Peskin&Schroeder Solution Manual

Solutions to Peskin & Schroeder

Chapter 2

Zhong-Zhi Xianyu

Institute of Modern Physics and Center for High Energy Physics,

Tsinghua University, Beijing, 100084

Draft version: November 8, 2012

1 Classical electromagnetism

In this problem we do some simple calculation on classical electrodynamics. The

action without source term is given by:

S = 14

Zd4xFF

; with F = @A @A: (1)

(a) Maxwell's equations We now derive the equations of motion from the action.

Note that

@F@(@A)

= ;

@F@A

= 0:

Then from the rst equality we get:

@

@(@A)

FF

= 4F:

Now substitute this into Euler-Lagrange equation, we have

0 = @

@L@(@A)

@L@A

= @F (2)

This is sometimes called the \second pair" Maxwell's equations. The so-called \rst

pair" comes directly from the denition of F = @A @A, and reads

@F + @F + @F = 0: (3)

The familiar electric and magnetic eld strengths can be written as Ei = F 0i andijkBk = F ij , respectively. From this we deduce the Maxwell's equations in terms ofEi and Bi:

@iEi = 0; ijk@jBk @0Ei = 0; ijk@jEk = 0; @iBi = 0: (4)E-mail: [email protected]

1

Notes by Zhong-Zhi Xianyu Solution to P&S, Chapter 2 (draft version)

(b) The energy-momentum tensor The energy-momentum tensor can be dened

to be the Nother current of the space-time translational symmetry. Under space-time

translation the vector A transforms as,

A = @A : (5)

Thus~T =

@L@(@A)

@A L = F@A + 14FF

: (6)

Obviously, this tensor is not symmetric. However, we can add an additional term @K

to ~T with K being antisymmetric to its rst two indices. It's easy to see that this

term does not aect the conservation of ~T . Thus if we choose K = FA , then:

T = ~T + @K = FF +

1

4FF

: (7)

Now this tensor is symmetric. It is called the Belinfante tensor in literature. We can

also rewrite it in terms of Ei and Bi:

T 00 =1

2(EiEi +BiBi); T i0 = T 0i = ijkEjBk; etc. (8)

2 The complex scalar eld

The Lagrangian is given by:

L = @@m2: (9)

(a) The conjugate momenta of and :

=@L@ _

= _; ~ =@L@ _

= _ = : (10)

The canonical commutation relations:

[(x); (y)] = [(x); (y)] = i(x y); (11)

The rest of commutators are all zero.

The Hamiltonian:

H =

Zd3x

_+ _ L = Z d3x +r r+m2: (12)

(b) Now we Fourier transform the eld as:

(x) =

Zd3p

(2)31p2Ep

ape

ipx + bypeipx

; (13)

thus:

(x) =Z

d3p

(2)31p2Ep

bpe

ipx + aypeipx

: (14)

2


Feed all these into the Hamiltonian:

H =

Zd3x

_ _+r r+m2

=

Zd3x

Zd3p

(2)3p2Ep

d3q

(2)3p2Eq

EpEq

aype

ipx bpeipx

aqeiqx byqeiqx

+ p q

aype

ipx bpeipx

aqeiqx byqeiqx

+m2

aype

ipx + bpeipx

aqeiqx + byqe

iqx

=

Zd3x

Zd3p

(2)3p2Ep

d3q

(2)3p2Eq

(EpEq + p q+m2)

aypaqe

i(pq)x + bpbyqei(pq)x

(EpEq + p qm2)

bqaqe

i(p+q)x + aypbyqe

i(p+q)x

=

Zd3p

(2)3p2Ep

d3q

(2)3p2Eq

(EpEq + p q+m2)

aypaqe

i(EpEq)t + bpbyqei(EpEq)t

(2)3(3)(p q)

(EpEq + p qm2)bqaqe

i(Ep+Eq)t + aypbyqe

i(Ep+Eq)t(2)3(3)(p+ q)

=

Zd3x

E2p + p2 +m2

2Ep

aypap + bpb

yp

=

Zd3xEp

aypap + b

ypbp + [bp; b

yp]: (15)

Note that the last term contributes an innite constant. It is normally explained as the

vacuum energy. We simply drop it:

H =

Zd3xEp

aypap + b

ypbp

: (16)

Where we have used the mass-shell condition: Ep =pm2 + p2. Hence we at once nd

two sets of particles with the same mass m.

(c) The theory is invariant under the global transformation: ! ei, ! ei.The corresponding conserved charge is:

Q = i

Zd3x

_ _: (17)

3


Rewrite this in terms of the creation and annihilation operators:

Q = i

Zd3x

_ _

= i

Zd3x

Zd3p

(2)3p2Ep

d3q

(2)3p2Eq

bpe

ipx + aypeipx

@@t

aqe

iqx + byqeiqx

@@t

bpe

ipx + aypeipx

aqe

iqx + byqeiqx

=

Zd3x

Zd3p

(2)3p2Ep

d3q

(2)3p2Eq

Eq

bpe

ipx + aypeipx

aqe

iqx byqeiqx

Epbpe

ipx aypeipx

aqeiqx + byqe

iqx

=

Zd3x

Zd3p

(2)3p2Ep

d3q

(2)3p2Eq

(Eq Ep)

bpaqe

i(p+q)x aypbyqei(p+q)x

+ (Eq + Ep)aypaqe

i(pq)x bpbyqei(pq)x

=

Zd3p

(2)3p2Ep

d3q

(2)3p2Eq

(Eq Ep)

bpaqe

i(Ep+Eq)t aypbyqei(Ep+Eqt)(2)3(3)(p+ q)

+ (Eq + Ep)aypaqe

i(EpEq)t bpbyqei(EpEq)t(2)3(3)(p q)

=

Zd3p

(2)32Ep 2Ep(aypap bpbyp)

=

Zd3p

(2)3aypap bypbp

; (18)

where the last equal sign holds up to an innitely large constant term, as we did when

calculating the Hamiltonian in (b). Then the commutators follow straightforwardly:

[Q; ay] = ay; [Q; by] = by: (19)We see that the particle a carries one unit of positive charge, and b carries one unit of

negative charge.

(d) Now we consider the case with two complex scalars of same mass. In this case the

Lagrangian is given by

L = @yi@i m2yii; (20)where i with i = 1; 2 is a two-component complex scalar. Then it is straightforward to

see that the Lagrangian is invariant under the U(2) transformation i ! Uijj with Uija matrix in fundamental (self) representation of U(2) group. The U(2) group, locally

isomorphic to SU(2)U(1), is generated by 4 independent generators 1 and 12 a, witha Pauli matrices. Then 4 independent Nother currents are associated, which are given

by

j = @L@(@i)

i @L@(@i )

i = (@i )(ii) (@i)(ii )

ja =@L

@(@i)ai @L

@(@i )ai =

i

2

h(@

i )ijj (@i)ijj

i: (21)

4


The overall sign is chosen such that the particle carry positive charge, as will be seen in

the following. Then the corresponding Nother charges are given by

Q = iZ

d3x_ii i _i

;

Qa = i2

Zd3x

_i (

a)ijj i (a)ij _j: (22)

Repeating the derivations above, we can also rewrite these charges in terms of creation

and annihilation operators, as

Q =

Zd3p

(2)3

ayipaip byipbip

;

Qa =1

2

Zd3p

(2)3

ayip

aijaip byipaijbip

: (23)

The generalization to n-component complex scalar is straightforward. In this case

we only need to replace the generators a=2 of SU(2) group to the generators ta in the

fundamental representation with commutation relation [ta; tb] = ifabctc.

Then we are ready to calculate the commutators among all these Nother charges and

the Hamiltonian. Firstly we show that all charges of the U(N) group commute with the

Hamiltonian. For the U(1) generator, we have

[Q;H] =

Zd3p

(2)3d3q

(2)3Eq

hayipaip byipbip

;ayjqajq + b

yjqbjq

i=

Zd3p

(2)3d3q

(2)3Eq

ayip[aip; a

yjq]ajq + a

yjq[a

yip; ajq]aip + (a! b)

=

Zd3p

(2)3d3q

(2)3Eq

ayipaiq ayiqaip + (a! b)

(2)3(3)(p q)

= 0: (24)

Similar calculation gives [Qa;H] = 0. Then we consider the commutation among internal

U(N) charges:

[Qa; Qb] =

Zd3p

(2)3d3q

(2)3

hayipt

aijajp byiptaijbjp

;aykqt

bk`a`q bykqtbk`b`q

i=

Zd3p

(2)3d3q

(2)3

ayipt

aijt

bj`a`q aykqtbk`ta`jajp + (a! b)

(2)3(3)(p q)

= ifabcZ

d3p

(2)3

ayipt

cijajp byiptcijbjp

= ifabcQc; (25)

and similarly, [Q;Q] = [Qa; Q] = 0.

3 The spacelike correlation function

We evaluate the correlation function of a scalar eld at two points:

D(x y) = h0j(x)(y)j0i; (26)with x y being spacelike. Since any spacelike interval x y can be transformed to aform such that x0 y0 = 0, thus we will simply take:

x0 y0 = 0; and jx yj2 = r2 > 0: (27)

5


Now:

D(x y) =Z

d3p

(2)31

2Epeip(xy) =

Zd3p

(2)31

2pm2 + p2

eip(xy)

=1

(2)3

Z 20

d'

Z 11

d cos

Z 10

dpp2

2pm2 + p2

eipr cos

=i

2(2)2r

Z 11

dppeiprpm2 + p2

(28)

Now we make the path deformation on p-complex plane, as is shown in Figure 2.3 in

Peskin & Schroeder. Then the integral becomes

D(x y) = 142r

Z 1m

derp2 m2 =

m

42rK1(mr): (29)

6


Chapter 3

Zhong-Zhi Xianyu




1 Lorentz group

The Lorentz group can be generated by its generators via exponential mappings.

The generators satisfy the following commutation relation:

[J ; J] = i(gJ gJ gnuJ + gJ): (1)

(a) Let us redene the generators as Li = 12 ijkJjk (All Latin indices denote spatial

components), where Li generate rotations, and Ki generate boosts. The commutators

of them can be deduced straightforwardly to be:

[Li; Lj ] = iijkLk; [Ki;Kj ] = iijkLk: (2)

If we further dene J i =12 (L

i iKi), then the commutators become

[J i; Jj] = i

ijkJk; [Ji+; J

j] = 0: (3)

Thus we see that the algebra of the Lorentz group is a direct sum of two identical algebra

su(2).

(b) It follows that we can classify the nite dimensional representations of the Lorentz

group by a pair (j+; j), where j = 0; 1=2; 1; 3=2; 2; are labels of irreducible repre-sentations of SU(2).

We study two specic cases.

1. ( 12 ; 0). Following the denition, we have Ji+ represented by

12

i and J i representedby 0. This implies

Li = (J i+ + Ji) =

12

i; Ki = i(J i+ J i) = i2i: (4)

Hence a eld under this representation transforms as:

! eiii=2ii=2 : (5)E-mail: [email protected]

1


2. ( 12 ; 0). In this case, Ji+ ! 0, J i ! 12 i. Then

Li = (J i+ + Ji) =

12

i; Ki = i(J i+ J i) = i2i: (6)Hence a eld under this representation transforms as:

! eiii=2+ii=2 : (7)

We see that a eld under the representation (12 ; 0) and (0;12 ) are precisely the left-handed

spinor L and right-handed spinor R, respectively.

(c) Let us consider the case of ( 12 ;12 ). To put the eld associated with this represen-

tation into a familiar form, we note that a left-handed spinor can also be rewritten as

row, which transforms under the Lorentz transformation as:

TL2 ! TL2

1 + i2

ii + 12 ii: (8)

Then the eld under the representation (12 ;12 ) can be written as a tensor with spinor

indices:

R TL

2 V = V 0 + V 3 V 1 iV 2V 1 + iV 2 V 0 V 3

!: (9)

In what follows we will prove that V is in fact a Lorentz vector.

A quantity V is called a Lorentz vector, if it satises the following transformation

law:

V ! V ; (10)where =

i2!(J) in its innitesimal form. We further note that:

(J) = i(

): (11)

and also, !ij = ijkk, !0i = !i0 = i, then the combination V = V ii + V 0

transforms according to

V ii !ij

i

2!mn(J

mn)ij

V ji +

i

2!0n(J

0n)i0 i

2!n0(J

n0)0i

V 0i

=ij i2 mnkk(i)(mi nj mj ni )

V ji +

ii(i)(ni )V 0i=V ii ijkV ijk + V 0ii;

V 0 ! V 0 + i

2!0n(J

0n)0i i2!n0(J

n0)0i

V i

= V 0 + ii(ini )V i = V 0 + iV i:

In total, we have

V !i ijkjk + iV i + (1 + ii)V 0: (12)

If we can reach the same conclusion by treating the combination V a matrix trans-

forming under the representation (12 ;12 ), then our original statement will be proved. In

fact:

V !1 i

2jj +

1

2jj

V

1 +

i

2jj +

1

2jj

=i +

i

2j [i; j ] +

1

2jfi; jg

V i + (1 + ii)V 0

=i ijkjk + iV i + (1 + ii)V 0; (13)

as expected. Hence we proved that V is a Lorentz vector.

2


2 The Gordon identity

In this problem we derive the Gordon identity,

u(p0)u(p) = u(p0) p0 + p

2m+

i(p0 p)2m

u(p): (14)

Let us start from the right hand side:

RHS. =1

2mu(p0)

(p0 + p) + i(p0 p)

u(p)

=1

2mu(p0)

(p0 + p)

1

2[; ](p0 p)

u(p)

=1

2mu(p0)

12f; g(p0 + p)

1

2[; ](p0 p)

u(p)

=1

2mu(p0)

=p0 + =p

u(p) = u(p0)u(p) = LHS;

where we have used the commutator and anti-commutators of gamma matrices, as well

as the Dirac equation.

3 The spinor products

In this problem, together with the Problems 5.3 and 5.6, we will develop a formalism

that can be used to calculating scattering amplitudes involving massless fermions or

vector particles. This method can profoundly simplify the calculations, especially in the

calculations of QCD. Here we will derive the basic fact that the spinor products can be

treated as the square root of the inner product of lightlike Lorentz vectors. Then, in

Problem 5.3 and 5.6, this relation will be put in use in calculating the amplitudes with

external spinors and external photons, respectively.

To begin with, let k0 and k1 be xed four-vectors satisfying k

20 = 0, k

21 = 1 and

k0 k1 = 0. With these two reference momenta, we dene the following spinors:

1. Let uL0 be left-handed spinor with momentum k0;

2. Let uR0 = =k1uL0;

3. For any lightlike momentum p (p2 = 0), dene:

uL(p) =1p

2p k0 =puR0; uR(p) =

1p2p k=puL0: (15)

(a) We show that =k0uR0 = 0 and =puL(p) = =puR(p) = 0 for any lightlike p. That is,

uR0 is a massless spinor with momentum k0, and uL(p), uR(p) are massless spinors with

momentum p. This is quite straightforward,

=k0uR0 = =k0=k1uL0 = (2g )k0k1uL0 = 2k0 k1uL0 =k1=k0uL0 = 0; (16)

and, by denition,

=puL(p) =1p

2p k0 =p=puR0 =

1p2p k0

p2uR0 = 0: (17)

In the same way, we can show that =puR(p) = 0.

3


(b) Now we choose k0 = (E; 0; 0;E) and k1 = (0; 1; 0; 0). Then in the Weylrepresentation, we have:

=k0uL0 = 0 )

0BBB@0 0 0 0

0 0 0 2E

2E 0 0 0

0 0 0 0

1CCCAuL0 = 0: (18)Thus uL0 can be chosen to be (0;

p2E; 0; 0)T , and:

uR0 = =k1uL0 =

0BBB@0 0 0 1

0 0 1 0

0 1 0 01 0 0 0

1CCCAuL0 =0BBB@

0

0

p2E0

1CCCA : (19)Let p = (p0; p1; p2; p3), then:

uL(p) =1p

2p k0 =puR0

=1p

2E(p0 + p3)

0BBB@0 0 p0 + p3 p1 ip20 0 p1 + ip2 p0 p3

p0 p3 p1 + ip2 0 0p1 ip2 p0 + p3 0 0

1CCCAuR0

=1p

p0 + p3

0BBB@(p0 + p3)(p1 + ip2)

0

0

1CCCA : (20)In the same way, we get:

uR(p) =1p

p0 + p3

0BBB@0

0

p1 + ip2p0 + p3

1CCCA : (21)

(c) We construct explicitly the spinor product s(p; q) and t(p; q).

s(p; q) = uR(p)uL(q) =(p1 + ip2)(q0 + q3) (q1 + iq2)(p0 + p3)p

(p0 + p3)(q0 + q3); (22)

t(p; q) = uL(p)uR(q) =(q1 iq2)(p0 + p3) (p1 ip2)(q0 + q3)p

(p0 + p3)(q0 + q3): (23)

It can be easily seen that s(p; q) = s(q; p) and t(p; q) = (s(q; p)).Now we calculate the quantity js(p; q)j2:

js(p; q)j2 =p1(q0 + q3) q1(p0 + p3)

2+p2(q0 + q3) q2(p0 + p3)

2(p0 + p3)(q0 + q3)

=(p21 + p22)q0 + q3p0 + p3

+ (q21 + q22)

p0 + p3q0 + q3

2(p1q1 + p2q2)

=2(p0q0 p1q1 p2q2 p3q3) = 2p q: (24)Where we have used the lightlike properties p2 = q2 = 0. Thus we see that the spinor

product can be regarded as the square root of the 4-vector dot product for lightlike

vectors.

4


4 Majorana fermions

(a) We at rst study a two-component massive spinor lying in (12 ; 0) representation,

transforming according to ! UL(). It satises the following equation of motion:

i@ im2 = 0: (25)

To show this equation is indeed an admissible equation, we need to justify: 1) It is

relativistically covariant; 2) It is consistent with the mass-shell condition (namely the

Klein-Gordon equation).

To show the condition 1) is satised, we note that is invariant under the simultane-

ous transformations of its Lorentz indices and spinor indices. That is U()U(1) =

. This implies

UR()UL(

1) = ;

as can be easily seen in chiral basis. Then, the combination @ transforms as @ !

UR()@UL(

1). As a result, the rst term of the equation of motion transforms as

i@! iUR()@UL(1)UL() = UR()i@

: (26)

To show the full equation of motion is covariant, we also need to show that the second

term i2 transforms in the same way. To see this, we note that in the innitesimalform,

UL = 1 iii=2 ii=2; UR = 1 iii=2 + ii=2:Then, under an innitesimal Lorentz transformation, transforms as:

! (1 iii=2 ii=2); ) ! (1 + iii=2 ii=2)

) 2 ! 2(1 + ii()i=2 i()i=2) = (1 iii=2 + ii=2)2:That is to say, 2 is a right-handed spinor that transforms as 2 ! UR()2.Thus we see the the two terms in the equation of motion transform in the same way

under the Lorentz transformation. In other words, this equation is Lorentz covariant.

To show the condition 2) also holds, we take the complex conjugation of the equation:

i()@ im2 = 0:

Combining this and the original equation to eliminate , we get

(@2 +m2) = 0; (27)

which has the same form with the Klein-Gordon equation.

(b) Now we show that the equation of motion above for the spinor can be derived

from the following action through the variation principle:

S =

Zd4x

yi @+ im

2(T2 y2)

: (28)

Firstly, let us check that this action is real, namely S = S. In fact,

S =Z

d4x

T i @ im

2(y2 T2)

5


The rst term can be rearranged as

T i @ = (T i @)T = (@y) i = yi @+ total derivative:Thus we see that S = S.

Now we vary the action with respect to y, that gives

0 =S

y= i @ im

2 22 = 0; (29)

which is exactly the Majorana equation.

(c) Let us rewrite the Dirac Lagrangian in terms of two-component spinors:

L = (i=@ m)

=y1 iT2 2

0 11 0

! m i@i@ m

! 1

i22

!= iy1

@1 + iT2

@2 imT2

21 y122

= iy1@1 + i

y2

@2 imT2

21 y122; (30)

where the equality should be understood to hold up to a total derivative term.

(d) The familiar global U(1) symmetry of the Dirac Lagrangian ! ei now be-comes 1 ! ei1, 2 ! ei2. The associated Nother current is

J = = y11 y22: (31)

To show its divergence @J vanishes, we make use of the equations of motion:

i@1 im22 = 0;i@2 im21 = 0;i(@

y1)

imT2 2 = 0;i(@

y2)

imT1 2 = 0:Then we have

@J = (@

y1)

1 + y1

@1 (@y2)2 y2@2= m

T2

21 + y1

22 T1 22 y221= 0: (32)

In a similar way, one can also show that the Nother currents associated with the global

symmetries of Majorana elds have vanishing divergence.

(e) To quantize the Majorana theory, we introduce the canonical anticommutation

relation, a(x);

yb(y)

= ab

(3)(x y);and also expand the Majorana eld into modes. To motivate the mode expansion, we

note that the Majorana Langrangian can be obtained by replacing the spinor 2 in the

Dirac Lagrangian (30) with 1. Then, according to our experience in Dirac theory, it

can be found that

(x) =

Zd3p

(2)3

rp 2Ep

Xa

haaa(p)e

ipx + (i2)aaya(p)eipxi: (33)

6


Then with the canonical anticommutation relation above, we can nd the anticommu-

tators between annihilation and creation operators:

faa(p); ayb(q)g = ab(3)(p q); faa(p); ab(q)g = faya(p); ayb(q)g = 0: (34)On the other hand, the Hamiltonian of the theory can be obtained by Legendre trans-

forming the Lagrangian:

H =

Zd3x

L _

_ L=

Zd3x

iy r+ im

2

y2 T2: (35)

Then we can also represent the Hamiltonian H in terms of modes:

H =

Zd3x

Zd3pd3q

(2)6p2Ep2Eq

Xa;b

yaa

ya(p)e

ipx + Ta (i2)aa(p)e

ipx

(pp )y(q )pq bab(q)e

iqx (i2)bayb(q)eiqx

+im

2

yaa

ya(p)e

ipx + Ta (i2)aa(p)e

ipx

(pp )y2(pq )ba

yb(q)e

iqx + (i2)bab(q)eiqx

im2

Ta aa(p)e

ipx + ya(i2)aya(p)e

ipx

(pp )T2pq bab(q)e

iqx + (i2)bayb(q)eiqx

=

Zd3x

Zd3pd3q

(2)6p2Ep2Eq

Xa;b

aya(p)ab(q)

ya

(pp )y(q )pq

+im

2(pp )y2(pq )(i2) im

2(i2)(

pp )T2pq

be

i(pq)x

+ aya(p)ayb(q)

ya

(pp )y(q )pq (i2) + im

2(pp )y2(pq )

im2(i2)(

pp )T2pq (i2)

b e

i(p+q)x

+ aa(p)ab(q)Ta

(i2)(

pp )y(q )pq + im

2(i2)(

pp )y2(pq )(i2)

im2(pp )T2pq

be

i(p+q)x

+ aa(p)ayb(q)

Ta

(i2)(pp )y(q )pq (i2) + im

2(i2)(

pp )y2(pq )

im2(pp )T2pq (i2)

b e

i(pq)x

=

Zd3p

(2)32Ep

Xa;b

aya(p)ab(p)

ya

(pp )y(p )pp

+im

2(pp )y2(pp )(i2) im

2(i2)(

pp )T2pp

b

+ aya(p)ayb(p)ya

(pp )y(p )

p~p (i2) + im

2(pp )y2(

p~p )

im2(i2)(

pp )T2

p~p (i2)

b

+ aa(p)ab(p)Ta(i2)(

pp )y(p )

p~p + im

2(i2)(

pp )y2(

p~p )(i2)

7


im2(pp )T2

p~p

b

+ aa(p)ayb(p)

Ta

(i2)(pp )y(p )pp (i2) + im

2(i2)(

pp )y2(pp )

im2(pp )T2pp (i2)

b

=

Zd3p

(2)32Ep

Xa;b

1

2

E2p + jpj2 +m2

haya(p)ab(p)

yab aa(p)ayb(p)Ta b

i=

Zd3p

(2)3Ep2

Xa

haya(p)aa(p) aa(p)aya(p)

i=

Zd3p

(2)3EpXa

aya(p)aa(p): (36)

In the calculation above, each step goes as follows in turn: (1) Substituting the mode

expansion for into the Hamiltonian. (2) Collecting the terms into four groups, charac-

terized by aya, ayay, aa and aay. (3) Integrating over d3x to produce a delta function,with which one can further nish the integration over d3q. (4) Using the following

relations to simplify the spinor matrices:

(p )2 = (p )2 = E2p + jpj2; (p )(p ) = p2 = m2; p = 12 (p p ):

In this step, the ayay and aa terms vanish, while the aay and aya terms remain. (5)Using the normalization yab = ab to eliminate spinors. (6) Using the anticommutatorfaa(p); aya(p)g = (3)(0) to further simplify the expression. In this step we have throwaway a constant term 12 Ep(3)(0) in the integrand. The minus sign of this termindicates that the vacuum energy contributed by Majorana eld is negative. With these

steps done, we nd the desired result, as shown above.

5 Supersymmetry

(a) In this problem we briey study the Wess-Zumino model. Maybe it is the simplest

supersymmetric model in 4 dimensional spacetime. Firstly let us consider the massless

case, in which the Lagrangian is given by

L = @@+ yi@+ F F; (37)

where is a complex scalar eld, is a Weyl fermion, and F is a complex auxiliary

scalar eld. By auxiliary we mean a eld with no kinetic term in the Lagrangian and

thus it does not propagate, or equivalently, it has no particle excitation. However, in

the following, we will see that it is crucial to maintain the o-shell supersymmtry of the

theory.

The supersymmetry transformation in its innitesimal form is given by:

= iT2; (38a) = F + (@)

2; (38b)

F = iy@; (38c)

where is a 2-component Grassmann variable. Now let us show that the Lagrangian

is invariant (up to a total divergence) under this supersymmetric transformation. This

8


can be checked term by term, as follows:

(@@) = i

@

y2@+ (@

) iT2@;

(yi@) =F y + T2@

i@+

yi@F +

2@@

= iF y2@+ i@T2(@

) iT2(@@)

+ iy@F + iy2@@

= iF y2@+ i@T2(@

) iT2(@2)

+ iy@F + iy2@2;

(F F ) = i(@y)F iF y@;

where we have used @@ = @2. Now summing the three terms above, we get:

L = i@hy2@+ yF + T2

@

i; (39)

which is indeed a total derivative.

(b) Now let us add the mass term in to the original massless Lagrangian:

L = mF + 12 imT2+ c.c. (40)Let us show that this mass term is also invariant under the supersymmetry transforma-

tion, up to a total derivative:

(L) = imT2F imy@+ 12 im[TF + y(2)T ()T@]2+ 12 im

T2[F + (@)2] + c.c.

= 12 imF (T2 T2) imy@ 12 im(@)y+ 12 im(@)T ()T + c.c.

= 12 imF (T2 T2) im@(y)+ 12 im(@)[

y+ T ()T ] + c.c

= im@(y) + c.c (41)

where we have used the following relations:

(2)T = 2; 2()T2 = ; T2 = T2; y = T ()T :

Now let us write down the Lagrangian with the mass term:

L = @@+ yi@+ F F +mF + 12 im

T2+ c.c.: (42)

Varying the Lagrangian with respect to F , we get the corresponding equation of motion:

F = m: (43)

Substitute this algebraic equation back into the Lagrangian to eliminate the eld F , we

get

L = @@m2+ yi@+ 12imT2+ c.c.

: (44)

Thus we see that the scalar eld and the spinor eld have the same mass.

9


(c) We can also include interactions into this model. Generally, we can write a La-

grangian with nontrivial interactions containing elds i, i and Fi (i = 1; ; n), as

L = @i @i + yi i@i + F i Fi +Fi

@W []

@i+

i

2

@2W []

@i@jTi

2j + c.c.

; (45)

where W [] is an arbitrary function of i.

To see this Lagrangian is supersymmetry invariant, we only need to check the inter-

actions terms in the square bracket:

Fi

@W []

@i+

i

2

@2W []

@i@jTi

2j + c.c.

= iy(@i) @W

@i+ Fi

@2W

@i@j(iT2j) + i

2

@3W

@i@j@k(iT2k)Ti 2j

+i

2

@2W

@i@j

hTFi +

y(2)T ()T@i2j +

Ti

2Fj +

@j2i

+ c.c.:

The term proportional to @3W=@3 vanishes. To see this, we note that the partial deriva-

tives with respect to i are commutable, hence @3W=@i@j@k is totally symmetric

on i; j; k. However, we also have the following identity:

(T2k)(Ti

2j) + (T2i)(

Tj

2k) + (T2j)(

Tk

2i) = 0; (46)

which can be directly checked by brute force. Then it can be easily seen that the

@3W=@3 term vanishes indeed. On the other hand, the terms containing F also sum

to zero, which is also straightforward to justify. Hence the terms left now are

iy(@i) @W@i

+ i@2W

@i@jy(2)T ()T (@i)2j

= i@yi

@W

@i

+ iyi

@2W

@i@j@j i @

2W

@i@jy(@i)j

= i@yi

@W

@i

; (47)

which is a total derivative. Thus we conclude that the Lagrangian (45) is supersymmet-

rically invariant up to a total derivative.

Let us end up with a explicit example, in which we choose n = 1 and W [] = g3=3.

Then the Lagrangian (45) becomes

L = @@+ yi@+ F F +gF2 + iT2+ c.c.

: (48)

We can eliminate F by solving it from its eld equation,

F + g()2 = 0: (49)

Substituting this back into the Lagrangian, we get

L = @@+ yi@ g2()2 + ig(T2 y2): (50)

This is a Lagrangian of massless complex scalar and a Weyl spinor, with 4 and Yukawa

interactions. The eld equations can be easily got from by variations.

10


6 Fierz transformations

In this problem, we derive the generalized Fierz transformation, with which one can

express (u1Au2)(u3

Bu4) as a linear combination of (u1Cu4)(u3

Du2), where A is

any normalized Dirac matrices from the following set:1; ; = i2 [

; ]; 5; 5 = i0123 :(a) The Dirac matrices A are normalized according to

tr (AB) = 4AB : (51)

For instance, the unit element 1 is already normalized, since tr (1 1) = 4. For Diracmatrices containing one , we calculate the trace in Weyl representation without loss

of generality. Then the representation of

=

0

0

!

gives tr () = 2 tr () (no sum on ). For = 0, we have tr (00) = 2 tr (122) =4, and for = i = 1; 2; 3, we have tr (ii) = 2 tr (ii) = 2 tr (122) = 4 (no sumon i). Thus the normalized gamma matrices are 0 and ii.

In the same way, we can work out the rest of the normalized Dirac matrices, as:

tr (0i0i) = 2 tr (ii) = 4; (no sum on i)tr (ijij) = 2 tr (kk) = 4; (no sum on i; j; k)

tr (55) = 4;

tr (5050) = 4; tr (5i5i) = 4:

Thus the 16 normalized elements are:1; 0; ii; i0i; ij ; 5; i50; 5i

: (52)

(b) Now we derive the desired Fierz identity, which can be written as:

(u1Au2)(u3

Bu4) =XC;D

CABCD(u1Cu4)(u3

Du2): (53)

Left-multiplying the equality by (u2Fu3)(u4

Eu1), we get:

(u2Fu3)(u4

Eu1)(u1Au2)(u3

Bu4) =XCD

CABCD tr (EC) tr (FD): (54)

The left hand side:

(u2Fu3)(u4

Eu1)(u1Au2)(u3

Bu4) = u4EAFBu4 = tr (

EAFB);

the right hand side:XC;D

CABCD tr (EC) tr (FD) =

XC;D

CABCD4EC4FD = 16CABEF ;

thus we conclude:

CABCD =116 tr (

CADB): (55)

11


(c) Now we derive two Fierz identities as particular cases of the results above. The

rst one is:

(u1u2)(u3u4) =XC;D

tr (CD)

16(u1

Cu4)(u3Du2): (56)

The traces on the right hand side do not vanish only when C = D, thus we get:

(u1u2)(u3u4) =XC

14 (u1

Cu4)(u3Cu2)

= 14

h(u1u4)(u3u2) + (u1

u4)(u3u2) +12 (u1

u4)(u3u2)

(u15u4)(u35u2) + (u15u4)(u35u2)i: (57)

The second example is:

(u1u2)(u3u4) =

XC;D

tr (CD)

16(u1

Cu4)(u3Du2): (58)

Again, the traces vanish if C 6= C / D with C a commuting number, whichimplies that C = D. That is,

(u1u2)(u3u4) =

XC

tr (CC)

16(u1

Cu4)(u3Cu2)

= 14

h4(u1u4)(u3u2) 2(u1u4)(u3u2)

2(u15u4)(u35u2) 4(u15u4)(u35u2)i: (59)

We note that the normalization of Dirac matrices has been properly taken into account

by raising or lowering of Lorentz indices.

7 The discrete symmetries P , C and T

(a) In this problem, we will work out the C, P and T transformations of the bilinear , with = i2 [

; ]. Firstly,

P (t;x) (t;x)P = i2 (t;x)0[; ]0 (t;x):

With the relations 0[0; i]0 = [0; i] and 0[i; j ]0 = [i; j ], we get:

P (t;x) (t;x)P =

( (t;x)0i (t;x); (t;x)ij (t;x): (60)

Secondly,

T (t;x) (t;x)T = i2 (t;x)(13)[; ](13) (t;x):Note that gamma matrices keep invariant under transposition, except 2, which changes

the sign. Thus we have:

T (t;x) (t;x)T =

( (t;x)0i (t;x); (t;x)ij (t;x): (61)

Thirdly,

C (t;x) (t;x)C = i2 (i02 )T(i 02)T = 02()T 02 :Note that 0 and 2 are symmetric while 1 and 3 are antisymmetric, we have

C (t;x) (t;x)C = (t;x) (t;x): (62)

12


(b) Now we work out the C, P and T transformation properties of a scalar eld .

Our starting point is

PapP = ap; TapT = ap; CapC = bp:

Then, for a complex scalar eld

(x) =

Zd3k

(2)31p2k0

hake

ikx + bykeikxi; (63)

we have

P(t;x)P =

Zd3k

(2)31p2k0

hakei(k

0tkx) + bykei(k0tkx)

i= (t;x): (64a)

T(t;x)T =

Zd3k

(2)31p2k0

hakei(k

0tkx) + bykei(k0tkx)

i= (t;x): (64b)

C(t;x)C =

Zd3k

(2)31p2k0

hbke

i(k0tkx) + aykei(k0tkx)

i= (t;x): (64c)

As a consequence, we can deduce the C, P , and T transformation properties of the

current J = i@ (@), as follows:

PJ(t;x)P = (1)s()i(t;x)@(t;x) @(t;x)(t;x)= (1)s()J(t;x); (65a)

where s() is the label for space-time indices that equals to 0 when = 0 and 1 when

= 1; 2; 3. In the similar way, we have

TJ(t;x)T = (1)s()J(t;x); (65b)CJ(t;x)C = J(t;x): (65c)

One should be careful when playing with T | it is antihermitian rather than hermitian,

and anticommutes, rather than commutes, withp1.

(c) Any Lorentz-scalar hermitian local operator O(x) constructed from (x) and (x)can be decomposed into groups, each of which is a Lorentz-tensor hermitian operator and

contains either (x) or (x) only. Thus to prove that O(x) is an operator of CPT = +1,it is enough to show that all Lorentz-tensor hermitian operators constructed from either

(x) or (x) have correct CPT value. For operators constructed from (x), this has been

done as listed in Table on Page 71 of Peskin & Schroeder; and for operators constructed

from (x), we note that all such operators can be decomposed further into a product

(including Lorentz inner product) of operators of the form

(@1 @my)(@1 @n) + c.c

together with the metric tensor . But it is easy to show that any operator of this

form has the correct CPT value, namely, has the same CPT value as a Lorentz tensor

of rank (m+n). Therefore we conclude that any Lorentz-scalar hermitian local operator

constructed from and has CPT = +1.

13


8 Bound states

(a) A positronium bound state with orbital angular momentum L and total spin S

can be build by linear superposition of an electron state and a positron state, with the

spatial wave function L(k) as the amplitude. Symbolically we have

jL; Si Xk

L(k)ay(k; s)by(k; s0)j0i:

Then, apply the space-inversion operator P , we get

P jL; Si =Xk

L(k)abay(k; s)by(k; s0)j0i = (1)LabXk

L(k)ay(k; s)by(k; s0)j0i:

(66)

Note that b = a, we conclude that P jL; Si = ()L+1jL; Si. Similarly,

CjL; Si =Xk

L(k)by(k; s)ay(k; s0)j0i = (1)L+S

Xk

L(k)by(k; s0)ay(k; s)j0i:

(67)

That is, CjL; Si = (1)L+S jL; Si. Then its easy to nd the P and C eigenvalues ofvarious states, listed as follows:

SL 1S 3S 1P 3P 1D 3D

P + + C + + +

(b) We know that a photon has parity eigenvalue 1 and C-eigenvalue 1. Thus wesee that the decay into 2 photons are allowed for 1S state but forbidden for 3S state

due to C-violation. That is, 3S has to decay into at least 3 photons.

14


Chapter 4

Zhong-Zhi Xianyu



Draft version: December 9, 2013

1 Scalar field with a classical source

In this problem we consider the theory with the following Hamiltonian:

H = H0

d3 j(t,x)(x), (1)

where H0 is the Hamiltonian for free Klein-Gordon field , and j is a classical source.

(a) Let us calculate the probability that the source creates no particles. Obviously,

the corresponding amplitude is given by the inner product of the needed in-state and

the out-state. In our case, both in- and out-state are vacuum state. Thus:

P (0) =out0|0in

2 = limt(1i)

0|ei2Ht|02

=0|T exp{ i

d4xHint

}|02 =

0|T exp{i

d4x j(x)I (x)}|02. (2)

(b) Now we expand this probability P (0) to j2. The amplitude reads:

0|T exp{i

d4x j(x)I (x)

}|0 =1 1

2

d4xd4y j(x)0|TI(x)I(y)|0j(y) +O(j4)

=1 12

d4xd4y j(x)j(y)

d3p

(2)31

2Ep+O(j4)

=1 12

d3p

(2)31

2Ep|j(p)|2 +O(j4). (3)

Thus:

P (0) = |1 12+O(j4)|2 = 1 +O(j4), (4)

where:

d3p

(2)31

2Ep|j(p)|2. (5)

E-mail: [email protected]

1


(c) We can calculate the probability P (0) exactly, to perform this, we calculate the

j2n term of the expansion:

i2n

(2n)!

d4x1 d4x2n j(x1) j(x2n)0|T(x1) (x2n)|0

=i2n(2n 1)(2n 3) 3 1

(2n)!

d4x1 d4x2n j(x1) j(x2n)

d3p1 d3pn

(2)3n1

2nEp1 Epneip1(x1x2) eipn(x2n1x2n)

=(1)n2nn!

( d3p(2)3

|j(p)|22Ep

)n=

(/2)2n!

. (6)

Thus:

P (0) =( n=0

(/2)nn!

)2= e. (7)

(d) Now we calculate the probability that the source creates one particle with mo-

mentum k. This time, we have:

P (k) =k|T exp{i

d4x j(x)I (x)

}|02 (8)

Expanding the amplitude to the first order in j, we get:

P (k) =k|0+ i

d4x j(x)

d3p

(2)3eipx2Ep

k|ap|0+O(j2)2

=i

d3p

(2)3j(p)2Ep

2Ep(2)

3(p k)2 = |j(k)|2 +O(j3). (9)

If we go on to work out all the terms, we will get:

P (k) =

n

i(2n+ 1)(2n+ 1)(2n 1) 3 1(2n+ 1)!

jn+1(k)2 = |j(k)|2e|j(k)|. (10)

(e) To calculate the probability that the source creates n particles, we write down the

relevant amplitude:

d3k1 d3kn

(2)3n2nEk1 Ekn

k1 kn|T exp{i

d4x j(x)I (x)

}|0. (11)

Expanding this amplitude in terms of j, we find that the first nonvanishing term is

the one of nth order in j. Repeat the similar calculations above, we can find that the

amplitude is:

in

n!

d3k1 d3kn

(2)3n2nEk1 Ekn

d4x1 d4xn j1 jnk1 kn|1 n|0+O(jn+2)

=in

n!

d3k1 d3kn jn(k)

(2)3n2nEk1 Ekn

n=0

(1)n2nn!

( d3p(2)3

|j(p)|22Ep

)n(12)

Then we see the probability is given by:

P (n) =n

n!e, (13)

which is a Poisson distribution.

2


(f) Its quite easy to check that the Poisson distribution P (n) satisfies the following

identities: n

P (n) = 1. (14)

N =n

nP (n) = . (15)

The first one is almost trivial, and the second one can be obtained by acting dd on

both sides of the first identity. If we apply dd again on the second identity, we get:

(N N)2 = N2 N2 = . (16)

2 Decay of a scalar particle

This problem is based on the following Lagrangian:

L = 12()

2 12M22 +

1

2()

2 12m22 . (17)

When M > 2m, a particle can decay into two particles. We want to calculate the

lifetime of the particle to lowest order in .

According to the eqn (4.86) on P.107, the decay rate is given by:d =

1

2M

d3p1d

3p2(2)6

1

4Ep1Ep2

M((0) (p1)(p2))2(2)4(4)(p p1 p2).(18)

To lowest order in , the amplitude M is given by:iM = 2i. (19)

The delta function in our case reads:

(4)(p p1 p2) = (M Ep1 Ep2)(3)(p1 + p2), (20)thus:

=1

2 2

2

M

d3p1d

3p2(2)6

1

4Ep1Ep2(2)4(M Ep1 Ep2)(3)(p1 + p2), (21)

where an additional factor of 1/2 takes account of two identical s in final state. Fur-

thermore, there are two mass-shell constraints:

m2 + p2i = E2pi

i = 1, 2. (22)

Hence:

=2

M

d3p1(2)3

1

4E2p1(2)(M 2Ep1) =

2

8M

(1 4m

2

M2

)1/2. (23)

Then the lifetime of is:

= 1 =8M

2

(1 4m

2

M2

)1/2. (24)

3 Linear sigma model

In this problem, we study the linear sigma model, provided by the following La-

grangian:

L = 12 ii 12 m2ii 14 (ii)2. (25)Where is a N -component scalar.

3


(a) We firstly compute the following differential cross sections to the leading order in

:

(12 12), (11 22), (11 11).Since the masses of all incoming and outgoing particles are identical, the cross section

is simply given by ( dd

)COM

=|M|2642s

, (26)

where s is the square of COM energy, and M is the scattering amplitude. With thehelp of Feynman rules, its quite easy to get

M(12 12) =M(11 22) = 2i;M(11 11) = 6i. (27)

Immediately, we get

(12 12) = M(11 22) = 2

162s;

(11 11) = 92

162s. (28)

(b) Now we study the symmetry broken case, that is, m2 = 2 < 0. Then, the scalarmultiplet can be parameterized as

= (1, , N1, + v)T , (29)

where v is the VEV of ||, and equals to2/ at tree level.

Substitute this into the Lagrangian, we get

L = 12 (k)2 + 12 ()2 12 (22)2 3

kk

4 4 2 2(kk) 4 (kk)2. (30)

Then its easy to read the Feynman rules from this expression:

k=

i

k2 22 ; (31a)

k=

iij

k2; (31b)

= 6iv; (31c)

i j

= 2ivij ; (31d)

= 6i; (31e)

4


i j

= 2iij ; (31f)

i j

k

= 2i(ijk + ikj + ijk). (31g)

(c) With the Feynman rules derived in (b), we can compute the amplitude

M[i(p1)j(p2) k(p3)(p4)],as:

M = (2iv)2[ is 22

ijk +i

t 22 ikj +

i

u 22 ijk

]

2i(ijk + ikj + ijk), (32)

where s, t, u are Mandelstam variables (See Section 5.4). Then, at the threshold pi = 0,

we have s = t = u = 0, and M vanishes.On the other hand, if N = 2, then there is only one component in , thus the

amplitude reduces to

M = 2i[ 22s 22 +

22

t 22 +22

u 22 + 3]

= 2i[ s+ t+ u

22+O(p4)

]. (33)

In the second line we perform the Taylor expansion on s, t and u, which are of order

O(p2). Note that s+ t+ u = 4m2 = 0, thus we see that O(p2) terms are also canceledout.

(d) We minimize the potential with a small symmetry breaking term:

V = 2ii + 4 (ii)2 aN , (34)

which yields the following equation that determines the VEV:

( 2 + ii)i = aiN . (35)Thus, up to linear order in a, the VEV i = (0, , 0, v) is

v =

2

+

a

22. (36)

Now we repeat the derivation in (b) with this new VEV, and write the Lagrangian in

terms of new field variable i and , as

L = 12 (k)2 + 12 ()2 12a

kk 12 (22)2

v3 vkk 14 4 2 2(kk) 4 (kk)2. (37)

5


The ij k amplitude is still given by

M = (2iv)2[ is 22

ijk +i

t 22 ikj +

i

u 22 ijk

]

2i(ijk + ikj + ijk). (38)

However this amplitude does not vanishes at the threshold. Since the vertices 6=

exactly even at tree level, and also s, t and u are not exactly zero in this case due to

nonzero mass of i. Both deviations are proportional to a, thus we conclude that the

amplitude M is also proportional to a.

4 Rutherford scattering

The Rutherford scattering is the scattering of an election by the coulomb field of a

nucleus. In this problem, we calculate the cross section by treating the electromagnetic

field as a given classical potential A(x). Then the interaction Hamiltonian is:

HI =

d3x eA. (39)

(a) We first calculate the T -matrix to lowest order. In fact:

outp|pin =p|T exp(i

d4xHI)|p = p|p ie

d4xA(x)p||p+O(e2)

=p|p ie

d4xA(x)u(p)u(p)ei(p

p)x +O(e2)

=(2)4(4)(p p) ieu(p)u(p)A(p p) +O(e2) (40)

But on the other hand,

outp|pin = p|S|p = p|p+ p|iT |p. (41)

Thus to the first order of e, we get:

p|iT |p = ieu(p)u(p)A(p p). (42)

(b) Now we calculate the cross section d in terms of the matrix elements iM.The incident wave packet | is defined to be:

| =

d3k

(2)3eibk2Ek

(k)|k, (43)

where b is the impact parameter.

The probability that a scattered electron will be found within an infinitesimal element

d3p centered at p is:

P =d3p

(2)31

2Ep

outp|in2

=d3p

(2)31

2Ep

d3kd3k

(2)62Ek2Ek

(k)(k)(outp|kin

)(outp|kin

)eib(kk

)

=d3p

(2)31

2Ep

d3kd3k

(2)62Ek2Ek

(k)(k)(p|iT |k

)(p|iT |k

)eib(kk

). (44)

6


In the last equality we have excluded the trivial scattering part from the S-matrix. Note

that:

p|iT |p = iM(2)(Ep Ep), (45)we have:

P =d3p

(2)31

2Ep

d3kd3k

(2)62Ek2Ek

(k)(k)|iM|2(2)2(EpEk)(EpEk)eib(kk)

(46)

The cross section d is given by:

d =

d2b P (b), (47)

thus the integration over b gives a delta function:

d2b eib(kk

) = (2)2(2)(k k). (48)

The other two delta functions in the integrand can be modified as follows:

(Ek Ek) = Ekk

(k k) =1

v(k k), (49)

where we have used |v| = v = v. Taking all these delta functions into account, we get:

d =d3p

(2)31

2Ep

d3k

(2)32Ek

1

v(k)(k)|iM|2(2)(Ep Ek). (50)

Since the momentum of the wave packet should be localized around its central value,

we can pull out the quantities involving energy Ek outside the integral:

d =d3p

(2)31

2Ep

1

2Ek

1

v(2)|M|2(Ep Ek)

d3k

(2)3(k)(k). (51)

Recall the normalization of the wave packet:

d3k

(2)3(k)(k) = 1, (52)

then:

d =d3p

(2)31

2Ep

1

2Ek

1

v|M(k p)|2(2)(Ep Ek). (53)

We can further integrate over |p| to get the differential cross section d/d:

d

d=

dp p2

(2)31

2Ep

1

2Ek

1

v|M(k p)|2(2)(Ep Ek)

=

dp p2

(2)31

2Ep

1

2Ek

1

v|M(k p)|2(2) Ek

k(p k)

=1

(4)2|M(k, )|2. (54)

In the last line we work out the integral by virtue of delta function, which constrains the

outgoing momentum |p| = |k| but leave the angle between p and k arbitrary. Thusthe amplitude M(k, ) is a function of momentum |k| and angle .

7


(c) We work directly for the relativistic case. Firstly the Coulomb potential A0 =

Ze/4r in momentum space is

A0(q) =Ze

|q|2 . (55)

This can be easily worked out by Fourier transformation, with a regulator emr in-serted:

A0(q,m)

d3x eipxemrZe

4r=

Ze

|q|2 +m2 . (56)

This is simply Yukawa potential, and Coulomb potential is a limiting case when m 0.The amplitude is given by

iM(k, ) = ieu(p)A(q)u(p) with q = p k. (57)

Then we have the module square of amplitude with initial spin averaged and final spin

summed (See 5.1 of Peskin & Schroeder for details), as:12

spin

|iM(k, )|2 = 12 e2A(q)A (q)spin

u(p)u(k)u(k)u(p)

= 12 e2A(q)A (q) tr

[(/p+m)

(/k +m)]

=2e2[2(p A)(k A) + (m2 (k p))A2]. (58)

Note that

A0(q) =Ze

|p k|2 =Ze

4|k|2 sin2(/2) , (59)

thus

12

spin

|iM(k, )|2 = Z2e4

(1 v2 sin2 2

)4|k|4v2 sin4(/2) , (60)

andd

d=

Z22(1 v2 sin2 2

)4|k|2v2 sin4(/2) (61)

In non-relativistic case, this formula reduces to

d

d=

Z22

4m2v4 sin4(/2)(62)

8


Chapter 5

Zhong-Zhi Xianyu




1 Coulomb scattering

In this problem we continue our study of the Coulomb scattering in Problem 4.4.

Here we consider the relativistic case. Let's rst recall some main points considered

before. The Coulomb potential A0 = Ze=4r in momentum space is

A0(q) =Ze

jqj2 : (1)

Then the scattering amplitude is given by

iM(k; ) = ieu(p) ~A(q)u(p) with q = p k: (2)

Then we can derive the squared amplitude with initial spin averaged and nal spin

summed, as:

1

2

Xspin

jiM(k; )j2 = 12e2 ~A(q) ~A(q)

Xspin

u(p)u(k)u(k)u(p)

=1

2e2 ~A(q) ~A(q) tr

h

(=p+m)

(=k +m)i

= 2e2h2(p ~A)(k ~A) + m2 (k p) ~A2i: (3)

Note that~A0(q) =

Ze

jp kj2 =Ze

4jkj2 sin2(=2) ; (4)

thus1

2

Xspin

jiM(k; )j2 = Z2e41 v2 sin2 2

4jkj4v2 sin4(=2) ; (5)

Now, from the result of Problem 4.4(b), we know that

d

d

=

1

(4)2

12

Xspin

jM(k; )j2=

Z221 v2 sin2 2

4jkj2v2 sin4(=2) (6)


1


% k1 k2 -

- p1 p2 %

e Z

e Z

Figure 1: The scattering of an electron by a charged heavy particle Z . All initial momentago inward and all nal momenta go outward.

This is the formula for relativistic electron scatted by Coulomb potential, and is called

Mott formula.

Now we give an alternative derivation of the Mott formula, by considering the cross

section of eZ ! eZ . When the mass of goes to innity and the charge of istaken to be Ze, this cross section will reduces to Mott formula. The relevant amplitude

is shown in Figure 1, which reads

iM = Z(ie)2u(p1)u(k1) it

U(p2)U(k2); (7)

where u is the spinor for electron and U is the spinor for muon, t = (k1 p1)2 is oneof three Mandelstam variables. Then the squared amplitude with initial spin averaged

and nal spin summed is

14

Xspin

jiMj2 = Z2e4

t2trh

(=k1 +m)

(=p1 +m)itrh

(=k2 +M)(=p2 +M)

i=

Z2e4

t2

h16m2M2 8M2(k1 p1) + 8(k1 p2)(k2 p1)

8m2(k2 p2) + 8(k1 k2)(p1 p2)i: (8)

Note that the cross section is given by dd

CM

=1

2Ee2Ejvk1 vk2 jjp1j

(2)24ECM

14

XjMj2

: (9)

When the mass of goes to innity, we have E ' ECM 'M , vk2 ' 0, and jp1j ' jk1j.Then the expression above can be simplied to d

d

CM

=1

16(2)2M2

14

XjMj2

: (10)

When M ! 1, only terms proportional to M2 are relevant in jMj2. To evaluate thissquared amplitude further, we assign each momentum a specic value in CM frame:

k1 = (E; 0; 0; k); p1 ' (E; sin ; 0; k cos );k2 ' (M; 0; 0;k); p2 ' (M;k sin ; 0;k cos ); (11)

then t = (k1 p1)2 = 4k2 sin2 2 , and

14

XjiMj2 = Z

2e4(1 v2 sin2 2 )k2v2 sin2 2

M2 +O(M): (12)

Substituting this into the cross section, and sendingM !1, we reach the Mott formulaagain:

d

d

=

Z221 v2 sin2 2

4jkj2v2 sin4(=2) (13)

2


k1

p2

k2

p1

e e+

e e+

k1

p2p1

k2e e+

e e+

Figure 2: Bhabha scattering at tree level. All initial momenta go inward and all nal momenta

go outward.

2 Bhabha scattering

The Bhabha scattering is the process e+e ! e+e. At the tree level, it consistsof two diagrams, as shown in Figure 2. The minus sign before the t-channel diagram

comes from the exchange of two fermion eld operators when contracting with in and

out states. In fact, the s- and t-channel diagrams correspond to the following two ways

of contraction, respectively:

hp1p2j =A =A jk1k2i; hp1p2j =A =A jk1k2i: (14)

In the high energy limit, we can omit the mass of electrons, then the amplitude for the

whole scattering process is:

iM = (ie)2v(k2)

u(k1)is

u(p1)v(p2) u(p1)u(k1) itv(k2)v(p2)

: (15)

Where we have used the Mandelstam variables s, t and u. They are dened as:

s = (k1 + k2)2; t = (p1 k1)2; u = (p2 k1)2: (16)

In the massless case, k21 = k22 = p

21 = p

22 = 0, thus we have:

s = 2k1 k2 = 2p1 p2; t = 2p1 k1 = 2p2 k2; u = 2p2 k1 = 2p1 k2: (17)

We want to get the unpolarized cross section, thus we must average the ingoing spins

and sum over outgoing spins. That is:

1

4

Xspin

jMj2 = e4

4s2

Xv(k2)u(k1)u(p1)v(p2)2+

e4

4t2

Xu(p1)u(k1)v(k2)v(p2)2 e

4

4st

Xhv(p2)u(p1)u(k1)

v(k2)u(p1)u(k1)v(k2)v(p2) + c.c.

i=

e4

4s2tr (=k1

=k2) tr (=p2=p1) +

e4

4t2tr (=k1

=p1

) tr (=p2=k2)

e4

4st

htr (=k1

=k2=p2=p1) + c.c.

i=

2e4(u2 + t2)

s2+

2e4(u2 + s2)

t2+

4e4u2

st

= 2e4t2

s2+

s2

t2+ u2

1s+

1

t

2: (18)

3


In the center-of-mass frame, we have k01 = k02 k0, and k1 = k2, thus the total

energy E2CM = (k01 + k

02)2 = 4k2 = s. According to the formula for the cross section in

the four identical particles' case (Eq.4.85): dd

CM

=1

642ECM

14

XjMj2

; (19)

thus dd

CM

=2

2s

t2

s2+

s2

t2+ u2

1s+

1

t

2; (20)

where = e2=4 is the ne structure constant. We integrate this over the angle ' to

get: dd cos

CM

=2

s

t2

s2+

s2

t2+ u2

1s+

1

t

2: (21)

3 The spinor products (2)

In this problem we continue our study of spinor product method in last chapter. The

formulae needed in the following are:

uL(p) =1p

2p k0 =puR0; uR(p) =

1p2p k0 =

puL0: (22)

s(p1; p2) = uR(p1)uL(p2); t(p1; p2) = uL(p1)uR(p2): (23)

For detailed explanation for these relations, see Problem 3.3

(a) Firstly, let us prove the following relation:

js(p1; p2)j2 = 2p1 p2: (24)

We make use of the another two relations,

uL0uL0 =1 5

2=k0; uR0uR0

1 + 5

2=k0: (25)

which are direct consequences of the familiar spin-sum formulaP

u0u0 = =k0. We now

generalize this to:

uL(p)uL(p) =1 5

2=p; uR(p)uR(p) =

1 + 5

2=p: (26)

We prove the rst one:

uL(p)uL(p) =1

2p k0 =puR0uR0=p =1

2p k0 =p1 + 5

2=k0=p

=1

2p k01 5

2=p=k0=p =

1

2p k01 5

2(2p k =k0=p)=p

=1 5

2=p 1

2p k01 5

2=k0p

2 =1 5

2=p: (27)

The last equality holds because p is lightlike. Then we get:

js(p1; p2)j2 =juR(p1)uL(p2)j2 = truL(p2)uL(p2)uR(p1)uR(p1)

=1

4tr(1 5)=p2(1 5)=p1

= 2p1 p2: (28)

4


(b) Now we prove the relation:

tr (12 n) = tr (n 21); (29)

where i = 0; 1; 2; 3; 5.

To make things easier, let us perform the proof in Weyl representation, without loss

of generality. Then it's easy to check that

()T =

(

; = 0; 2; 5;

; = 1; 3: (30)

Then, we dene M = 13, and it can be easily shown that M1M = ()T , andM1M = 1. Then we have:

tr (12 n) = tr (M11MM12M M1nM)= tr

(1)T

2)T (n)T = tr (n 21)T

= tr (n 21): (31)

With this formula in hand, we can derive the equality

uL(p1)uL(p2) = uR(p2)

uR(p1) (32)

as follows

LHS = CuR0=p1=p2uR0 = C tr (=p1

=p2)

= C tr (=p2=p1) = CuL0=p2

=p1uL0 = RHS;

in which C 2p(p1 k0)(p2 k0)1.(c) The way of proving the Fierz identity

uL(p1)uL(p2)[]ab = 2

uL(p2)uL(p1) + uR(p1)uR(p2)

ab

(33)

has been indicated in the book. The right hand side of this identity, as a Dirac matrix,

which we denoted by M , can be written as a linear combination of 16 matrices listed

in Problem 3.6. In addition, it is easy to check directly that M = M5. Thus Mmust have the form

M = 1 5

2

V

+ 1 + 5

2

W

:

Each of the coecients V and W can be determined by projecting out the other one

with the aid of trace technology, that is,

V =1

2tr

1 5

2

M= uL(p1)

uL(p2); (34)

W =1

2tr

1 + 5

2

M= uR(p2)

uR(p1) = uL(p1)uL(p2): (35)

The last equality follows from (32). Substituting V and W back, we nally get the

left hand side of the Fierz identity, which nishes the proof.

5


(d) The amplitude for the process at leading order in is given by

iM = (ie2)uR(k2)uR(k1) isvR(p1)vR(p2): (36)

To make use of the Fierz identity, we multiply (33), with the momenta variables changed

to p1 ! k1 and p2 ! k2, byvR(p1)

aand

vR(p2)

b, and also take account of (32),

which leads to

uR(k2)uR(k1)vR(p1)vR(p2)

= 2vR(p1)uL(k2)uL(k1)vR(p2) + vR(p1)uR(k1)uR(k2)vR(p2)

= 2s(p1; k2)t(k1; p2): (37)

Then,

jiMj2 = 4e4

s2js(p1; k2)j2jt(k1; p2)j2 = 16e

4

s2(p1 k2)(k1 p2) = e4(1 + cos )2; (38)

andd

d

(e+Le

R ! +LR) =

jiMj2642Ecm

=2

4Ecm(1 + cos )2: (39)

It is straightforward to work out the dierential cross section for other polarized pro-

cesses in similar ways. For instance,

d

d

(e+Le

R ! +RL ) =

e4jt(p1; k1)j2js(k2; p2)j2642Ecm

=2

4Ecm(1 cos )2: (40)

(e) Now we recalculate the Bhabha scattering studied in Problem 5.2, by evaluating

all the polarized amplitudes. For instance,

iM(e+LeR ! e+LeR)

= (ie)2uR(k2)

uR(k1)isvR(p1)vR(p2)

uR(p1)uR(k1) itvR(k2)vR(p2)

= 2ie2

s(p1; k2)t(k1; p2)

s s(k2; p1)t(k1; p2)

t

: (41)

Similarly,

iM(e+LeR ! e+ReL ) = 2ie2t(p1; k1)s(k2; p2)

s; (42)

iM(e+ReL ! e+LeR) = 2ie2s(p1; k1)t(k2; p2)

s; (43)

iM(e+ReL ! e+ReL ) = 2ie2t(p1; k2)s(k1; p2)

s t(k2; p1)s(k1; p2)

t

; (44)

iM(e+ReR ! e+ReR) = 2ie2t(k2; k1)s(p1; p2)

t; (45)

iM(e+LeL ! e+LeL ) = 2ie2s(k2; k1)t(p1; p2)

t: (46)

Squaring the amplitudes and including the kinematic factors, we nd the polarized

dierential cross sections as

d

d

(e+Le

R ! e+LeR) =

d

d

(e+Re

L ! e+ReL ) =

2u2

2s

1s+

1

t

2; (47)

6


d

d

(e+Le

R ! e+ReL ) =

d

d

(e+Re

L ! e+LeR) =

2

2s

t2

s2; (48)

d

d

(e+Re

R ! e+ReR) =

d

d

(e+Le

L ! e+LeL ) =

2

2s

s2

t2: (49)

Therefore we recover the result obtained in Problem 5.2:

d

d

(e+e ! e+e) =

2

2s

t2

s2+

s2

t2+ u2

1s+

1

t

2: (50)

4 Positronium lifetimes

In this problem we study the decay of positronium (Ps) in its S and P states. To

begin with, we recall the formalism developed in the Peskin & Schroeder that treats

the problem of bound states with nonrelativistic quantum mechanics. The positronium

state jPsi, as a bound state of an electron-positron pair, can be represented in terms ofelectron and positron's state vectors, as

jPsi =p2MP

Zd3k

(2)3 (k)Cab

1p2m

jea (k)i1p2m

je+b (k)i; (51)

where m is the electron's mass, MP is the mass of the positronium, which can be taken

to be 2m as a good approximation, a and b are spin labels, the coecient Cab depends

on the spin conguration of jPsi, and (k) is the momentum space wave function forthe positronium in nonrelativistic quantum mechanics. In real space, we have

100(r) =

r(mr)

3

exp(mrr); (52)

21i(r) =

r(mr=2)

5

xi exp(mrr=2): (53)

where mr = m=2 is the reduced mass. Then the amplitude of the decay process Ps! 2can be represented in terms of the amplitude for the process e+e ! 2 as

M(Ps! 2) = 1pm

Zd3k

(2)3 (k)CabcMea (k)e+b (k)! 2: (54)

We put a hat on the amplitude of e+e ! 2. In the following, a hatted amplitudealways refer to this process.

(a) In this part we study the decay of the S-state positronium. As stated above, we

have to know the amplitude of the process e+e ! 2, which is illustrated in Figure 3with the B replaced with , and is given by

icM = (ie)2(p1)(p2) v(k2)

i(=k1 =p1 +m)(k1 p1)2 m2

+ i(=k1 =p2 +m)(k1 p2)2 m2

u(k1); (55)

where the spinors can be written in terms of two-component spinors and 0 in thechiral representation as

u(k1) =

pk1 pk1

; v(k2) =

pk2 0

pk2 0: (56)

7


We also write as:

=

0

0

!;

where = (1; i) and = (1;i) with i the three Pauli matrices. Then theamplitude can be brought into the following form,

icM = ie2(p1)(p2)0y t(k1 p1)2 m2 + u

(k1 p2)2 m2; (57)

with

t =p

k2 pk1

pk2

pk1

m

+p

k2 pk1

pk2

pk1

(k1 p1);

u =p

k2 pk1

pk2

pk1

m

+p

k2 pk1

pk2

pk1

(k1 p2):

In the rest of the part (a), we take the nonrelativistic limit, with the momenta chosen

to be

k1 = k2 = (m; 0; 0; 0); p

1 = (m; 0; 0;m); p

2 = (m; 0; 0;m): (58)

Accordingly, we can assign the polarization vectors for nal photons to be

(p1) =1p2(0; 1;i; 0); (p2) = 1p2 (0;1;i; 0): (59)

Now substituting the momenta (58) into (57), noticing thatpki =

pki =

pm (i =

1; 2), and (k1 p1)2 = (k1 p2)2 = m2, and also using the trick that one can freelymake the substitution ! since the temporal component of the polarizationvectors always vanishes, we get a much more simplied expression,

icM = ie2(p1)(p2)0y3 3: (60)The positronium can lie in spin-0 (singlet) state or spin-1 (triplit) state. In the former

case, we specify the polarizations of nal photons in all possible ways, and also make

the substitution 0y ! 1p2(See (5.49) of Peskin & Schroeder), which leads to

icMs++ = icMs = i2p2e2; icMs+ = icMs+ = 0; (61)where the subscripts denote nal photons' polarizations and s means singlet. We show

the mid-step for calculating iMs++ as an example:

icMs++ = ie22 tr h(1 + i2)3(1 + i2) (1 + i2)3(1 + i2)0yi = i2p2e2:In the same way, we can calculate the case of triplet initial state. This time, we make

the substitution 0y ! n =p2, with n = (x^ iy^)=p2 or n = z^, corresponding tothree independent polarizations. But it is straightforward to show that the amplitudes

with these initial polarizations all vanish, which is consistent with our earlier results by

using symmetry arguments in Problem 3.8.

8


Therefore it is enough to consider the singlet state only. The amplitude for the decay

of a positronium in its 1S0 state into 2 then follows directly from (54), as

M(1S0 ! 2) = (x = 0)pm

cMs; (62)where (x = 0) =

p(m=2)3= according to (52). Then the squared amplitude with

nal photons' polarizations summed isXspin

M(1S0 ! 2)2 = j (0)j22m

jMs++j2 + jMsj2

= 165m2: (63)

Finally we nd the decay width of the process Ps(1S0)! 2, to be

(1S0 ! 2) = 12

1

4m

Zd3p1d

3p2(2)62E12E2

PM(1S0 ! 2)2(2)4(4)(pPs p1 p2)=

1

2

1

4m

Zd3p1

(2)34m2PM(1S0 ! 2)2(2)(m E1)

=1

25m; (64)

where an additional factor of 1=2 follows from the fact that the two photons in the nal

state are identical particles.

(b) To study the decay of P state (l = 1) positronium, we should keep one power of

3-momenta of initial electron and positron. Thus we set the momenta of initial and nal

particles, and also the polarization vectors of the latter, in ee+ ! 2, to be

k1 = (E; 0; 0; k); k2 = (E; 0; 0;k);

p1 = (E;E sin ; 0; E cos ); p2 = (E;E sin ; 0;E cos );

(p1) =1p2(0; cos ;i; sin ); (p2) = 1p2 (0; cos ;i; sin ): (65)

Here we have the approximate expression up to linear order in k:pk1 =

pk2 =

pm k

2pm3 +O(k2);p

k2 =pk1 =

pm+

k

2pm3 +O(k2);

1

(k1 p1)2 m2 = 1

2m2 k cos

2m3+O(k2);

1

(k1 p2)2 m2 = 1

2m2+

k cos

2m3+O(k2):

Consequently,

t = 2m2(1s +

3c) mk3 + 3 + 23+O(k2);

u = 2m2(1s + 3c) mk3 + 3 + 23

+O(k2);

where we use the shorthand notation s = sin and c = cos . We can use these

expansion to nd the terms in the amplitude icM of linear order in k, to beicMO(k) = ie2(p1)(p2) k2m0yh 2c(1s + 3c) 2c(1s + 3c)

9


+3 + 3 + 23

+3 + 3 + 23

i;

(66)

Feeding in the polarization vectors of photons, and also make the substitution 0y !n =p2 or 1=p2 for triplet and singlet positronium, respectively, as done in last part,we get

icM##jO(k) = 0; icM##jO(k) = i2s(1 + c)e2k=m;icM"#+#" jO(k) = i2p2e2k=m; icM"#+#" jO(k) = i2p2s2e2k=m;icM""jO(k) = 0; icM""jO(k) = i2s(1 + c)e2k=m;icM"##" jO(k) = 0; icM"##" jO(k) = 0: (67)

The vanishing results in the last line indicate that S = 0 state of P -wave positronium

cannot decay to two photons.

(c) Now we prove that the state

jB(k)i =p2MP

Zd3p

(2)3 i(p)a

yp+k=2

ibyp+k=2j0i (68)

is a properly normalized state for the P -wave positronium. In fact,

hB(k)jB(k)i = 2MPZ

d3p0

(2)3d3p

(2)3 j (p

0) i(p)

h0jbp0+k=2jyap0+k=2ayp+k=2ibyp0+k=2j0i

= 2MP

Zd3p0

(2)3d3p

(2)3 j (p

0) i(p)

h0jbp0+k=2jyibyp0+k=2j0i(2)3(3)(p0 p)

= 2MP

Zd3p

(2)3 j (p) i(p)h0jbp+k=2jyibyp+k=2j0i

= 2MP

Zd3p

(2)3 j (p) i(p)h0j tr (jyi)j0i(2)3(3)(0)

= 2MP (2)3(3)(0); (69)

which is precisely the needed normalization of a state. In this calculation we have used

the anticommutation relations of creation and annihilation operators, as well as the

normalization of the wave function and the matrices.

(d) Now we evaluate the partial decay rate of the S = 1 P -wave positronium of

denite J into two photons. The states for the positronium is presented in (c), with the

matrices chosen as

=

8>>>>>>>>>>>:

1p6i; J = 0;

1

2ijknjk; J = 1;

1p3hijj ; J = 2;

(70)

and the wave function given by (53).

10


Firstly, consider the J = 0 state, in which case we have

iM(3P0 ! ) = 1pm

Zd3k

(2)3 i(k)

1p6iabicMea (k)e+b (k)! ; (71)

where ; = + or are labels of photons' polarizations and a; b =" or # are spinorindices. For amplitude icM, we only need the terms linear in k, as listed in (67). Let usrewrite this as

icMea (k)e+b (k)! = F ab;iki:In the same way, the wave function can also be put into the form of i(x) = x

if(r),

with r = jxj. Then the integration above can be carried out to be

iM(3P0 ! ) = ip6m

iabFab;j

i@

@xj i(x)

x=0

=ip6m

iabFab;if(0): (72)

On the other hand, we have chose the direction of k to be in the x3-axis, then F ab;1 =

F ab;2 = 0 as a consequence. Therefore,

iM(3P0 ! ) = ip6m

f(0)F "";3 F ##;3

=

r7

24m sin : (73)

Square these amplitudes, sum over the photons' polarizations, and nish the phase space

integration in the same way as what we did in (a), we nally get the partial decay rate

of the J = 0 P -wave positronium into two photons to be

(3P0) =1

5767m: (74)

The positronium in 3P1 state, namely the case J = 1, cannot decay into two photons

by the conservation of the angular momentum, since the total angular momentum of

two physical photons cannot be 1. Therefore let us turn to the case of J = 2. In this

case we should average over the initial polarizations of the positronium, which can be

represented by the symmetric and traceless polarization tensors hijn , with n = 1; 2; ; 5the labeled of 5 independent polarizations. Let us choose these tensors to be

hij1 =1p2(i2j3 + i3j2); hij2 =

1p2(i1j3 + i3j1);

hij3 =1p2(i1j2 + i2j1); hij4 =

1p2(i1j1 i2j2);

hij5 =1p2(i1j1 i3j3): (75)

Then the decay amplitude for a specic polarization of J = 2 Ps can be represented as

iMn(3P2 ! ) = 1pm

Zd3k

(2)3 i(k)

1p3hijn

jabicMea (k)e+b (k)!

=1p3m

hijn jabF

ab;if(0): (76)

Now substituting all stus in, we nd the nonvanishing components of the decay ampli-

tude to be

iM2(3P2 ! ) =r

7

48im;

11


iM2(3P2 ! ) =r

7

48im sin2 ;

iM5(3P2 ! ) = 2r

7

48im sin2 : (77)

Squaring these amplitudes, summing over photon's polarizations and averaging the ini-

tial polarization of the positronium (by dividing the squared and summed amplitude by

5), we get

1

5

Xspin

Mn(3P2 ! 2)2 = 7m2120

(1 + sin2 + 4 sin4 ): (78)

Finally, we nish the phase space integration and get the partial decay rate of 3P2positronium into 2 photons to be

(3P2) =19

192007m: (79)

5 Physics of a massive vector boson

In this problem, the mass of electron is always set to zero.

(a) We rstly compute the cross section (e+e ! B) and the decay rate (B !e+e). For the cross section, the squared amplitude can be easily found to be

1

4

Xspin

jiMj2 = 14

Xspin

ig(i) v(p0)u(p)2 = 2g2(p p0): (80)Note that we have set the mass of electrons to be zero. Then the cross section can be

deduced from (4.79). Let's take the initial momenta to be

p = 12 (E; 0; 0; E); p0 = 12 (E; 0; 0;E); (81)

with E being the center-of-mass energy. Then it's easy to get

=g2

4E(2)(MB E) = g

2

4E(2)2MB(M

2B s) = g2(M2B s); (82)

where s = E2.

To deduce the decay rate, we should average polarizations of massive vector B instead

of two electrons. Thus the squared amplitude in this case reads

1

3

Xspin

jiMj2 = 83g2(p p0): (83)

The decay rate can be found from (4.86):

=1

2MB

Zd3p

(2)3d3p0

231

2Ep

1

2Ep0

13

PjMj2(2)4(4)(pB p p0)=

1

2MB

Zd3p

(2)31

4E2p

163 g

2E2p

(2)(MB 2Ep)

=4

(2)22MB

Zdp 43 g

2E2p12 (

12 MB Ep) =

g2MB12

: (84)

We see the cross section and the decay rate satisfy the following relation, as expected:

(e+e ! B) = 122

MB(B ! e+e)(sM2): (85)

12


k1

p2p1

k2e e+

B

+

k1

p1

k2

p2

e e+

B

Figure 3: The tree diagrams of the process ee+ ! +B. All initial momenta go inward andall nal momenta go outward.

(b) Now we calculate the cross section (ee+ ! +B) in COM frame. The relateddiagrams are shown in Figure 3. The amplitude reads:

iM = (ie)(ig)(p1)e(p2)v(k2)h

i

=k1 =p1

+

i

=k1 =p2

iu(k1); (86)

where is the polarization of photon while e is the polarization for B. Now we square

this amplitude,

1

4

Xspin

jiMj2 = 14e2g2gg tr

(=k1 =p1)t

+

(=p1 =k2)

u

=k1

(=k1 =p1)

t+

(=p1 =k2)u

=k2

= 8e2g2

(k1 p1)(k2 p1)

t2+

(k1 p1)(k2 p1)u2

+2(k1 k2)(k1 k2 k1 p1 k2 p1)

tu

= 2e2g2

u

t+

t

u+

2s(s+ t+ u)

tu

= 2e2g2

u

t+

t

u+

2sM2Btu

(87)

Then the cross section can be evaluated as dd

CM

=1

2Ek12Ek2 jvk1 vk2 jjp1j

(2)24ECM

14

PjMj2=

e2g2

322s

1 M

2B

s

ut+

t

u+

2sM2Btu

: (88)

We can also write this dierential cross section in terms of squared COM energy s and

scattering angle . To do this, we note that

s = E2CM; t = (M2B E2CM) sin2 2 ; u = (M2B E2CM) cos2 2 : (89)

Then we have dd

CM

=e2g2(1M2B=s)162s sin2

1 + cos2 +

4sM2B(sM2B)2

; (90)

and dd cos

CM

=g2(1M2B=s)

2s sin2

1 + cos2 +

4sM2B(sM2B)2

: (91)

13


(c) The dierential cross obtained in (b) diverges when ! 0 or ! . Now let usstudy the former case, namely ! 0.

If we cut of the integral from 2c ' m2e=s, then we have:Zc

dd cos

CM

sin d ' g2(1M2B=s)

2s

2 +

4sM2B(sM2B)2

Z 1m2e=s dt1 t2

' g2(1M2B=s)

4s

2 +

4sM2B(sM2B)2

log

sm2e

=

g2

2

1 +M4B=s2

sM2Blog

sm2e

(92)

Now we calculate the following expression:Z 10

dx f(x)(e+e ! B)ECM=(1x)s

=

Z 10

dx

2

1 + (1 x)2x

log sm2e

g2

M2B (1 x)s

=

g2

2

1 +M4B=s2

sM2Blog

sm2e

(93)

6 The spinor products (3)

This problem generalize the spinor product formalism to the processes involving

external photons.

(a) Firstly we can represent photon's polarization vectors in terms of spinors of denite

helicity. Let the momentum of the photon be k, and p be a lightlike momentum such

that p k 6= 0. Then, the polarization vector (k) of the photon can be taken to be

+(k) =1p4p k uR(k)

uR(p); (k) =

1p4p k uL(k)

uL(p); (94)

where the spinors uL;R(k) have been introduced in Problems 3.3 and 5.3. Now we use

this choice to calculate the polarization sum:

+(k)+ (k) +

(k)

(k)

=1

4p khuR(k)

uR(p)uR(p)uR(k) + uL(k)

uL(p)uL(p)uL(k)

i=

1

4p k tr=p

=k= g + p

k + pk

p k : (95)

When dotted into an amplitude with external photon, the second term of the result

vanishes. This justies the denitions above for photon's polarization vectors.

(b) Now we apply the formalism to the process e+e ! 2 in the massless limit. Therelevant diagrams are similar to those in Figure 3, except that one should replace the

label `B' by `'. To simplify expressions, we introduce the standard shorthand notations

as follows:

pi = uR(p); p] = uL(p); hp = uL(p); [p = uR(p): (96)

14


Then the spin products become s(p1; p2) = [p1p2] and t(p1; p2) = hp1p2i. Variousexpressions get simplied with this notation. For example, the Fierz identity (37) now

reads [k2k1i[p1p2i = 2[p1k2]hk1p2i. Similarly, we also have hk1k2]hp1p2] =

2hk1p1i[p2k2].Now we write down the expression for tree amplitude of e+Re

L ! RL. For illustra-

tion, we still keep the original expression as well as all explicit mid-steps. The auxiliary

lightlike momenta used in the polarization vectors are arbitrarily chosen such that the

calculation can be mostly simplied.

iM(e+ReL ! LR)

= (ie)2(p1)+(p2)uL(k2)

i

=k1 =p1

+

i

=k1 =p1

uL(k1)

= ie2 hk2p1][k1p2i4p(k2 p1)(k1 p2)

hk2(=k1 =p1)k1]t

+hk2(=k1 =p2)k1]

u

= ie2 hk2p1][k1p2i

2u

hk2k1]hk1k1] hk2p1]hp1k1]t

+hk2k1]hk1k1] hk2p2]hp2k1]

u

=2ie2u

hk1k2i[p1k1]hk2p2i[k1k1] hk2p1i[k1p1]hk2p2i[k1p1]t

+hk2k2i[k1p1]hk1p2i[k1k1] hk2k2i[p2p1]hp2p2i[k1k1]

u

= 2ie2

hk2p1i[k1p1]hk2p2i[k1p1]tu

; (97)

where we have used the spin sum identity =p = p]hp+ pi[p in the third equality, and alsothe Fierz transformations. Note that all spinor products like hppi and [pp], or hpkiand [pk] vanish. Square this amplitude, we getiM(e+ReL ! LR)2 = 4e4 tu : (98)In the same way, we calculate other polarized amplitudes:

iM(e+ReL ! RL)

= ie2 [k1p1ihk2p2]4p(k1 p1)(k2 p2)

[k2

(=k1 =p1)k1it

+[k2

(=k1 =p2)k1iu

= 2ie2

hk2p1i[k1p2]hk2p2i[k1p2]tu

(99)

Note that we have used a dierent set of auxiliary momenta in photons' polarizations.

After evaluating the rest two nonvanishing amplitudes, we get the squared polarized

amplitudes, as follows:M(e+ReL ! LR)2 = M(e+LeR ! RL)2 = 4e4 tu ; (100)M(e+LeR ! LR)2 = M(e+LeR ! LR)2 = 4e4 ut : (101)Then the dierential cross section follows straightforwardly,

d

d cos =

1

16s

1

4

Xspin

jiMj2=

22

s

tu+

u

t

; (102)

which is in accordance with (5.107) of Peskin & Schroeder.

15


Chapter 6

Zhong-Zhi Xianyu




1 Rosenbluth formula

In this problem we derive the dierential cross section for the electron-proton scatter-

ing in the lab frame, assuming that the scattering energy is much higher than electron's

mass, and taking account of the form factors of the proton. The result is known as

Rosenbluth formula. The relevant diagram is shown in Figure 1. Let us rstly work out

k1 k2

p1 p2

e p

e p

Figure 1: The electron-proton scattering. The blob denotes form factors that includes the

eect of strong interaction. All initial momenta go inward and all nal momenta go outward.

the kinematics. In the lab frame, the momenta can be parameterized as

k1 = (E; 0; 0; E); p1 = (E0; E0 sin ; 0; E0 cos ); k2 = (M; 0; 0; 0); (1)

and p2 can be found by momentum conservation, k1 + k2 = p1 + p2. With the on-shell

condition p22 =M2, we nd that

E0 =ME

M + 2E sin2 2: (2)

We also use q = k1 p1 to denote the momentum transfer and t = q2 its square. Notethat we have set the electron mass to zero.

Now we write down the amplitudeM.

iM = (ie)2 U(p2)

F1(q

2) +iq2M

F2(q2)

U(k2)

itu(p1)u(k1); (3)


1


where U is the spinor for the proton and u is for the electron, M is the mass of the

proton. At this stage, we convert this expression into a more convenient form by means

of the Gordon identity (see Problem 3.2):

iM = (ie)2 U(p2)

(F1 + F2) (p2 + k2)

2MF2

U(k2)

itu(p1)u(k1); (4)

Now, the modular-squared amplitude with initial spins averaged and nal spins

summed, is

1

4

XjMj2 = e

4

4q4tr

(F1 + F2) (p2 + k2)

2MF2

(=k2 +M)

(F1 + F2) (p2 + k2)

2MF2

(=p2 +M)

tr

=k1=p1

=

4e4M2

q4

2E2 + 2E02 + q2

(F1 + F2)

2

2F1F2 + F

22

1 + q

2

4M2

(E + E0)2 + q2

1 q24M2

: (5)

There are two terms in the square bracket in the last expression. We rewrite the rst

factor in the second term as

2F1F2 + F22

1 + q

2

4M2

= (F1 + F2)

2 F 21 + q2

4M2F22 ;

and combine the (F1 + F2)2 part into the rst term, which leads to

1

4

XjMj2 = 4e

4M2

q4

q4

2M2 (F1 + F2)2 + 4

F 21 q

2

4M2F22

EE0 cos2 2

;

where we have used the following two relations which can be easily justied:

E0 E = q22m (6)q2 = 4E0E sin2 2 : (7)

Now we can put the squared amplitude into its nal form:

1

4

XjMj2 = 16e

4E2M3

q4M + 2E sin2 2

F 21 q

2

4M2F22

cos2 2

q2

2M2(F1 + F2)

2 sin2 2

; (8)

On the other hand, we can derive the A+B ! 1+ 2 dierential cross section in thelab frame as

dL =1

2EA2EB jvA vB jZ

d3p1d3p2

(2)62E12E2jMj2(2)4(4)(p1 + p2 pA pB): (9)

In our case, EA = E, EB =M , E1 = E0, and jvA vB j ' 1, thus

dL =1

4EM

Zd3p1d

3p2(2)62E12E2

jMj2(2)4(4)(p1 + p2 pA pB)

=1

4EM

ZE02dE0d cos d'(2)32E02E2

jMj2(2)E0 + E2(E0) E M=

1

4EM

ZE02dE0d cos d'(2)22E02E2

jMj21 +

E0 E cos E2(E0)

1

E0 ME

M + 2E sin2 2

2


=1

4EM

Zd cos

8jMj2 E

0

M + 2E sin2 2

where we use the notation E2 = E2(E0) to emphasize that E2 is a function of E0. That

is,

E2 =pM2 + E2 + E02 2E0E cos :

Then, dd cos

L=

1

32M + 2E sin2 2

2 jMj2 (10)So nally we get the dierential cross section, the Rosenbluth formula, as d

d cos

L=

2

2E21 + 2EM sin

2 2

sin4 2

F 21 q

2

4M2F22

cos2 2

q2

2M2(F1 + F2)

2 sin2 2

: (11)

2 Equivalent photon approximation

In this problem we study the scattering of a very high energy electron from a target

in the forward scattering limit. The relevant matrix element is

M = (ie)u(p0)u(p) igq2

cM(q): (12)(a) First, the spinor product in the expression above can be expanded as

u(p0)u(p) = A q +B q + C 1 +D 2 : (13)

Now, using the fact that qu(p0)u(p) = 0, we have

0 = Aq2 +Bq ~q ' 4AEE0 sin2 2 +Bq ~q ) B 2:

(b) It is easy to nd that

1 = N(0; p0 cos p; 0;p0 sin ); 2 = (0; 0; 1; 0);

where N = (E2 + E02 2EE0 cos )1=2 is the normalization constant. Then, for theright-handed electron with spinor u+(p) =

p2E(0; 0; 1; 0)T and left-handed electron

with u(p) =p2E(0; 1; 0; 0)T , it is straightforward to show that

u+(p0) =

p2E0(0; 0; cos 2 ; sin

2 )

T ; u(p0) =p2E0( sin 2 ; cos 2 ; 0; 0); (14)

and,

u(p0) 1u(p) ' pEE0

E + E0

jE E0j; (15)

u(p0) 2u(p) ' ipEE0; (16)

u(p0) 1u(p) = u(p0) 2u(p) = 0: (17)

That is to say, we have

C = pEE0

E + E0

jE E0j; D = ipEE0: (18)

3


(c) The squared amplitude is given by

jMj2 = e2

(q2)2cM(q)cM(q)C1 +D2 C1 +D2 (19)

Averaging and summing over the initial and nal spins of the electron respectively, we

get

1

2

XjMj2 = e

2

2(q2)2cM(q)cM(q)jC+j2 + jCj21 1 + jD+j2 + jDj22 2

+C+D

+ + CD

1

2 +

C+D+ + C

D

2

1

=

e2

(q2)2cM(q)cM(q)EE02 E + E0

E E021

1 +

2 2

(20)

Then the cross section readsZd =

1

2E2Mt

Zd3p0

(2)32E0d3pt

(23)2Et

1

2

XjMj2

(2)4(4)

Ppi

=e2

2E2Mt

Zd3p0

(2)32E0EE02

3(q2)2

E + E0E E0

2+ 1

Z

d3pt(23)2Et

jcM(q)j2(2)4(4)Ppi= 1

2E2Mt

2

Zdxh1 +

2 xx

2i Z 0

d2 sin

4(1 cos )2

Z

d3pt(23)2Et

jcM(q)j2(2)4(4)Ppi: (21)where we have used the trick described in the nal project of Part I (radiation of gluon

jets) to separate the contractions of Lorentz indices, and x (E E0)=E. Now letus focus on the integral over the scattering angle in the last expression, which is

contributed from the following factor:Z 0

d2 sin

4(1 cos )4 Z0

d

which is logarithmically divergent as ! 0.

(d) We reintroduce the mass of the electron into the denominator to cut o the diver-

gence, namely, let q2 = 2(EE0 pp0 cos ) + 2m2. Then we can expand q2, treatingm2 and as small quantities, as

q2 ' (1 x)E22 x2

1 xm2:

Then the polar angle integration near = 0 becomesZ0

d 32 +

x2

(1 x)2m2

E2

2 1

2log

E2

m2: (22)

(e) Combining the results above, the cross section can be expressed asZd = 1

2E2Mt

2

Zdxh1 +

2 xx

2i Z0

d 32 +

x2

(1 x)2m2

E2

24


Z

d3pt(23)2Et

jcM(q)j2(2)4(4)Ppi=

1

2E2Mt

2

Zdx

1 + (1 x)2x2

logE2

m2

Z

d3pt(23)2Et

jcM(q)j2(2)4(4)Ppi: (23)3 Exotic contributions to g 2(a) The 1-loop vertex correction from Higgs boson is

u(p0)u(p) = ip

2

2 Z ddk(2)d

i

(k p)2 m2hu(p0)

i

=k + =q m i

=k mu(p)

=i2

2

Z 10

dx

Z 1x0

dy

Zddk0

(2)d2u(p0)Nu(p)(k02 )3 ; (24)

with

N = (=k + =q +m)(=k +m);

k0 = k xp+ yq; = (1 x)m2 + xm2h x(1 x)p2 y(1 y)q2 + 2xyp q:

To put this correction into the following form:

= F1(q) +iq2m

F2(q); (25)

we rst rewrite N as

N = A +B(p0 + p) + C(p0 p);

where term proportional to (p0 p) can be thrown away by Ward identity q(q) = 0.This can be done by gamma matrix calculations, leading to the following result:

N =h

2d 1

k02 + (3 + 2x x2)m2 + (y xy y2)q2

i

+ (x2 1)m(p0 + p):

Then, use Gordon identity, we nd

N =h

2d 1

k02 + (x+ 1)2m2 + (y y2 xy)q2

i

+

i

2m 2m2(1 x2):

Comparing this with (25), we see that

F2(q = 0) = 2i2m2

Z 10

dx

Z 1x0

dy

Zd4k

Date post:	04-Oct-2015
Category:	Documents
Upload:	jerad-williams
View:	111 times
Download:	0 times

Peskin&Schroeder Solution Manual

Documents