Introduction to Renormalization A lecture given by Prof. Wayne Repko Notes: Michael Flossdorf April 4th 2007 Contents 1 Introduction 1 2 RenormalizationoftheQEDLagrangian 2 3 One Loop Correction to the Fermion Propagator 3 4 Dimensional Regularization 6 4.1 SuperficialDegreeofDivergence. 6 4.2 TheProcedureofDimensionalRegularization . .. 8 4.3 AUsefulGeneralIntegral. 9 4.4 Dirac Matrices in n Dimensions .................. 11 4.5 Dimensional Regularization of Σ2 ................. 11 5 Renormalization 12 5.1 Summary ofLastSection and OutlineoftheNextSteps . .. 12 5.2 MomentumSubtractionSchemes. 13 5.3 BacktoRenormalizationinQED. 14 5.4 Leading Order Expression of Σ2 for Small ǫ ............ 17 6 Outline of the Following Procedure 19 7 The Vertex Correction 19 8 Vacuum Polarization 22 CONTENTS 2 9 The Beta Function of QED 25 Appendix 27 References 30 1 INTRODUCTION 1 First Part of the Lecture 1 Introduction Starting with the Lagrangian of any Quantum Field Theory, we have seen in some of the previous lectures how to obtain the Feynman rules, which sufficiently de- scribe how to do pertubation theory in that particular theory. But beyond tree level, the naive calculation of diagrams involving loops will often yield infinity, since the integrals have to be performed over the whole momentum space. Renormal- ization Theory deals with the systematic isolation and removing of these infinities from physical observables. The first important insight is, that it is not the fields or the coupling constants which represent measurable quantities. Measured are cross sections, decay width, etc.. As long as we make sure that this observables are finite in the end and can be unambiguously derived from the Lagrangian, we are free to introduce new quantities, called renormalized quantities for every, so called, bare quantity. We can then go a step further and consider the bare quantities to be infinite in a way which would just cancel the infinities coming from our loop calculation. In other words, if we would be able to arrange the infinities of the bare quantities and the infinities coming from the divergent integrals to cancel each other systematically, we would be left with a finite, physically meaningful theory. This is, roughly speaking, what Renormalization Theory does. The general procedure is now done in two different steps. At first, to give the above men- tioned cancelation of infinities mathematically meaning, we need to regularize the divergent integrals. That is, we first have to make the integrals finite, e.g. by im- posing a momentum cut-off. Then we are free to manipulate them. Secondly, the divergent parts of the integrals have to be absorbed into the bare quantities of the Lagrangian. We have to expect that there are different ways to treat the finite parts of the integrals, since a finite quantity can always be absorbed into an infi- nite one, and also that the way the divergences reside in the regularized integrals will depend on the Regularization procedure we choose. As we will work out, the freedom in the treatment of the finite parts is reflected in the existence of differ- ent renormalization schemes and also in the occurrence of scale quantities, which will have mass dimension. The fact that one choice of scale should be as good as any other and should not affect measurable quantities, will lead to the so called Renormalization Group consisting of transformations between different scales. If there exists a consistent way to perform this procedure for a particular the- 2 RENORMALIZATION OF THE QED LAGRANGIAN 2 ory, this theory is said to be renormalizable. QED was historically the first impor- tant theory which could be showed to be renormalizable and we shall see how this can be done explicitly in next to leading order. 2 Renormalization of the QED Lagrangian In the following we will denote the bare, unrenormalized quantities m0, ψ0, etc. to distinguish them from the renormalized ones. 1 = ψ¯ i∂/ m ψ e ψ¯ γµψ A F 0µνF + LQED 0 − 0 0 − 0 0 0 0µ − 4 0µν LG.F. µ 1 µν Z ψi¯ ∂ψ/ Z′ mψψ¯ Z eψγ¯ ψA Z F F + (1) ≡ 2 − 2 − 1 µ − 3 4 µν LG.F. 1 µ Here G.F. = 2α (∂ Aµ) denotes the Faddeev and Popov Gauge Fixing term. In the followingL we− will not take into account this term in our Renormalization pro- cedure. For a renormalization including this term, see for example [3]. Whenever we will have to choose a gauge, we will pick the Feynman gauge, i.e. set α =1. From the first and last term we see ψ0 = Z2ψ, A0µ = Z3Aµ. Similarly from the mass term:p p ¯ Z2′ ¯ Z2′ mψψ = mψ0ψ0 Z2 Z2 m = m0. ⇒ Z2′ We now define m = m δm and get 0 − m0 m δm δm Z′ = Z = Z − = 1 Z . 2 2 m 2 m − m 2 The remaining term yields: e Z1 ψ¯ γµψ A = e ψ¯ γµψ A Z2√Z3 0 0 0µ 0 0 0 0µ (2) e = Z2 √Z e . ⇒ Z1 3 0 Note that the covariant derivative now reads ren = ∂ ie Z1 A . In order to Dµ µ − Z2 µ have gauge invariance preserved, Z1 = Z2 should hold. We will find this relation in the next lecture. 3 ONE LOOP CORRECTION TO THE FERMION PROPAGATOR 3 Summarizing the results from above, we have: δm Z′ = 1 Z , (3) 2 − m 2 Z2 e = Z3e0, (4) Z1 p ψ0 = Z2ψ, (5) A0µ = pZ3Aµ. (6) We can now rewrite the Lagrangian in thep form = + , L Lren Lcounter where 1 ψ¯ i∂/ m ψ eψγ¯ µψA F µνF (7) Lren ≡ − − µ − 4 µν (Z 1) ψ¯ i∂/ m ψ + δmZ ψψ¯ (8) Lcounter ≡ 2 − − 2 1 (Z 1)eψγ¯ µψA (Z 1) F µνF (9) − 1 − µ − 3 − 4 µν (10) We now introduce the interacting part of the hamiltonian: µ H′ eψγ¯ ψA . (11) ≡ µ −Lcounter Note that the whole counter-term Lagrangian is treated as a perturbation. Remark: Divergences of processes which do not have a tree level correspon- dence, as the photon-photon scattering (see Fig. 1) would also have to be renor- malized by counter. By naive power counting, see section 4, this diagram seems to be logarithmicL divergent. It would now be hardly imaginable to absorb this divergence into a counter term, since there is no corresponding tree level process. But as it turns out when one actually does the calculation, this diagram is not divergent at all. 3 One Loop Correction to the Fermion Propagator In this section we will look at the one loop correction to the fermion propaga- tor. The relation to the counter-term Lagrangian will be worked out in the later sections. 3 ONE LOOP CORRECTION TO THE FERMION PROPAGATOR 4 Figure 1: Box diagram of photon-photon scattering. The scattering amplitude is given by 4 S = T exp i d xH′(x) − Z 2 4 ( i) 4 4 =1 i d xH′(x)+ − d x d x′ T [H′(x)H′(x)] + ... − 2! Z Z Z We will use Wick’s theorem to evaluate the time ordered products. See (A.1- A.4) in the appendix for the definition of the field contractions. We will introduce a photon mass λ to regularize the infrared divergence in the loop integral. Fig. 2 shows the process for which we will calculate the amplitude now. To use Wicks theorem we now have to take into account all possible contrac- tions which correspond to this process. The factor of two in the calculation below is due to the fact that, in our case, there are actually two equivalent ways to do the contractions. k p q = p + k p′ = p Figure 2: One loop correction to the fermion propagator. 3 ONE LOOP CORRECTION TO THE FERMION PROPAGATOR 5 p′,s′ S 1 p,s h | − | i 2 ( ie) 4 4 µ ν =2 − p′,s′ d x d x′ψ¯(x)γ ψ(x)A (x) ψ¯(x′)γ ψ(x′)A (x′) p,s 2! h | µ ν | i Z 2 4 4 ip′x s′ µ ν ipx′ s = e d x d x′ e u¯ (p′)γ iS (x x′)iD (x x′)γ e− u (p) − Fαβ − Fµν − Z Z 2 4 4 s′ µ /q + m 1 ν s = e (i)( i) d qd k u¯ (p′)γ g γ u (p) − − q2 m2 + iǫ µν k2 λ2 + iǫ Z − − 1 4 4 ip′x ipx′ iq(x x′) i( k)(x x′) d x d x′ e e− e− − e− − − ×(2π)8 Z Z This gives us two δ-functions: δ(q p k) and δ(p′ q + k). Note that the sign of k depends on how we choose the− momentum− direction− in the loop and that we choose k to be consistent with our assignments in Fig. 2. Now we perform the q-integration:− 2 s′ 4 µ /p+k/+m 1 s = e u¯ (p )δ(p p) d kγ 2 2 2 2 γ u (p) ′ ′ (p+k) m +iǫ k λ +iǫ µ − − ′ 4 4 − s − s = (2π) δ(p′ p) iM (2π) δ(p′ p) u¯ (p′)( iΣ ) u (p) . − ≡R − − 2 Where M denotes the matrix element which we could have obtained directly by using the Feynman rules: iM = sum of all connected, amputated diagrams. At this point we should stress one subtlety: We introduced a counter-term La- grangian which will lead to additional Feynman rules. Also, we did not take into account the counter-terms in the interacting Hamiltonian we used to calculate the scattering amplitude. We shall come back to that in section 5. Note that we in general want the calculated diagram to be part of a bigger one, so we don’t include the Dirac spinors from the external lines in our definition of Σ2.