Quantum field theory

Quantum field theory combines special relativity with quantum theory by treating fields as operator-valued distributions and particles as field excitations.

A quantum field is an operator-valued distribution on spacetime: smearing $\Phi (x)$ against a smooth test function gives an operator, whereas pointwise products can require renormalization. For example, a free complex scalar field contains particle annihilation operators and antiparticle creation operators. Its microcausality condition makes field commutators vanish at spacelike separation. For fermions, the corresponding locality condition uses anticommutators. This combines quantum operators with the spacetime locality of a quantum field theory.

Algebraic quantum field theory assigns an observable algebra $\mathcal{A}(O)$ to each suitable spacetime region $O$. Inclusion of regions implies inclusion of algebras, and spacelike-separated algebras commute under microcausality. A state is a positive normalized functional on the observable algebra, so its restriction to a region specifies expectation values and correlations for that algebra. Spectrum conditions restrict translation representations, while primitive causality and time-slice axioms constrain the dynamics; these are different assumptions.

The time-slice axiom says that an algebra generated in an arbitrarily thin suitable neighbourhood of a Cauchy surface generates the algebra of its causal development. For $O$ such a neighbourhood in its development $D(O)$, this is $\mathcal{A}(O)=\mathcal{A}(D(O))$. Precise global and local versions can differ in the regions to which they apply. For a local hyperbolic field equation, later fields can be represented using the initial field and its conjugate momentum in the corresponding domain of dependence.

Primitive causality is a dynamical completeness condition: the state restricted to a suitable initial spacetime region determines its expectations throughout that region's domain of dependence. Algebraically one requires the later domain-of-dependence observable algebra to be contained in the initial algebra, under the identifications supplied by the dynamics. A local time-slice axiom gives this property for a neighbourhood of a Cauchy slice. Knowledge of the full restricted state, including correlations, is required; knowing only mean initial fields is insufficient.

In a translation-covariant quantum field theory on Minkowski spacetime, the joint spectrum of the energy-momentum generators is required to lie in the closed future light cone. With signature $(+,-,-,-)$ this means $p^0\ge 0$ and $p^2\ge 0$. The spectrum condition implements a positive-energy stability requirement in every inertial frame. It constrains the Fourier support and analyticity of vacuum correlation functions but does not by itself assert spacelike commutation or an initial-value property.

The joint energy-momentum spectrum lies in the closed forward light cone. In axiomatic quantum field theory, this condition permits analytic continuation of Wightman functions in suitable complex spacetime domains. Together with Lorentz invariance, microcausality and the positivity of the physical state space, it is one of the hypotheses behind the CPT theorem.

A scalar field theory describes one or more Lorentz scalar fields through an action. The displayed Lagrangian density is a common classical normalization for a real scalar field; a complex scalar field has two real components. Canonical quantization turns these classical degrees of freedom into operators. The free real scalar field gives spin-zero particles, while interactions arise from the potential $V$.

For real scalar fields with quartic potential energy density $V_4=\lambda ({\varphi }_1^4+{\varphi }_2^4)+2\mu {\varphi }_1^2{\varphi }_2^2$, nonnegativity in every direction is equivalent to the displayed conditions. Necessity follows on an axis and on ${\varphi }_1={\varphi }_2$. Sufficiency follows by writing $V_4=\lambda ({\varphi }_1^2-{\varphi }_2^2{)}^2+2(\lambda +\mu ){\varphi }_1^2{\varphi }_2^2$. With a positive quadratic mass term, equality cases still give a stable isolated classical vacuum at the origin. The quartic alone is strictly positive away from the origin precisely when $\lambda >0$ and $\mu >-\lambda$. Rotation symmetry additionally requires $\mu =\lambda$.

This first-order-in-time field theory has an annihilation-field expansion $\Psi =\int a_{\mathbf{p}}{e}^{-i{\mathbf{p}}^{2}t/(2m)+i\mathbf{p}\cdot \mathbf{x}}{d}^{3}p/(2\pi {)}^3$. Its adjoint creates particles; an antiparticle creation term is not required by a relativistic mass shell or Lorentz-invariant microcausality. It is often an effective low-energy sector of a relativistic charged theory after the antiparticle sector is omitted. This does not mean that antimatter cannot be treated nonrelativistically: one may introduce a separate nonrelativistic field for that species.

A free real scalar field obeys the Klein-Gordon equation and creates neutral spin-zero particles.

A free real scalar field with zero mass has the displayed wave equation in Minkowski spacetime. In curved spacetime its dynamics also depend on curvature coupling: a minimally coupled massless scalar field and a conformally coupled scalar field need not have the same cosmological modes.

The displayed action functional contains no mass term and no $R{\varphi }^2$ curvature coupling. Its Euler-Lagrange field equation is ${\partial }_{\mu }(\sqrt{-g}{g}^{\mu \nu }{\partial }_{\nu }\varphi )=0$. In a flat FLRW metric, a Fourier mode obeys ${\varphi }_k^{\prime \prime }+2(a^{\prime }/a){\varphi }_k^{\prime }+k^2{\varphi }_k=0$ in four dimensions. The expansion can therefore freeze these modes; the curvature term of a conformally coupled scalar field changes that conclusion.

A real scalar triplet groups three real scalar components into a vector under SO(3). Equivalently it is the three-dimensional Adjoint representation of SU(2), whose center acts trivially. Its rotationally invariant quadratic and quartic potentials depend on $\Phi \cdot \Phi$. A nonzero vacuum direction leaves rotations around that direction unbroken.

A relativistic real scalar field theory with a periodic cosine potential. In physical coordinates $y^{\mu }$, one normalization is $\mathcal{L}=\frac{1}{2}{\partial }_{\mu }\phi {\partial }^{\mu }\phi -m^2(1-\cos \beta \phi )/{\beta }^2$. With $\varphi =\beta \phi$ and $x^{\mu }=m{y}^{\mu }$, its action is $S={\beta }^{-2}\int d^{2}x[\frac{1}{2}(\partial \varphi {)}^2+\cos \varphi -1]$. Its Euler-Lagrange field equation is the Sine-Gordon equation. Distinct scalar-field vacua differ by $2\pi$ in $\varphi$, permitting a Sine-Gordon kink.

At renormalized coupling $\gamma =8\pi /n$, the kink-antikink transmission poles lie at $\vartheta =i\pi (1-k/n)$. The relativistic bound-state mass from a rapidity pole gives the displayed increasing breather masses. The kink mass is $M=mn/\pi$, and $k=n$ is an excluded threshold state. No breathers occur at $n=1$; scattering involving a physical second breather requires $n\ge 3$.

For the transmitting product with factors $\cosh [(\theta -i\pi j/N)/2]/\cosh [(\theta +i\pi j/N)/2]$ and prefactor $(-1{)}^N$, the continuous phase on $\theta \ge 0$ is $\pi N-2{\sum }_{j=1}^{N-1}\arctan [\tanh (\theta /2)\tan (\pi j/(2N))]$. Its derivative is $-{\sum }_j\sin (\pi j/N)/[\cosh \theta +\cos (\pi j/N)]$. A Riemann sum at fixed positive rapidity gives $(2N/\pi )\log \tanh (\theta /2)$ at leading order. The endpoints depend on the stated unwrapped branch; setting the high-energy value to zero silently changes the constant.

The second breather is a bound pair of first breathers with constituent shifts $\pm ia$. The bound-state fusion of factorized S-matrices gives $S_{2,1}=S_{1,1}(\vartheta +ia)S_{1,1}(\vartheta -ia)$, which factors as displayed. The nearest physical-strip pole at $\vartheta =ia$ is a crossed-channel exchange of the first breather. At $n=3$ the more distant center pole is double, without changing that nearest-pole interpretation.

At the closest physical-strip pole of $S_{2,1}$, the difference of the analytically continued external momenta has squared mass $m_1^2$. Using $m_2=2m_1\cos a$ makes the identity immediate. Thus the exchanged t-channel one-particle state is the lightest Sine-Gordon breather, not a new member of the spectrum.

For the angular field $u=\beta \varphi$ and coordinates $\xi =m(x+t)/2$, $\eta =m(x-t)/2$, let $w=(v+u)/2$ and $d=(v-u)/2$. The displayed relations imply $u_{\xi \eta }=\sin u$ and $v_{\xi \eta }=\sin v$. They define a Bäcklund transformation with the reciprocal parameter convention used in the Bianchi permutability for sine-Gordon Bäcklund transformations. Physical-field trigonometric arguments include the coupling $\beta$.

For $\tau =x+t$, $\rho =x-t$, the unit-mass Sine-Gordon equation is ${\varphi }_{\tau \rho }=\frac{1}{4}\sin \varphi$. An auto-Bäcklund transformation with these coordinates isDifferentiate each relation in the other coordinate and add or subtract; the sine addition formula gives ${\varphi }_{j,\tau \rho }=\frac{1}{4}\sin {\varphi }_j$ for $j=0,1$. Doubling both right-hand sides instead gives ${\varphi }_{j,\tau \rho }=\sin {\varphi }_j$, hence a sine-Gordon equation with mass squared four. Those doubled coefficients are appropriate for the half-scaled coordinates $(x+t)/2,(x-t)/2$.

Two compatible Bäcklund steps commute after integration constants are matched. The displayed superposition relation constructs their common output algebraically from the seed and the two one-step outputs, in the reciprocal-parameter convention of the Sine-Gordon Bäcklund transformation. It uses the angular field $u=\beta \varphi$. Smooth inverse-tangent branch continuation is required to retain the correct vacuum labels.

Write the transformed angular field as $v=u+2d$. The Sine-Gordon Bäcklund transformation gives $\sin d=a(u_{\xi }+d_{\xi })$, determining $d$ as a formal local derivative expansion in $a$. The displayed exact current identity yields a conservation law at each order. Formal convergence is unnecessary because each coefficient obeys an exact identity on solutions.

The Bäcklund expansion produces local differential-polynomial currents. In the coordinates $\xi =m(x+t)/2$, $\eta =m(x-t)/2$, their charges are $\int (P_j+Q_j)dx$ with vanishing boundary flux. Derivative improvements contribute no new charge. The first nontrivial higher current can be written $P=u_{\xi \xi }^2-u_{\xi }^4/4$, $Q=\cos u{u}_{\xi }^2$. Continuing, and exchanging light-cone directions, produces the infinite higher-spin hierarchy characterizing classical integrability.

With canonical field $\phi =\varphi /\beta$, the quartic interaction has coupling $-m^2{\beta }^2$. Cancelling the vacuum tadpole diagram requires the divergent mass counterterm $\delta m^2=(m^2{\beta }^2/4){\int }_{-\Lambda }^{\Lambda }dk/[2\pi \sqrt{k^2+1}]$. Its kink energy is $\Delta M_{\mathrm{ct}}=4\delta m^2/(m{\beta }^2)$ because $\int (1-\cos {\varphi }_K)dx=4$. This cancels the logarithmic ultraviolet divergence in the one-loop soliton mass correction. Finite parts depend on the renormalization condition.

For real parameters with ${\kappa }_i^2-b_i^2=1$, set $E_i=e^{{\kappa }_{i}x-b_{i}t+{\gamma }_i}$ and $a_{ij}=[({\kappa }_i-{\kappa }_j{)}^2-(b_i-b_j{)}^2]/[({\kappa }_i+{\kappa }_j{)}^2-(b_i+b_j{)}^2]$. Sum ${\prod }_i{E}_i^{{\mu }_i}{\prod }_{i<j}{a}_{ij}^{{\mu }_i{\mu }_j}$ over binary vectors of even parity for $f$ and odd parity for $g$. The Sine-Gordon equation solution is the continuous field $4\arg (f+ig)$. In the all-kink sector, ${\kappa }_i=\cosh {\theta }_i$, $b_i=\sinh {\theta }_i$ and $a_{ij}=-{\tanh }^2[({\theta }_i-{\theta }_j)/2]$. Treating these coefficients as positive would change the solution. Distinct rapidities give separated incoming and outgoing solitons.

The static kink ${\varphi }_K(x)=4\arctan e^{x-a}$ joins adjacent scalar-field vacua $0$ and $2\pi$. It obeys ${\varphi }_K^{\prime }=2\mathrm{sech}(x-a)$ and has topological charge $Q=[\varphi (+\infty )-\varphi (-\infty )]/(2\pi )=1$. In the Sine-Gordon theory normalization with physical mass scale $m$ and coupling $\beta$, its classical mass is $8m/{\beta }^2$. A Lorentz boost gives ${\varphi }_K=4\arctan \exp [(x-vt-a)/\sqrt{1-v^2}]$. Spatial reflection gives an antikink.

The identically conserved topological current $j$ couples to an external electromagnetic gauge potential through $-A_{\mu }{j}^{\mu }=-A_0{\theta }_x+A_1{\theta }_t$. A gauge transformation changes this coupling only by a divergence. The Euler-Lagrange field equation becomes ${\theta }_{tt}-{\theta }_{xx}+\sin \theta +E=0$, with $E={\partial }_t{A}_1-{\partial }_x{A}_0$. A unit Sine-Gordon kink has unnormalized charge $Q=2\pi$ and mass $8$. For spatially uniform forcing, its field momentum balance is $\dot{P}=QE$ when the boundary stress contributions cancel.

This localized two-soliton field has zero net winding and describes an elastic kink-antikink collision for $0<v<1$. Two reciprocal positive Bäcklund parameters produce it from the vacuum. Analytic continuation of $v$ to an imaginary value gives a real Sine-Gordon breather. The overall field sign and spacetime translations change conventions, not the field equation.

The zero-relative-speed limit of the Sine-Gordon kink-antikink scattering solution is a separatrix at twice the rest-kink energy. A Sine-Gordon Bäcklund transformation of the static kink with equal parameter gives $p_x=-p\tanh x$, $p_t=\mathrm{sech}x$ for $p=\tan (\varphi /4)$, hence the displayed solution. The two asymptotic transitions separate logarithmically in time and their speeds tend to zero; this is not a finite-period Sine-Gordon breather.

For $0<v<1$ and $\Gamma =(1-v^2{)}^{-1/2}$, this real two-soliton solution has angular winding $4\pi$. Its separated kink velocities are $\pm v$. At large times the centers satisfy $|m\Gamma x|=|m\Gamma vt|-\log v+o(1)$, giving a right-moving shift $-2\log v/(m\Gamma )$. It follows from Bianchi permutability for sine-Gordon Bäcklund transformations with oppositely signed seed parameters.

For two equal-charge Sine-Gordon kinks with speeds $\pm v$, $0<v<1$, the positive spatial transition satisfies $X_{+}(T)=v|T|-\log v/(m\gamma )+o(1)$, where $\gamma =(1-v^2{)}^{-1/2}$. Labeling a soliton by its preserved rapidity, the right-moving incoming and outgoing intercepts differ by $-2\log v/(m\gamma )$. The soliton time delay at a fixed distant location is minus this shift divided by $v$, giving the displayed negative value. Reversing the entire field gives two antikinks with the same shifts. The sign means an advance relative to extrapolation of the incoming line; labels tied to left and right positions instead exchange velocities during reflection.

The second variation of the Sine-Gordon theory action about its static kink gives ${\Delta }_x=-{\partial }_x^2+1-2{\mathrm{sech}}^{2}x$. Let $A={\partial }_x+\tanh x$. Then ${\Delta }_x=A^{†}A$ and $A{A}^{†}=-{\partial }_x^2+1$. Thus ${\Delta }_x$ is nonnegative, with its translational zero mode of a sine-Gordon kink and a continuum at ${\omega }^2=k^2+1$. The construction is the supersymmetric factorization of the one-soliton potential.

The normalized eigenfunction ${\psi }_0(x)=\mathrm{sech}x/\sqrt{2}$ has zero eigenvalue under the Sine-Gordon kink fluctuation operator. It is proportional to ${\partial }_x{\varphi }_K$ and comes from shifting the collective coordinate of the kink. Its frequency is zero, so it contributes no oscillator zero-point energy; it must be handled separately from a Gaussian functional determinant.

A scalar field configuration eigenstate assigns eigenvalues of the equal-time field operator at every spatial point. These generalized states play the role of position eigenstates for a system of infinitely many coordinates. Inserting their regulated completeness relations on time slices constructs the scalar field path integral.

For a free massive scalar in a Euclidean path integral, inversion of the quadratic kernel gives ${\Delta }_E(p)=1/(p^2+m^2)$. The Minkowski version follows by Wick rotation.

For a Hermitian scalar field in a positive-norm Hilbert space with a Lorentz-invariant vacuum, the spectrum condition and completeness give the connected vacuum two-point function as a superposition of free mass-shell functions. Define the Källén–Lehmann spectral density by $\tilde{W}(p)=2\pi \theta (p^0)\rho (p^2)$, where $W(x)=⟨0|\varphi (x)\varphi (0)|0⟩$ and its vacuum expectation value has been subtracted if nonzero. Squared intermediate-state matrix elements make $\rho$ a nonnegative measure supported on nonnegative mass squared. Applying time ordering to each free mass shell gives the displayed Feynman propagator representation, using $i\Delta =⟨T\varphi \varphi ⟩$ and signature $(+---)$. Its positivity does not automatically extend to gauge-variant fields in an indefinite auxiliary space.

The Källén–Lehmann spectral density is the nonnegative invariant-mass measure in a scalar vacuum two-point function. With complete intermediate states it satisfies $2\pi \theta (p^0)\rho (p^2)=(2\pi {)}^4{\sum }_{\alpha }|⟨\alpha |\varphi (0)|0⟩{|}^2{\delta }^4(p-p_{\alpha })$, with continuous state measures included in the sum. A stable one-particle state of mass $m$, with covariant normalization and squared field overlap $Z$, contributes $Z\delta (\sigma -m^2)$. Other particles and multiparticle states contribute additional nonnegative weight. The isolated pole of the scalar propagator then has residue $Z$.

For a real scalar field with canonical field normalization, impose $[\dot{\varphi }(t,\mathbf{x}),\varphi (t,\mathbf{y})]=-i{\delta }^3(\mathbf{x}-\mathbf{y})$. The Källén–Lehmann spectral representation also represents its vacuum commutator as the weighted sum of free commutators. The equal-time derivative of each free commutator is $-i{\delta }^3(\mathbf{x}-\mathbf{y})$ independently of mass. Comparison therefore gives the displayed sum rule, whenever the canonical commutator and spectral integral admit this distributional limit. Since the measure is nonnegative, an isolated one-particle weight obeys $0\le Z\le 1$, and is strictly positive if the field couples to that particle. Arbitrary rescaling of the field rescales the total spectral weight, so positivity alone does not imply this normalization or upper bound.

A smooth-cutoff scalar propagator is the inverse of a positive regulated quadratic kernel. It agrees with the unregulated scalar propagator at low momentum and decreases rapidly at high momentum. Smooth suppression is not identical to vanishing support. Its cutoff derivative is the line weight in an exact renormalization-group flow.

The Gaussian contraction used while integrating out a momentum shell. Its Fourier support lies entirely in that shell. In particular, convolution with a purely low-momentum field is zero. This distinguishes shell Feynman diagrams from unrestricted loop diagrams.

A nonminimal scalar coupling includes an interaction between the field and spacetime curvature, such as $-\xi R{\Phi }^2$ in the Lagrangian density. The Euler-Lagrange field equation for $\mathcal{L}=-\frac{1}{2}(\nabla \Phi {)}^2-\xi R{\Phi }^2$ is $\square \Phi -2\xi R\Phi =0$. The coefficient is convention dependent: an action written with $-\frac{1}{2}\xi R{\Phi }^2$ has $\xi$ instead of $2\xi$.

For two free real scalar fields, the internal rotation $\delta {\varphi }_1=\theta {\varphi }_2$, $\delta {\varphi }_2=-\theta {\varphi }_1$ is a symmetry precisely when their squared masses agree. Its Noether current is $j^{\mu }={\varphi }_2{\partial }^{\mu }{\varphi }_1-{\varphi }_1{\partial }^{\mu }{\varphi }_2$. Without mass degeneracy, ${\partial }_{\mu }{j}^{\mu }=(m_2^2-m_1^2){\varphi }_1{\varphi }_2$.

For the internal rotation symmetry of two real scalar fields with $\delta {\varphi }_1=\alpha {\varphi }_2$ and $\delta {\varphi }_2=-\alpha {\varphi }_1$, the Noether charge is $Q=i\int (a_1^{†}{a}_2-a_2^{†}{a}_1)d^{3}p/(2\pi {)}^3$. Its commutators satisfy $[Q,b_{\pm }^{†}]=\pm b_{\pm }^{†}$, so these creation operators produce charge eigenstates in the one-particle state sector. Substitution gives $Q=N_{+}-N_{-}$, converting the rotation of real components into opposite charges in a complex basis.

The equal-time canonical commutation relation is $[\varphi (t,\mathbf{x}),\pi (t,\mathbf{y})]=i{\delta }^3(\mathbf{x}-\mathbf{y})$ in units $\hslash =1$, with the two field-field and momentum-momentum commutators zero. The quadratic Hamiltonian operator gives $\dot{\varphi }=\pi$ and $\dot{\pi }={\nabla }^2\varphi -m^2\varphi$, hence the Klein-Gordon equation as an operator identity.

With $d{\Pi }_p=d^{3}p/[(2\pi {)}^{3}2E_p]$, a free real scalar field has mode expansion of a free field $\varphi =\int d{\Pi }_p(a{e}^{-ipx}+a^{†}{e}^{ipx})$ and commutator $[a(p),a^{†}(q)]=(2\pi {)}^{3}2E_p{\delta }^3(\mathbf{p}-\mathbf{q})$. Substitution into the quadratic Hamiltonian density gives $H=\frac{1}{2}\int d{\Pi }_p{E}_p(a^{†}a+a{a}^{†})$. The terms with two annihilation operators or two creation operators cancel by $E_p^2={\mathbf{p}}^2+m^2$. Normal ordering removes the constant vacuum energy. The commutators $[H,a(p)]=-E_{p}a(p)$ and $[H,a^{†}(p)]=E_p{a}^{†}(p)$ give the boson energy without changing under this subtraction. Replacing $a(p)$ by $\sqrt{2E_p}c(p)$ recovers the noncovariant normalization $[c(p),c^{†}(q)]=(2\pi {)}^3{\delta }^3(\mathbf{p}-\mathbf{q})$.

Insert the real scalar field mode expansion into its quadratic Hamiltonian density and use the Fourier representation of the Dirac delta function. The two-annihilator and two-creator terms cancel by $E_{\mathbf{p}}^2={\mathbf{p}}^2+m^2$. The remaining symmetric expression is $H=\frac{1}{2}\int d^{3}p{E}_{\mathbf{p}}(a^{†}a+a{a}^{†})/(2\pi {)}^3$. Normal ordering removes its constant vacuum energy, yielding the displayed operator. The canonical commutation relations give $[:H:,a_{\mathbf{p}}]=-E_{\mathbf{p}}{a}_{\mathbf{p}}$ and $[:H:,a_{\mathbf{p}}^{†}]=E_{\mathbf{p}}{a}_{\mathbf{p}}^{†}$.

For the real scalar field expansion $\varphi =\int d^{3}p[a_{\mathbf{p}}{e}^{-ipx}+a_{\mathbf{p}}^{†}{e}^{ipx}]/((2\pi {)}^3\sqrt{2E_{\mathbf{p}}})$, the equal-time fields extract $a_{\mathbf{p}}=e^{iEt}\int d^{3}x{e}^{-i\mathbf{p}\cdot \mathbf{x}}[\sqrt{E/2}\varphi +i\pi /\sqrt{2E}]$. The two mixed canonical commutation relation terms yield $[a_{\mathbf{p}},a_{\mathbf{q}}^{†}]=(2\pi {)}^3{\delta }^3(\mathbf{p}-\mathbf{q})$, while the other oscillator commutators vanish. This verifies the normalization of the mode expansion directly from the canonical fields.

Phi-fourth theory has a scalar interaction $-\lambda {\varphi }^4/4!$. In four dimensions its coupling is classically marginal, and its one-loop four-point function receives bubble corrections in the three Mandelstam channels.

For a real scalar field with interaction $-\lambda {\varphi }^4/4!$, the first Dyson series term is $-i\lambda \int {\varphi }^4{d}^{4}x/4!$. Four labeled external legs can attach to the four fields in $4!$ ways, or in $\left(\genfrac{}{}{0ex}{}{4}{2}\right)2!2!=24$ ways after separating incoming and outgoing legs. The factorial cancels and spacetime integration supplies the four-momentum conservation delta function. This gives the connected tree scattering amplitude $-\lambda$. The separate identical final-state symmetry factor belongs in the event phase-space integral and is not another vertex factor.

For a connected six-external-leg Feynman diagram made from quartic vertices, $4V=6+2I$ and $L=I-V+1$ give $V=L+2$, $I=2L+1$. A tree therefore has two vertices and one internal line; partitioning six labelled external legs into two triples gives ten tree channels. A one-loop triangle has three vertices, three internal lines and two external legs at each vertex. With external labels fixed, that triangle has symmetry factor one. Other one-loop topologies also contribute to the full amplitude.

With interaction $-{\mu }^{ϵ}\lambda {\varphi }^4/4!$ and $d=4-ϵ$, the two-point tadpole diagram has symmetry factor $1/2$. The Euclidean massive loop integral gives its pole through $I_{d,1}=-m^2/(8{\pi }^2ϵ)+O(1)$. The three four-point bubble diagrams each have symmetry factor $1/2$, and the massive scalar bubble pole in four dimensions gives the quartic counterterm. The tadpole is independent of external momentum, so no one-loop wave-function renormalization is required. A vacuum-energy counterterm is additionally needed if vacuum diagrams are retained.

With interaction $-{\lambda }_0{\varphi }^4/4!$, the tadpole diagram has symmetry factor one half. After Wick rotation and a rotationally invariant Euclidean cutoff, its self-energy is $({\lambda }_0/2){\int }_{|k|<\Lambda }{d}^{4}k(2\pi {)}^{-4}(k^2+m_0^2{)}^{-1}$. The sphere area $2{\pi }^2$ and radial integral ${\int }_0^{\Lambda }{k}^{3}dk/(k^2+m_0^2)=\frac{1}{2}[{\Lambda }^2-m_0^2\log (1+{\Lambda }^2/m_0^2)]$ give the displayed shift. It is independent of external momentum, so no one-loop momentum-dependent wave-function renormalization arises from this graph. Different regulator schemes can treat power divergences differently while defining the same chosen pole mass.

The interaction vertex $-i\lambda$ and scalar propagator $-i/(p^2-iϵ)$ give one four-point bubble diagram for each of three momentum channels, each with Feynman-diagram symmetry factor $1/2$. The two-point tadpole diagram is zero as a scaleless integral in dimensional regularization. If full proper vertices include the free quadratic kernel, ${\hat{\tau }}_2^{(0)}=-p^2$; the interaction self-energy convention instead starts at zero.

With interaction $\lambda {\varphi }^4/4!$ for one real scalar, the three one-loop four-point bubble channels give the four-dimensional renormalization-group beta function $\beta (\lambda )=3{\lambda }^2/(16{\pi }^2)+O({\lambda }^3)$. Its positive sign for $\lambda >0$ excludes asymptotic freedom and produces a perturbative Landau pole on extrapolation.

For a positive weak quartic scalar field theory coupling in four dimensions, the one-loop renormalization-group beta function is $\beta =a{\lambda }^2$, $a=3/(16{\pi }^2)>0$. Integration gives the displayed inverse-coupling law, so $\lambda$ tends to zero logarithmically as $\mu \to 0$ and grows toward the ultraviolet. The interacting Wilson-Fisher fixed point of $4-ϵ$ dimensions merges with the Gaussian fixed point at $ϵ=0$. This describes the quartic direction in the massless or critically tuned infrared regime, not disappearance of every finite-energy interaction in a massive theory. The ultraviolet Landau pole signals breakdown of that perturbative extrapolation; one loop alone does not prove a nonperturbative triviality theorem.

At quadratic order, $\mu d\lambda /d\mu =-ϵ\lambda +3{\lambda }^2/(16{\pi }^2)$ for $ϵ>0$. Solving the linear equation for $1/\lambda$ gives the displayed expression. If $0<\lambda ({\mu }_0)<{\lambda }_{\ast }$, the ultraviolet limit is zero and the infrared limit is ${\lambda }_{\ast }$. Initial data above ${\lambda }_{\ast }$ reach a Landau pole at a finite ultraviolet scale. The interacting value is the Wilson-Fisher fixed point along the quartic coupling direction; the full scalar theory also has a thermal relevant direction at the Wilson-Fisher fixed point, so a critical infrared limit requires mass tuning.

Phi cubed theory has interaction $-g{\varphi }^3/3!$, giving a cubic vertex $-ig$. In four spacetime dimensions the scalar field has mass dimension one and $g$ has mass dimension one.

The two-vertex cubic bubble in $d=6-ϵ$ dimensions has the displayed amputated pole. A Feynman parameter gives mass $m^2+x(1-x)k^2$, and $\Gamma (-1+ϵ/2)$ supplies the pole. The full propagator receives $D_0^{2}I$, while its inverse receives $-I$. This sign distinguishes a two-point insertion from an inverse-propagator correction.

With additive counterterms $\frac{1}{2}\delta Z(\partial \varphi {)}^2+\frac{1}{2}\delta m^2{\varphi }^2$, their Euclidean propagator insertion is $-(\delta Z{k}^2+\delta m^2)$. These minimal subtraction scheme coefficients cancel the cubic bubble insertion. A bare mass shift also subtracts $m^2\delta Z$ after wavefunction renormalization; it is not the same coefficient as the additive mass counterterm.

A canonically normalized real scalar field in six dimensions has mass dimension two. The cubic coupling is classically a marginal coupling, whereas a local quartic coupling has mass dimension minus two. This theory is perturbatively renormalizable by power counting in quantum field theory, but a real cubic Euclidean potential is unbounded below. Its expansion about a massive Gaussian theory is therefore formal perturbation theory, not a convergent positive functional integral at nonzero real $g$.

The self-energy bubble diagram at loop order one in six-dimensional cubic scalar field theory has Feynman-diagram symmetry factor two. With positive Euclidean scalar propagators, its amputated Green's function is the displayed integral. Its superficial degree of divergence is two. Expanding at large internal momentum and taking the angular average shows that its divergent part is a constant plus a multiple of $p^2$, precisely the mass counterterm and wave-function renormalization structures.

For a spherical ultraviolet cutoff $|k|\le \Lambda$ and $m^2>0$, its value at $p=0$ isThe divergent $p^2$ coefficient is $-g^2\log ({\Lambda }^2/m^2)/(768{\pi }^3)$. Finite terms depend on the regularization convention. Tadpole diagram insertions are excluded from the one-particle-irreducible Feynman diagram, and are removed from the connected two-point function by imposing a vanishing one-point function.

For a cubic graph with $E$ external legs, $3V=2I+E$ and $L=I-V+1$. Its superficial degree of divergence is therefore $6L-2I=6-2E$. Divergences in vacuum, one-, two-, and three-point functions can be absorbed into vacuum energy, a linear term, mass, wave-function renormalization, and cubic-coupling counterterms. Higher-point proper diagrams are superficially convergent after subtraction of divergent subgraphs. This establishes perturbative renormalizability, without asserting a nonperturbative positive measure for the unstable real cubic potential.

In a six-dimensional cubic scalar field theory, three labelled box Feynman diagrams give the one-loop local quartic vertex at zero external momentum. With an effective-action term $g_4{\varphi }^4/4!$, their contribution is $\delta g_4=-3g^4{I}_4$. To check its sign and multiplicity, expand the one-loop scalar effective action as $\frac{1}{2}\mathrm{Tr}\log (D+g\varphi )$, where $D=-{\partial }^2+m^2$. Its fourth-order term is $-g^4\mathrm{Tr}(D^{-1}\varphi {)}^4/8$, giving the stated coefficient. The associated amputated connected diagram insertion has the opposite sign.