram can now end either on an X vertex, or on an H vertex, and only one line enters an H vertex. The Feynman rule for an H vertex is that a line from an H with momentum k gets a factor of h(k).

The sum of the connected diagrams in the presence of sources includes a term for each connected diagram in the absence of sources, except now the diagrams can end on the source. Traditionally, a source is represented by a little "×" with one line extending out, exactly as an insertion.

${\displaystyle \log (Z[h])=\sum _{n,C}h(k_1)h(k_2)\cdots h(k_n)C(k_1,\cdots ,k_n)}$

where C(k₁,...,k_n) is the connected diagram with n external lines carrying momentum as indicated. The sum is over all connected diagrams, as before.

The field h is not dynamical, which means that there is no path integral over h: h is just a parameter in the Lagrangian, which varies from point to point. The path integral for the field is:

${\displaystyle Z[h]=\int e^{iS+i\int h\varphi }D\varphi }$

and it is a function of the values of h at every point. One way to interpret this expression is that it is taking the Fourier transform in field space. If there is a probability density on Rⁿ, the Fourier transform of the probability density is:

${\displaystyle \int \rho (y)e^{iky}{d}^{n}y=⟨e^{iky}⟩=⟨\prod _{i=1}^n{e}^{i{h}_i{y}_i}⟩}$

The Fourier transform is the expectation of an oscillatory exponential. The path integral in the presence of a source h(x) is:

${\displaystyle Z[h]=\int e^{iS}{e}^{i{\int }_{x}h(x)\varphi (x)}D\varphi =⟨e^{ih\varphi }⟩}$

which, on a lattice, is the product of an oscillatory exponential for each field value:

${\displaystyle ⟨\prod _x{e}^{i{h}_x{\varphi }_x}⟩}$

The Fourier transform of a delta-function is a constant, which gives a formal expression for a delta function:

${\displaystyle \delta (x-y)=\int e^{ik(x-y)}dk}$

This tells you what a field delta function looks like in a path-integral. For two scalar fields φ and η,

${\displaystyle \delta (\varphi -\eta )=\int e^{ih(x)(\varphi (x)-\eta (x))d^{d}x}Dh,}$

which integrates over the Fourier transform coordinate, over h. This expression is useful for formally changing field coordinates in the path integral, much as a delta function is used to change coordinates in an ordinary multi-dimensional integral.

The partition function is now a function of the field h, and the physical partition function is the value when h is the zero function:

The correlation functions are derivatives of the path integral with respect to the source:

${\displaystyle ⟨\varphi (x)⟩=\frac{1}{Z}\frac{\mathrm{\partial }}{\mathrm{\partial }h(x)}Z[h]=\frac{\mathrm{\partial }}{\mathrm{\partial }h(x)}\log (Z[h]).}$

In Euclidean space, source contributions to the action can still appear with a factor of i, so that they still do a Fourier transform.

Spin ⁠1/2⁠; "photons" and "ghosts"

Spin ⁠1/2⁠: Grassmann integrals

The field path integral can be extended to the Fermi case, but only if the notion of integration is expanded. A Grassmann integral of a free Fermi field is a high-dimensional determinant or Pfaffian, which defines the new type of Gaussian integration appropriate for Fermi fields.

The two fundamental formulas of Grassmann integration are:

${\displaystyle \int e^{M_{ij}{\overline{\psi }}^i{\psi }^j}D\overline{\psi }D\psi =\mathrm{D}\mathrm{e}\mathrm{t}(M),}$

where M is an arbitrary matrix and ψ, ψ are independent Grassmann variables for each index i, and

${\displaystyle \int e^{\frac{1}{2}{A}_{ij}{\psi }^i{\psi }^j}D\psi =\mathrm{P}\mathrm{f}\mathrm{a}\mathrm{f}\mathrm{f}(A),}$

where A is an antisymmetric matrix, ψ is a collection of Grassmann variables, and the ⁠1/2⁠ is to prevent double-counting (since ψⁱψ^j = −ψ^jψⁱ).

In matrix notation, where ψ and η are Grassmann-valued row vectors, η and ψ are Grassmann-valued column vectors, and M is a real-valued matrix:

${\displaystyle Z=\int e^{\overline{\psi }M\psi +\overline{\eta }\psi +\overline{\psi }\eta }D\overline{\psi }D\psi =\int e^{\left(\overline{\psi }+\overline{\eta }{M}^{-1}\right)M\left(\psi +M^{-1}\eta \right)-\overline{\eta }{M}^{-1}\eta }D\overline{\psi }D\psi =\mathrm{D}\mathrm{e}\mathrm{t}(M)e^{-\overline{\eta }{M}^{-1}\eta },}$

where the last equality is a consequence of the translation invariance of the Grassmann integral. The Grassmann variables η are external sources for ψ, and differentiating with respect to η pulls down factors of ψ.

${\displaystyle ⟨\overline{\psi }\psi ⟩=\frac{1}{Z}\frac{\mathrm{\partial }}{\mathrm{\partial }\eta }\frac{\mathrm{\partial }}{\mathrm{\partial }\overline{\eta }}Z{|}_{\eta =\overline{\eta }=0}=M^{-1}}$

again, in a schematic matrix notation. The meaning of the formula above is that the derivative with respect to the appropriate component of η and η gives the matrix element of M⁻¹. This is exactly analogous to the bosonic path integration formula for a Gaussian integral of a complex bosonic field:

${\displaystyle \int e^{{\varphi }^{\ast }M\varphi +h^{\ast }\varphi +{\varphi }^{\ast }h}D{\varphi }^{\ast }D\varphi =\frac{e^{h^{\ast }{M}^{-1}h}}{\mathrm{D}\mathrm{e}\mathrm{t}(M)}}$

${\displaystyle ⟨{\varphi }^{\ast }\varphi ⟩=\frac{1}{Z}\frac{\mathrm{\partial }}{\mathrm{\partial }h}\frac{\mathrm{\partial }}{\mathrm{\partial }{h}^{\ast }}Z{|}_{h=h^{\ast }=0}=M^{-1}.}$

So that the propagator is the inverse of the matrix in the quadratic part of the action in both the Bose and Fermi case.

For real Grassmann fields, for Majorana fermions, the path integral is a Pfaffian times a source quadratic form, and the formulas give the square root of the determinant, just as they do for real Bosonic fields. The propagator is still the inverse of the quadratic part.

The free Dirac Lagrangian:

${\displaystyle \int \overline{\psi }\left({\gamma }^{\mu }{\mathrm{\partial }}_{\mu }-m\right)\psi }$

formally gives the equations of motion and the anticommutation relations of the Dirac field, just as the Klein Gordon Lagrangian in an ordinary path integral gives the equations of motion and commutation relations of the scalar field. By using the spatial Fourier transform of the Dirac field as a new basis for the Grassmann algebra, the quadratic part of the Dirac action becomes simple to invert:

${\displaystyle S={\int }_k\overline{\psi }\left(i{\gamma }^{\mu }{k}_{\mu }-m\right)\psi .}$

The propagator is the inverse of the matrix M linking ψ(k) and ψ(k), since different values of k do not mix together.

${\displaystyle ⟨\overline{\psi }(k^{\prime })\psi (k)⟩=\delta (k+k^{\prime })\frac{1}{\gamma \cdot k-m}=\delta (k+k^{\prime })\frac{\gamma \cdot k+m}{k^2-m^2}}$

The analog of Wick's theorem matches ψ and ψ in pairs:

${\displaystyle ⟨\overline{\psi }(k_1)\overline{\psi }(k_2)\cdots \overline{\psi }(k_n)\psi (k_1^{\prime })\cdots \psi (k_n)⟩=\sum _{\mathrm{p}\mathrm{a}\mathrm{i}\mathrm{r}\mathrm{i}\mathrm{n}\mathrm{g}\mathrm{s}}(-1{)}^S\prod _{\mathrm{p}\mathrm{a}\mathrm{i}\mathrm{r}\mathrm{s}i,j}\delta \left(k_i-k_j\right)\frac{1}{\gamma \cdot k_i-m}}$

where S is the sign of the permutation that reorders the sequence of ψ and ψ to put the ones that are paired up to make the delta-functions next to each other, with the ψ coming right before the ψ. Since a ψ, ψ pair is a commuting element of the Grassmann algebra, it does not matter what order the pairs are in. If more than one ψ, ψ pair have the same k, the integral is zero, and it is easy to check that the sum over pairings gives zero in this case (there are always an even number of them). This is the Grassmann analog of the higher Gaussian moments that completed the Bosonic Wick's theorem earlier.

The rules for spin-⁠1/2⁠ Dirac particles are as follows: The propagator is the inverse of the Dirac operator, the lines have arrows just as for a complex scalar field, and the diagram acquires an overall factor of −1 for each closed Fermi loop. If there are an odd number of Fermi loops, the diagram changes sign. Historically, the −1 rule was very difficult for Feynman to discover. He discovered it after a long process of trial and error, since he lacked a proper theory of Grassmann integration.

The rule follows from the observation that the number of Fermi lines at a vertex is always even. Each term in the Lagrangian must always be Bosonic. A Fermi loop is counted by following Fermionic lines until one comes back to the starting point, then removing those lines from the diagram. Repeating this process eventually erases all the Fermionic lines: this is the Euler algorithm to 2-color a graph, which works whenever each vertex has even degree. The number of steps in the Euler algorithm is only equal to the number of independent Fermionic homology cycles in the common special case that all terms in the Lagrangian are exactly quadratic in the Fermi fields, so that each vertex has exactly two Fermionic lines. When there are four-Fermi interactions (like in the Fermi effective theory of the weak nuclear interactions) there are more k-integrals than Fermi loops. In this case, the counting rule should apply the Euler algorithm by pairing up the Fermi lines at each vertex into pairs that together form a bosonic factor of the term in the Lagrangian, and when entering a vertex by one line, the algorithm should always leave with the partner line.

To clarify and prove the rule, consider a Feynman diagram formed from vertices, terms in the Lagrangian, with Fermion fields. The full term is Bosonic, it is a commuting element of the Grassmann algebra, so the order in which the vertices appear is not important. The Fermi lines are linked into loops, and when traversing the loop, one can reorder the vertex terms one after the other as one goes around without any sign cost. The exception is when you return to the starting point, and the final half-line must be joined with the unlinked first half-line. This requires one permutation to move the last ψ to go in front of the first ψ, and this gives the sign.

This rule is the only visible effect of the exclusion principle in internal lines. When there are external lines, the amplitudes are antisymmetric when two Fermi insertions for identical particles are interchanged. This is automatic in the source formalism, because the sources for Fermi fields are themselves Grassmann valued.

Spin 1: photons

The naive propagator for photons is infinite, since the Lagrangian for the A-field is:

${\displaystyle S=\int \frac{1}{4}{F}^{\mu \nu }{F}_{\mu \nu }=\int -\frac{1}{2}\left({\mathrm{\partial }}^{\mu }{A}_{\nu }{\mathrm{\partial }}_{\mu }{A}^{\nu }-{\mathrm{\partial }}^{\mu }{A}_{\mu }{\mathrm{\partial }}_{\nu }{A}^{\nu }\right).}$

The quadratic form defining the propagator is non-invertible. The reason is the gauge invariance of the field; adding a gradient to A does not change the physics.

To fix this problem, one needs to fix a gauge. The most convenient way is to demand that the divergence of A is some function f, whose value is random from point to point. It does no harm to integrate over the values of f, since it only determines the choice of gauge. This procedure inserts the following factor into the path integral for A:

${\displaystyle \int \delta \left({\mathrm{\partial }}_{\mu }{A}^{\mu }-f\right)e^{-\frac{f^2}{2}}Df.}$

The first factor, the delta function, fixes the gauge. The second factor sums over different values of f that are inequivalent gauge fixings. This is simply

${\displaystyle e^{-\frac{{\left({\mathrm{\partial }}_{\mu }{A}_{\mu }\right)}^2}{2}}.}$

The additional contribution from gauge-fixing cancels the second half of the free Lagrangian, giving the Feynman Lagrangian:

${\displaystyle S=\int {\mathrm{\partial }}^{\mu }{A}^{\nu }{\mathrm{\partial }}_{\mu }{A}_{\nu }}$

which is just like four independent free scalar fields, one for each component of A. The Feynman propagator is:

${\displaystyle ⟨A_{\mu }(k)A_{\nu }(k^{\prime })⟩=\delta \left(k+k^{\prime }\right)\frac{g_{\mu \nu }}{k^2}.}$

The one difference is that the sign of one propagator is wrong in the Lorentz case: the timelike component has an opposite sign propagator. This means that these particle states have negative norm—they are not physical states. In the case of photons, it is easy to show by diagram methods that these states are not physical—their contribution cancels with longitudinal photons to only leave two physical photon polarization contributions for any value of k.

If the averaging over f is done with a coefficient different from ⁠1/2⁠, the two terms do not cancel completely. This gives a covariant Lagrangian with a coefficient ${\displaystyle \lambda }$, which does not affect anything:

${\displaystyle S=\int \frac{1}{2}\left({\mathrm{\partial }}^{\mu }{A}^{\nu }{\mathrm{\partial }}_{\mu }{A}_{\nu }-\lambda {\left({\mathrm{\partial }}_{\mu }{A}^{\mu }\right)}^2\right)}$

and the covariant propagator for QED is:

${\displaystyle ⟨A_{\mu }(k)A_{\nu }(k^{\prime })⟩=\delta \left(k+k^{\prime }\right)\frac{g_{\mu \nu }-\lambda \frac{k_{\mu }{k}_{\nu }}{k^2}}{k^2}.}$

Spin 1: non-Abelian ghosts

To find the Feynman rules for non-Abelian gauge fields, the procedure that performs the gauge fixing must be carefully corrected to account for a change of variables in the path-integral.

The gauge fixing factor has an extra determinant from popping the delta function:

${\displaystyle \delta \left({\mathrm{\partial }}_{\mu }{A}_{\mu }-f\right)e^{-\frac{f^2}{2}}detM}$

To find the form of the determinant, consider first a simple two-dimensional integral of a function f that depends only on r, not on the angle θ. Inserting an integral over θ:

${\displaystyle \int f(r)dxdy=\int f(r)\int d\theta \delta (y)\left|\frac{dy}{d\theta }\right|dxdy}$

The derivative-factor ensures that popping the delta function in θ removes the integral. Exchanging the order of integration,

${\displaystyle \int f(r)dxdy=\int d\theta \int f(r)\delta (y)\left|\frac{dy}{d\theta }\right|dxdy}$

but now the delta-function can be popped in y,

${\displaystyle \int f(r)dxdy=\int d{\theta }_0\int f(x)\left|\frac{dy}{d\theta }\right|dx.}$

The integral over θ just gives an overall factor of 2π, while the rate of change of y with a change in θ is just x, so this exercise reproduces the standard formula for polar integration of a radial function:

${\displaystyle \int f(r)dxdy=2\pi \int f(x)xdx}$

In the path-integral for a nonabelian gauge field, the analogous manipulation is:

${\displaystyle \int DA\int \delta (F(A))det\left(\frac{\mathrm{\partial }F}{\mathrm{\partial }G}\right)DG{e}^{iS}=\int DG\int \delta (F(A))det\left(\frac{\mathrm{\partial }F}{\mathrm{\partial }G}\right)e^{iS}}$

The factor in front is the volume of the gauge group, and it contributes a constant, which can be discarded. The remaining integral is over the gauge fixed action.

${\displaystyle \int det\left(\frac{\mathrm{\partial }F}{\mathrm{\partial }G}\right)e^{i{S}_{GF}}DA}$

To get a covariant gauge, the gauge fixing condition is the same as in the Abelian case:

${\displaystyle {\mathrm{\partial }}_{\mu }{A}^{\mu }=f,}$

Whose variation under an infinitesimal gauge transformation is given by:

${\displaystyle {\mathrm{\partial }}_{\mu }{D}_{\mu }\alpha ,}$

where α is the adjoint valued element of the Lie algebra at every point that performs the infinitesimal gauge transformation. This adds the Faddeev Popov determinant to the action:

${\displaystyle det\left({\mathrm{\partial }}_{\mu }{D}_{\mu }\right)}$

which can be rewritten as a Grassmann integral by introducing ghost fields:

${\displaystyle \int e^{\overline{\eta }{\mathrm{\partial }}_{\mu }{D}^{\mu }\eta }D\overline{\eta }D\eta }$

The determinant is independent of f, so the path-integral over f can give the Feynman propagator (or a covariant propagator) by choosing the measure for f as in the abelian case. The full gauge fixed action is then the Yang Mills action in Feynman gauge with an additional ghost action:

${\displaystyle S=\int \mathrm{Tr}{\mathrm{\partial }}_{\mu }{A}_{\nu }{\mathrm{\partial }}^{\mu }{A}^{\nu }+f_{jk}^i{\mathrm{\partial }}^{\nu }{A}_i^{\mu }{A}_{\mu }^j{A}_{\nu }^k+f_{jr}^i{f}_{kl}^r{A}_i{A}_j{A}^k{A}^l+\mathrm{Tr}{\mathrm{\partial }}_{\mu }\overline{\eta }{\mathrm{\partial }}^{\mu }\eta +\overline{\eta }{A}_j\eta }$

The diagrams are derived from this action. The propagator for the spin-1 fields has the usual Feynman form. There are vertices of degree 3 with momentum factors whose couplings are the structure constants, and vertices of degree 4 whose couplings are products of structure constants. There are additional ghost loops, which cancel out timelike and longitudinal states in A loops.

In the Abelian case, the determinant for covariant gauges does not depend on A, so the ghosts do not contribute to the connected diagrams.

Particle-path representation

Feynman diagrams were originally discovered by Feynman, by trial and error, as a way to represent the contribution to the S-matrix from different classes of particle trajectories.

Schwinger representation

The Euclidean scalar propagator has a suggestive representation:

${\displaystyle \frac{1}{p^2+m^2}={\int }_0^{\mathrm{\infty }}{e}^{-\tau \left(p^2+m^2\right)}d\tau }$

The meaning of this identity (which is an elementary integration) is made clearer by Fourier transforming to real space.

${\displaystyle \mathrm{\Delta }(x)={\int }_0^{\mathrm{\infty }}d\tau e^{-m^2\tau }\frac{1}{(4\pi \tau {)}^{d/2}}{e}^{\frac{-x^2}{4\tau }}}$

The contribution at any one value of τ to the propagator is a Gaussian of width ̀̀√τ. The total propagation function from 0 to x is a weighted sum over all proper times τ of a normalized Gaussian, the probability of ending up at x after a random walk of time τ.

The path-integral representation for the propagator is then:

${\displaystyle \mathrm{\Delta }(x)={\int }_0^{\mathrm{\infty }}d\tau \int DX{e}^{-\underset{0}{\overset{\tau }{\int }}\left(\frac{{\dot{x}}^2}{2}+m^2\right)d{\tau }^{\prime }}}$

which is a path-integral rewrite of the Schwinger representation.

The Schwinger representation is both useful for making manifest the particle aspect of the propagator, and for symmetrizing denominators of loop diagrams.

Combining denominators

The Schwinger representation has an immediate practical application to loop diagrams. For example, for the diagram in the φ⁴ theory formed by joining two xs together in two half-lines, and making the remaining lines external, the integral over the internal propagators in the loop is:

${\displaystyle {\int }_k\frac{1}{k^2+m^2}\frac{1}{(k+p{)}^2+m^2}.}$

Here one line carries momentum k and the other k + p. The asymmetry can be fixed by putting everything in the Schwinger representation.

${\displaystyle {\int }_{t,t^{\prime }}{e}^{-t(k^2+m^2)-t^{\prime }\left((k+p{)}^2+m^2\right)}dtd{t}^{\prime }.}$

Now the exponent mostly depends on t + t′,

${\displaystyle {\int }_{t,t^{\prime }}{e}^{-(t+t^{\prime })(k^2+m^2)-t^{\prime }2p\cdot k-t^{\prime }{p}^2},}$

except for the asymmetrical little bit. Defining the variable u = t + t′ and v = ⁠t′/u⁠, the variable u goes from 0 to ∞, while v goes from 0 to 1. The variable u is the total proper time for the loop, while v parametrizes the fraction of the proper time on the top of the loop versus the bottom.

The Jacobian for this transformation of variables is easy to work out from the identities:

${\displaystyle d(uv)=d{t}^{\prime }du=dt+d{t}^{\prime },}$

and "wedging" gives

${\displaystyle udu\wedge dv=dt\wedge d{t}^{\prime }}$.

This allows the u integral to be evaluated explicitly:

${\displaystyle {\int }_{u,v}u{e}^{-u\left(k^2+m^2+v2p\cdot k+v{p}^2\right)}=\int \frac{1}{{\left(k^2+m^2+v2p\cdot k-v{p}^2\right)}^2}dv}$

leaving only the v-integral. This method, invented by Schwinger but usually attributed to Feynman, is called combining denominator. Abstractly, it is the elementary identity:

${\displaystyle \frac{1}{AB}={\int }_0^1\frac{1}{(vA+(1-v)B{)}^2}dv}$

But this form does not provide the physical motivation for introducing v; v is the proportion of proper time on one of the legs of the loop.

Once the denominators are combined, a shift in k to k′ = k + vp symmetrizes everything:

${\displaystyle {\int }_0^1\int \frac{1}{{\left(k^2+m^2+2vp\cdot k+v{p}^2\right)}^2}dkdv={\int }_0^1\int \frac{1}{{\left(k^{\prime 2}+m^2+v(1-v)p^2\right)}^2}d{k}^{\prime }dv}$

This form shows that the moment that p² is more negative than four times the mass of the particle in the loop, which happens in a physical region of Lorentz space, the integral has a cut. This is exactly when the external momentum can create physical particles.

When the loop has more vertices, there are more denominators to combine:

${\displaystyle \int dk\frac{1}{k^2+m^2}\frac{1}{(k+p_1{)}^2+m^2}\cdots \frac{1}{(k+p_n{)}^2+m^2}}$

The general rule follows from the Schwinger prescription for n + 1 denominators:

${\displaystyle \frac{1}{D_0{D}_1\cdots D_n}={\int }_0^{\mathrm{\infty }}\cdots {\int }_0^{\mathrm{\infty }}{e}^{-u_0{D}_0\cdots -u_n{D}_n}d{u}_0\cdots d{u}_n.}$

The integral over the Schwinger parameters u_i can be split up as before into an integral over the total proper time u = u₀ + u₁ ... + u_n and an integral over the fraction of the proper time in all but the first segment of the loop v_i = ⁠u_i/u⁠ for i ∈ {1,2,...,n}. The v_i are positive and add up to less than 1, so that the v integral is over an n-dimensional simplex.

The Jacobian for the coordinate transformation can be worked out as before:

${\displaystyle du=d{u}_0+d{u}_1\cdots +d{u}_n}$

${\displaystyle d(u{v}_i)=d{u}_i.}$

Wedging all these equations together, one obtains

${\displaystyle u^{n}du\wedge d{v}_1\wedge d{v}_2\cdots \wedge d{v}_n=d{u}_0\wedge d{u}_1\cdots \wedge d{u}_n.}$

This gives the integral:

${\displaystyle {\int }_0^{\mathrm{\infty }}{\int }_{\mathrm{s}\mathrm{i}\mathrm{m}\mathrm{p}\mathrm{l}\mathrm{e}\mathrm{x}}{u}^n{e}^{-u\left(v_0{D}_0+v_1{D}_1+v_2{D}_2\cdots +v_n{D}_n\right)}d{v}_1\cdots d{v}_{n}du,}$

where the simplex is the region defined by the conditions

${\displaystyle v_i>0\text{and}\sum _{i=1}^n{v}_i<1}$

as well as

${\displaystyle v_0=1-\sum _{i=1}^n{v}_i.}$

Performing the u integral gives the general prescription for combining denominators:

${\displaystyle \frac{1}{D_0\cdots D_n}=n!{\int }_{\mathrm{s}\mathrm{i}\mathrm{m}\mathrm{p}\mathrm{l}\mathrm{e}\mathrm{x}}\frac{1}{{\left(v_0{D}_0+v_1{D}_1\cdots +v_n{D}_n\right)}^{n+1}}d{v}_{1}d{v}_2\cdots d{v}_n}$

Since the numerator of the integrand is not involved, the same prescription works for any loop, no matter what the spins are carried by the legs. The interpretation of the parameters v_i is that they are the fraction of the total proper time spent on each leg.

Scattering

The correlation functions of a quantum field theory describe the scattering of particles. The definition of "particle" in relativistic field theory is not self-evident, because if you try to determine the position so that the uncertainty is less than the compton wavelength, the uncertainty in energy is large enough to produce more particles and antiparticles of the same type from the vacuum. This means that the notion of a single-particle state is to some extent incompatible with the notion of an object localized in space.

In the 1930s, Wigner gave a mathematical definition for single-particle states: they are a collection of states that form an irreducible representation of the Poincaré group. Single particle states describe an object with a finite mass, a well defined momentum, and a spin. This definition is fine for protons and neutrons, electrons and photons, but it excludes quarks, which are permanently confined, so the modern point of view is more accommodating: a particle is anything whose interaction can be described in terms of Feynman diagrams, which have an interpretation as a sum over particle trajectories.

A field operator can act to produce a one-particle state from the vacuum, which means that the field operator φ(x) produces a superposition of Wigner particle states. In the free field theory, the field produces one particle states only. But when there are interactions, the field operator can also produce 3-particle, 5-particle (if there is no +/− symmetry also 2, 4, 6 particle) states too. To compute the scattering amplitude for single particle states only requires a careful limit, sending the fields to infinity and integrating over space to get rid of the higher-order corrections.

The relation between scattering and correlation functions is the LSZ-theorem: The scattering amplitude for n particles to go to m particles in a scattering event is the given by the sum of the Feynman diagrams that go into the correlation function for n + m field insertions, leaving out the propagators for the external legs.

For example, for the λφ⁴ interaction of the previous section, the order λ contribution to the (Lorentz) correlation function is:

${\displaystyle ⟨\varphi (k_1)\varphi (k_2)\varphi (k_3)\varphi (k_4)⟩=\frac{i}{k_1^2}\frac{i}{k_2^2}\frac{i}{k_3^2}\frac{i}{k_4^2}i\lambda }$

Stripping off the external propagators, that is, removing the factors of ⁠i/k²⁠, gives the invariant scattering amplitude M:

${\displaystyle M=i\lambda }$

which is a constant, independent of the incoming and outgoing momentum. The interpretation of the scattering amplitude is that the sum of |M|² over all possible final states is the probability for the scattering event. The normalization of the single-particle states must be chosen carefully, however, to ensure that M is a relativistic invariant.

Non-relativistic single particle states are labeled by the momentum k, and they are chosen to have the same norm at every value of k. This is because the nonrelativistic unit operator on single particle states is:

${\displaystyle \int dk|k⟩⟨k|.}$

In relativity, the integral over the k-states for a particle of mass m integrates over a hyperbola in E,k space defined by the energy–momentum relation:

${\displaystyle E^2-k^2=m^2.}$

If the integral weighs each k point equally, the measure is not Lorentz-invariant. The invariant measure integrates over all values of k and E, restricting to the hyperbola with a Lorentz-invariant delta function:

${\displaystyle \int \delta (E^2-k^2-m^2)|E,k⟩⟨E,k|dEdk=\int \frac{dk}{2E}|k⟩⟨k|.}$

So the normalized k-states are different from the relativistically normalized k-states by a factor of

${\displaystyle \sqrt{E}={\left(k^2-m^2\right)}^{\frac{1}{4}}.}$

The invariant amplitude M is then the probability amplitude for relativistically normalized incoming states to become relativistically normalized outgoing states.

For nonrelativistic values of k, the relativistic normalization is the same as the nonrelativistic normalization (up to a constant factor ${\displaystyle \sqrt{m}}$). In this limit, the φ⁴ invariant scattering amplitude is still constant. The particles created by the field φ scatter in all directions with equal amplitude.

The nonrelativistic potential, which scatters in all directions with an equal amplitude (in the Born approximation), is one whose Fourier transform is constant—a delta-function potential. The lowest order scattering of the theory reveals the non-relativistic interpretation of this theory—it describes a collection of particles with a delta-function repulsion. Two such particles have an aversion to occupying the same point at the same time.

Nonperturbative effects

Thinking of Feynman diagrams as a perturbation series, nonperturbative effects like tunneling do not show up, because any effect that goes to zero faster than any polynomial does not affect the Taylor series. Even bound states are absent, since at any finite order particles are only exchanged a finite number of times, and to make a bound state, the binding force must last forever.

But this point of view is misleading, because the diagrams not only describe scattering, but they also are a representation of the short-distance field theory correlations. They encode not only asymptotic processes like particle scattering, they also describe the multiplication rules for fields, the operator product expansion. Nonperturbative tunneling processes involve field configurations that on average get big when the coupling constant gets small, but each configuration is a coherent superposition of particles whose local interactions are described by Feynman diagrams. When the coupling is small, these become collective processes that involve large numbers of particles, but where the interactions between each of the particles is simple. (The perturbation series of any interacting quantum field theory has zero radius of convergence, complicating the limit of the infinite series of diagrams needed (in the limit of vanishing coupling) to describe such field configurations.)

This means that nonperturbative effects show up asymptotically in resummations of infinite classes of diagrams, and these diagrams can be locally simple. The graphs determine the local equations of motion, while the allowed large-scale configurations describe non-perturbative physics. But because Feynman propagators are nonlocal in time, translating a field process to a coherent particle language is not completely intuitive, and has only been explicitly worked out in certain special cases. In the case of nonrelativistic bound states, the Bethe–Salpeter equation describes the class of diagrams to include to describe a relativistic atom. For quantum chromodynamics, the Shifman–Vainshtein–Zakharov sum rules describe non-perturbatively excited long-wavelength field modes in particle language, but only in a phenomenological way.

The number of Feynman diagrams at high orders of perturbation theory is very large, because there are as many diagrams as there are graphs with a given number of nodes. Nonperturbative effects leave a signature on the way in which the number of diagrams and resummations diverge at high order. It is only because non-perturbative effects appear in hidden form in diagrams that it was possible to analyze nonperturbative effects in string theory, where in many cases a Feynman description is the only one available.

In popular culture

The use of the above diagram of the virtual particle producing a quark–antiquark pair was featured in the television sit-com The Big Bang Theory, in the episode "The Bat Jar Conjecture".

PhD Comics of January 11, 2012, shows Feynman diagrams that visualize and describe quantum academic interactions, i.e. the paths followed by Ph.D. students when interacting with their advisors.

Vacuum Diagrams, a science fiction story by Stephen Baxter, features the titular vacuum diagram, a specific type of Feynman diagram.

Feynman and his wife, Gweneth Howarth, bought a Dodge Tradesman Maxivan in 1975, and had it painted with Feynman diagrams. The van is currently owned by video game designer and physicist Seamus Blackley. Qantum was the license plate ID.

Variational Monte Carlo

From Wikipedia, the free encyclopedia

https://en.wikipedia.org/wiki/Variational_Monte_Carlo 

In computational physics, variational Monte Carlo (VMC) is a quantum Monte Carlo method that applies the variational method to approximate the ground state of a quantum system.

The basic building block is a generic wave function ${\displaystyle |\mathrm{\Psi }(a)⟩}$ depending on some parameters ${\displaystyle a}$. The optimal values of the parameters ${\displaystyle a}$ is then found upon minimizing the total energy of the system.

In particular, given the Hamiltonian ${\displaystyle \mathcal{H}}$, and denoting with ${\displaystyle X}$ a many-body configuration, the expectation value of the energy can be written as:

$${\displaystyle E(a)=\frac{⟨\mathrm{\Psi }(a)|\mathcal{H}|\mathrm{\Psi }(a)⟩}{⟨\mathrm{\Psi }(a)|\mathrm{\Psi }(a)⟩}=\frac{\int |\mathrm{\Psi }(X,a){|}^2\frac{\mathcal{H}\mathrm{\Psi }(X,a)}{\mathrm{\Psi }(X,a)}dX}{\int |\mathrm{\Psi }(X,a){|}^{2}dX}.}$$

Following the Monte Carlo method for evaluating integrals, we can interpret ${\displaystyle \frac{|\mathrm{\Psi }(X,a){|}^2}{\int |\mathrm{\Psi }(X,a){|}^{2}dX}}$ as a probability distribution function, sample it, and evaluate the energy expectation value ${\displaystyle E(a)}$ as the average of the so-called local energy ${\displaystyle E_{\text{loc}}(X)=\frac{\mathcal{H}\mathrm{\Psi }(X,a)}{\mathrm{\Psi }(X,a)}}$. Once ${\displaystyle E(a)}$ is known for a given set of variational parameters ${\displaystyle a}$, then optimization is performed in order to minimize the energy and obtain the best possible representation of the ground-state wave-function.

VMC is no different from any other variational method, except that the many-dimensional integrals are evaluated numerically. Monte Carlo integration is particularly crucial in this problem since the dimension of the many-body Hilbert space, comprising all the possible values of the configurations ${\displaystyle X}$, typically grows exponentially with the size of the physical system. Other approaches to the numerical evaluation of the energy expectation values would therefore, in general, limit applications to much smaller systems than those analyzable thanks to the Monte Carlo approach.

The accuracy of the method then largely depends on the choice of the variational state. The simplest choice typically corresponds to a mean-field form, where the state ${\displaystyle \mathrm{\Psi }}$ is written as a factorization over the Hilbert space. This particularly simple form is typically not very accurate since it neglects many-body effects. One of the largest gains in accuracy over writing the wave function separably comes from the introduction of the so-called Jastrow factor. In this case the wave function is written as $\mathrm{\Psi }(X)=\exp (\sum u(r_{ij}))$, where ${\displaystyle r_{ij}}$ is the distance between a pair of quantum particles and ${\displaystyle u(r)}$ is a variational function to be determined. With this factor, we can explicitly account for particle-particle correlation, but the many-body integral becomes unseparable, so Monte Carlo is the only way to evaluate it efficiently. In chemical systems, slightly more sophisticated versions of this factor can obtain 80–90% of the correlation energy (see electronic correlation) with less than 30 parameters. In comparison, a configuration interaction calculation may require around 50,000 parameters to reach that accuracy, although it depends greatly on the particular case being considered. In addition, VMC usually scales as a small power of the number of particles in the simulation, usually something like N²⁻⁴ for calculation of the energy expectation value, depending on the form of the wave function.

Wave function optimization in VMC

QMC calculations crucially depend on the quality of the trial-function, and so it is essential to have an optimized wave-function as close as possible to the ground state. The problem of function optimization is a very important research topic in numerical simulation. In QMC, in addition to the usual difficulties to find the minimum of multidimensional parametric function, the statistical noise is present in the estimate of the cost function (usually the energy), and its derivatives, required for an efficient optimization.

Different cost functions and different strategies were used to optimize a many-body trial-function. Usually three cost functions were used in QMC optimization energy, variance or a linear combination of them. The variance optimization method has the advantage that the exact wavefunction's variance is known. (Because the exact wavefunction is an eigenfunction of the Hamiltonian, the variance of the local energy is zero). This means that variance optimization is ideal in that it is bounded from below, it is positive defined and its minimum is known. Energy minimization may ultimately prove more effective, however, as different authors recently showed that the energy optimization is more effective than the variance one.

There are different motivations for this: first, usually one is interested in the lowest energy rather than in the lowest variance in both variational and diffusion Monte Carlo; second, variance optimization takes many iterations to optimize determinant parameters and often the optimization can get stuck in multiple local minimum and it suffers of the "false convergence" problem; third energy-minimized wave functions on average yield more accurate values of other expectation values than variance minimized wave functions do.

The optimization strategies can be divided into three categories. The first strategy is based on correlated sampling together with deterministic optimization methods. Even if this idea yielded very accurate results for the first-row atoms, this procedure can have problems if parameters affect the nodes, and moreover density ratio of the current and initial trial-function increases exponentially with the size of the system. In the second strategy one use a large bin to evaluate the cost function and its derivatives in such way that the noise can be neglected and deterministic methods can be used.

The third approach, is based on an iterative technique to handle directly with noise functions. The first example of these methods is the so-called Stochastic Gradient Approximation (SGA), that was used also for structure optimization. Recently an improved and faster approach of this kind was proposed the so-called Stochastic Reconfiguration (SR) method.

VMC and deep learning

In 2017, Giuseppe Carleo and Matthias Troyer used a VMC objective function to train an artificial neural network to find the ground state of a quantum mechanical system. More generally, artificial neural networks are being used as a wave function ansatz (known as neural network quantum states) in VMC frameworks for finding ground states of quantum mechanical systems. The use of neural network ansatzes for VMC has been extended to fermions, enabling electronic structure calculations that are significantly more accurate than VMC calculations which do not use neural networks.

Alcubierre drive

From Wikipedia, the free encyclopedia

https://en.wikipedia.org/wiki/Alcubierre_drive 

Two-dimensional visualization of an Alcubierre drive, showing the opposing regions of expanding and contracting spacetime that displace the central region

The Alcubierre drive ([alkuˈβjere]) is a speculative warp drive idea according to which a spacecraft could achieve apparent faster-than-light travel by contracting space in front of it and expanding space behind it, under the assumption that a configurable energy-density field lower than that of vacuum (that is, negative mass) could be created. Proposed by theoretical physicist Miguel Alcubierre in 1994, the Alcubierre drive is based on a solution of Einstein's field equations. Since those solutions are metric tensors, the Alcubierre drive is also referred to as Alcubierre metric.

Objects cannot accelerate to the speed of light within normal spacetime; instead, the Alcubierre drive shifts space around an object so that the object would arrive at its destination more quickly than light would in normal space without breaking any physical laws.

Although the metric proposed by Alcubierre is consistent with the Einstein field equations, construction of such a drive is not necessarily possible. The proposed mechanism of the Alcubierre drive implies a negative energy density and therefore requires exotic matter or manipulation of dark energy. If exotic matter with the correct properties does not exist, then the drive cannot be constructed. At the close of his original article, however, Alcubierre argued (following an argument developed by physicists analyzing traversable wormholes) that the Casimir vacuum between parallel plates could fulfill the negative-energy requirement for the Alcubierre drive.

Another possible issue is that, although the Alcubierre metric is consistent with Einstein's equations, general relativity does not incorporate quantum mechanics. Some physicists have presented arguments to suggest that a theory of quantum gravity (which would incorporate both theories) would eliminate those solutions in general relativity that allow for backward time travel (see the chronology protection conjecture) and thus make the Alcubierre drive invalid.

History

In 1994, Miguel Alcubierre proposed a method for changing the geometry of space by creating a wave that would cause the fabric of space ahead of a spacecraft to contract and the space behind it to expand. The ship would then ride this wave inside a region of flat space, known as a warp bubble, and would not move within this bubble but instead be carried along as the region itself moves due to the actions of the drive. The local velocity relative to the deformed spacetime would be subluminal, but the speed at which a spacecraft could move would be superluminal, thereby rendering possible interstellar flight, such as a visit to Proxima Centauri within a few days.

Alcubierre metric

The Alcubierre metric defines the warp-drive spacetime. It is a Lorentzian manifold that, if interpreted in the context of general relativity, allows a warp bubble to appear in previously flat spacetime and move away at effectively faster-than-light speed. The interior of the bubble is an inertial reference frame and inhabitants experience no proper acceleration. This method of transport does not involve objects in motion at faster-than-light speeds with respect to the contents of the warp bubble; that is, a light beam within the warp bubble would still always move more quickly than the ship. Because objects within the bubble are not moving (locally) more quickly than light, the mathematical formulation of the Alcubierre metric is consistent with the conventional claims of the laws of relativity (namely, that an object with mass cannot attain or exceed the speed of light) and conventional relativistic effects such as time dilation would not apply as they would with conventional motion at near-light speeds.

An extension of the Alcubierre metric that eliminates the expansion of the volume elements and instead relies on the change in distances along the direction of travel is that of mathematician José Natário. In his metric, spacetime contracts towards the prow of the ship and expands in the direction perpendicular to the motion, meaning that the bubble actually "slides" through space, roughly speaking by "pushing space aside".

The Alcubierre drive remains a hypothetical concept with seemingly difficult problems, although the amount of energy required is no longer thought to be unobtainably large. However, Alexey Bobrick and Gianni Martire claim that, in principle, a class of subluminal, spherically symmetric warp drive spacetimes can be constructed based on physical principles presently known to humanity, such as positive energy. Furthermore, a study by Remo Garattini and Kirill Zatrimaylov shows that the amount of negative energy density required to sustain a warp bubble can in principle be reduced if the bubble is moving in an external gravitational field, such as that of a black hole.

Mathematics

Using the ADM formalism of general relativity, the spacetime is described by a foliation of space-like hypersurfaces of constant coordinate time t, with the metric taking the following general form:

${\displaystyle d{s}^2=-\left({\alpha }^2-{\beta }_i{\beta }^i\right)d{t}^2+2{\beta }_{i}d{x}^{i}dt+{\gamma }_{ij}d{x}^{i}d{x}^j,}$

where

• α is the lapse function that gives the interval of proper time between nearby hypersurfaces,
• βⁱ is the shift vector that relates the spatial coordinate systems on different hypersurfaces,
• γ_ij is a positive-definite metric on each of the hypersurfaces.

The particular form that Alcubierre studied is defined by:

${\displaystyle \begin{array}{rl}\alpha & =1,\\ {\beta }^x& =-v_s(t)f(r_s(t)),\\ {\beta }^y& ={\beta }^z=0,\\ {\gamma }_{ij}& ={\delta }_{ij},\end{array}}$

where

${\displaystyle \begin{array}{rl}{v}_s(t)& =\frac{d{x}_s(t)}{dt},\\ r_s(t)& =\sqrt{(x-x_s(t){)}^2+y^2+z^2},\\ f(r_s)& =\frac{\tanh (\sigma (r_s+R))-\tanh (\sigma (r_s-R))}{2\tanh (\sigma R)},\end{array}}$

with arbitrary parameters R > 0 and σ > 0. Alcubierre's specific form of the metric can thus be written:

${\displaystyle d{s}^2=\left(v_s(t{)}^{2}f(r_s(t){)}^2-1\right)d{t}^2-2v_s(t)f(r_s(t))dxdt+d{x}^2+d{y}^2+d{z}^2.}$

With this particular form of the metric, it can be shown that the energy density measured by observers whose 4-velocity is normal to the hypersurfaces is given by:

${\displaystyle -\frac{c^4}{8\pi G}\frac{v_s^2\left(y^2+z^2\right)}{4g^2{r}_s^2}{\left(\frac{df}{d{r}_s}\right)}^2,}$

where g is the determinant of the metric tensor.

Thus, because the energy density is negative, one needs exotic matter to travel more quickly than the speed of light. The existence of exotic matter is not theoretically ruled out; however, generating and sustaining enough exotic matter to perform feats such as faster-than-light travel (and to keep open the "throat" of a wormhole) is thought to be impractical. According to writer Robert Low, within the context of general relativity it is impossible to construct a warp drive in the absence of exotic matter.

Connection to dark energy and dark matter

See also: Negative mass

Astrophysicist Jamie Farnes from the University of Oxford has proposed a theory, published in the peer-reviewed scientific journal Astronomy & Astrophysics, that unifies dark energy and dark matter into a single dark fluid, and which is expected to be testable by the Square Kilometre Array around 2030. Farnes found that Albert Einstein had explored the idea of gravitationally repulsive negative masses while developing the equations of general relativity, an idea which leads to a "beautiful" hypothesis where the cosmos has equal amounts of positive and negative qualities. Farnes' theory relies on negative masses that behave identically to the physics of the Alcubierre drive, providing a natural solution for the current "crisis in cosmology" due to a time-variable Hubble parameter.

As Farnes' theory allows a positive mass (i.e. a ship) to reach a speed equal to the speed of light, it has been dubbed "controversial". If the theory is correct, which has been highly debated in the scientific literature, it would explain dark energy, dark matter, allow closed timelike curves (see time travel), and suggest that an Alcubierre drive is physically possible with exotic matter.

Physics

With regard to certain specific effects of special relativity, such as Lorentz contraction and time dilation, the Alcubierre metric has some apparently peculiar aspects. In particular, Alcubierre has shown that a ship using an Alcubierre drive travels on a free-fall geodesic even while the warp bubble is accelerating: its crew would be in free fall while accelerating without experiencing accelerational g-forces. Enormous tidal forces, however, would be present near the edges of the flat-space volume because of the large space curvature there, but a suitable specification of the metric would keep the tidal forces very small within the volume occupied by the ship.

The original warp-drive metric and simple variants of it happen to have the ADM form, which is often used in discussing the initial-value formulation of general relativity. This might explain the widespread misconception that this spacetime is a solution of the field equation of general relativity. Metrics in ADM form are adapted to a certain family of inertial observers, but these observers are not really physically distinguished from other such families. Alcubierre interpreted his "warp bubble" in terms of a contraction of space ahead of the bubble and an expansion behind, but this interpretation could be misleading, since the contraction and expansion actually refer to the relative motion of nearby members of the family of ADM observers.

In general relativity, one often first specifies a plausible distribution of matter and energy, and then finds the geometry of the spacetime associated with it; but it is also possible to run the Einstein field equations in the other direction, first specifying a metric and then finding the energy–momentum tensor associated with it, and this is what Alcubierre did in building his metric. This practice means that the solution can violate various energy conditions and require exotic matter. The need for exotic matter raises questions about whether one can distribute the matter in an initial spacetime that lacks a warp bubble in such a way that the bubble is created at a later time, although some physicists have proposed models of dynamical warp-drive spacetimes in which a warp bubble is formed in a previously flat space. Moreover, according to Serguei Krasnikov, generating a bubble in a previously flat space for a one-way faster-than-light trip requires forcing the exotic matter to move at local faster-than-light speeds, something that would require the existence of tachyons, although Krasnikov also notes that when the spacetime is not flat from the outset, a similar result could be achieved without tachyons by placing in advance some devices along the travel path and programming them to come into operation at preassigned moments and to operate in a preassigned manner. Some suggested methods avoid the problem of tachyonic motion, but would probably generate a naked singularity at the front of the bubble. Allen Everett and Thomas Roman comment on Krasnikov's finding (Krasnikov tube):

[The finding] does not mean that Alcubierre bubbles, if it were possible to create them, could not be used as a means of superluminal travel. It only means that the actions required to change the metric and create the bubble must be taken beforehand by some observer whose forward light cone contains the entire trajectory of the bubble.

For example, if one wanted to travel to Deneb (2,600 light-years away) and arrive less than 2,600 years in the future according to external clocks, it would be required that someone had already begun work on warping the space from Earth to Deneb at least 2,600 years ago:

A spaceship appropriately located with respect to the bubble trajectory could then choose to enter the bubble, rather like a passenger catching a passing trolley car, and thus make the superluminal journey ... as Krasnikov points out, causality considerations do not prevent the crew of a spaceship from arranging, by their own actions, to complete a round trip from Earth to a distant star and back in an arbitrarily short time, as measured by clocks on Earth, by altering the metric along the path of their outbound trip.

Difficulties

Mass–energy requirement

The metric of this form has significant difficulties because all known warp-drive spacetime theories violate various energy conditions. Nevertheless, an Alcubierre-type warp drive might be realized by exploiting certain experimentally verified quantum phenomena, such as the Casimir effect, that lead to stress–energy tensors that also violate the energy conditions, such as negative mass–energy, when described in the context of the quantum field theories.

If certain quantum inequalities conjectured by Ford and Roman hold, the energy requirements for some warp drives may be unfeasibly large as well as negative. For example, the energy equivalent of −10⁶⁴ kg might be required to transport a small spaceship across the Milky Way—an amount orders of magnitude greater than the estimated mass of the observable universe. Counterarguments to these apparent problems have also been offered, although the energy requirements still generally require a Type III civilization on the Kardashev scale.

Chris Van Den Broeck of the Katholieke Universiteit Leuven in Belgium, in 1999, tried to address the potential issues. By contracting the 3+1-dimensional surface area of the bubble being transported by the drive, while at the same time expanding the three-dimensional volume contained inside, Van Den Broeck was able to reduce the total energy needed to transport small atoms to less than three solar masses. Later in 2003, by slightly modifying the Van den Broeck metric, Serguei Krasnikov reduced the necessary total amount of negative mass to a few milligrams. Van Den Broeck detailed this by saying that the total energy can be reduced dramatically by keeping the surface area of the warp bubble itself microscopically small, while at the same time expanding the spatial volume inside the bubble. However, Van Den Broeck concludes that the energy densities required are still unachievable, as are the small size (a few orders of magnitude above the Planck scale) of the spacetime structures needed.

In 2012, physicist Harold White and collaborators announced that modifying the geometry of exotic matter could reduce the mass–energy requirements for a macroscopic space ship from the equivalent of the planet Jupiter to that of the Voyager 1 spacecraft (c. 700 kg) or less, and stated their intent to perform small-scale experiments in constructing warp fields. White proposed to thicken the extremely thin wall of the warp bubble, so the energy is focused in a larger volume, but the overall peak energy density is actually smaller. In a flat 2D representation, the ring of positive and negative energy, initially very thin, becomes a larger, fuzzy torus (donut shape). However, as this less energetic warp bubble also thickens toward the interior region, it leaves less flat space to house the spacecraft, which has to be smaller. Furthermore, if the intensity of the space warp can be oscillated over time, the energy required is reduced even more. According to White, a modified Michelson–Morley interferometer could test the idea: one of the legs of the interferometer would appear to have a slightly different length when the test devices were energised. Alcubierre has expressed skepticism about the experiment, saying: "from my understanding there is no way it can be done, probably not for centuries if at all".

In 2021, physicist Erik Lentz described a way warp drives sourced from known and familiar purely positive energy could exist—warp bubbles based on superluminal self-reinforcing "soliton" waves. The claim is controversial, with other physicists arguing that all physically reasonable warp drives violate the weak energy condition, as well as both the strong and dominant energy conditions.

Placement of matter

Krasnikov proposed that if tachyonic matter cannot be found or used, then a solution might be to arrange for masses along the path of the vessel to be set in motion in such a way that the required field was produced. But in this case, the Alcubierre drive vessel can only travel routes that, like a railroad, have first been equipped with the necessary infrastructure. The pilot inside the bubble is causally disconnected from its walls and cannot carry out any action outside the bubble: the bubble cannot be used for the first trip to a distant star because the pilot cannot place infrastructure ahead of the bubble while "in transit". For example, traveling to Vega (which is 25 light-years from Earth) requires arranging everything so that the bubble moving toward Vega with a superluminal velocity would appear; such arrangements will always take more than 25 years.

Coule has argued that schemes, such as the one proposed by Alcubierre, are infeasible because matter placed en route of the intended path of a craft must be placed at superluminal speed—that constructing an Alcubierre drive requires an Alcubierre drive even if the metric that allows it is physically meaningful. Coule further argues that an analogous objection will apply to any proposed method of constructing an Alcubierre drive.

Survivability inside the bubble

An article by José Natário (2002) argues that crew members could not control, steer or stop the ship in its warp bubble because the ship could not send signals to the front of the bubble.

A 2009 article by Carlos Barceló, Stefano Finazzi, and Stefano Liberati uses quantum theory to argue that the Alcubierre drive at faster-than-light velocities is impossible mostly because extremely high temperatures caused by Hawking radiation would destroy anything inside the bubble at superluminal velocities and destabilize the bubble itself; the article also argues that these problems are absent if the bubble velocity is subluminal, although the drive still requires exotic matter.

Damaging effect on destination

Brendan McMonigal, Geraint F. Lewis, and Philip O'Byrne have argued that were an Alcubierre-driven ship to decelerate from superluminal speed, the particles that its bubble had gathered in transit would be released in energetic outbursts akin to the infinitely-blueshifted radiation hypothesized to occur at the inner event horizon of a Kerr black hole; forward-facing particles would thereby be energetic enough to destroy anything at the destination directly in front of the ship.

Wall thickness

The amount of negative energy required for such a propulsion is not yet known. Pfenning and Allen Everett of Tufts hold that a warp bubble traveling at 10-times the speed of light must have a wall thickness of no more than 10⁻³² meters—close to the limiting Planck length, 1.6 × 10⁻³⁵ meters. In Alcubierre's original calculations, a bubble macroscopically large enough to enclose a ship of 200 meters would require a total amount of exotic matter greater than the mass of the observable universe, and straining the exotic matter to an extremely thin band of 10⁻³² meters is considered impractical. Similar constraints apply to Krasnikov's superluminal subway. Chris Van den Broeck constructed a modification of Alcubierre's model that requires much less exotic matter but places the ship in a curved spacetime "bottle" whose neck is about 10⁻³² meters.

Causality violation and semiclassical instability

Calculations by physicist Allen Everett show that warp bubbles could be used to create closed timelike curves in general relativity, meaning that the theory predicts that they could be used for backwards time travel. While it is possible that the fundamental laws of physics might allow closed timelike curves, the chronology protection conjecture hypothesizes that in all cases where the classical theory of general relativity allows them, quantum effects would intervene to eliminate the possibility, making these spacetimes impossible to realize. A possible type of effect that would accomplish this is a buildup of vacuum fluctuations on the border of the region of spacetime where time travel would first become possible, causing the energy density to become high enough to destroy the system that would otherwise become a time machine. Some results in semiclassical gravity appear to support the conjecture, including a calculation dealing specifically with quantum effects in warp-drive spacetimes that suggested that warp bubbles would be semiclassically unstable, but ultimately the conjecture can only be decided by a full theory of quantum gravity.

Alcubierre briefly discusses some of these issues in a series of lecture slides posted online, where he writes: "beware: in relativity, any method to travel faster than light can in principle be used to travel back in time (a time machine)". In the next slide, he brings up the chronology protection conjecture and writes: "The conjecture has not been proven (it wouldn't be a conjecture if it had), but there are good arguments in its favor based on quantum field theory. The conjecture does not prohibit faster-than-light travel. It just states that if a method to travel faster than light exists, and one tries to use it to build a time machine, something will go wrong: the energy accumulated will explode, or it will create a black hole."

Relation to Star Trek warp drive

The Star Trek television series and films use the term "warp drive" to describe their method of faster-than-light travel. Neither the Alcubierre theory, nor anything similar, existed when the series was conceived—the term "warp drive" and general concept originated with John W. Campbell's 1931 science fiction novel Islands of Space. Alcubierre stated in an email to William Shatner that his theory was directly inspired by the term used in the show and cites the "'warp drive' of science fiction" in his 1994 article. A USS Alcubierre appears in the Star Trek tabletop RPG Star Trek Adventures. Since the release of Star Trek: The Original Series, more recent Star Trek spin-off series have made closer use of the theory behind the Alcubierre Drive incorporating warp bubbles/fields into the in-universe science.