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Relativistic Quantum Fields 1 Mark Hindmarsh University of Sussex Autumn Term 2001 Tuesday 10:15am Arundel 1C Thursday 12:30pm Arundel 1C PostScript and PDF versions of these notes available: http://www.pact.cpes.susx.ac.uk/users/markh/RQF1/rqf1.ps http://www.pact.cpes.susx.ac.uk/users/markh/RQF1/rqf1.pdf Contents 0 Course information 3 1 Preliminaries 7 2 Relativistic wave equations 7 2.1 Special Relativity in relativistic notation .............. 7 2.1.1 Lorentz transformations and Space-time interval ...... 7 2.1.2 Relativistic notation ..................... 8 2.1.3 Matrix representation of Lorentz transformations ..... 8 2.1.4 Space-time metric ....................... 9 2.1.5 General Lorentz transformations: the Lorentz group .... 10 2.1.6 Derivatives .......................... 10 2.1.7 4-vectors and the scalar product ............... 11 2.1.8 The d’Alembertian ...................... 12 2.1.9 Lorentz covariance ...................... 12 2.2 Klein-Gordon equation ........................ 13 2.3 Maxwell’s equations in covariant form ................ 14 2.3.1 Gauges ............................. 17 1

# Relativistic Quantum Fields 1

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is moving at velocity v in the x direction relativeto another F. Special Relativity tells us that coordinates measured in F

arerelated to coordinates measured in F by a Lorentz transformationt t

= (t vx/c2) ,x x

= (x vt) ,y y

= y,z z

= z,(2.1)where = 1/

1 v2/c2. (2.2)Although neither the distance nor the time interval between two events is observer-independent in special relativity, there is a concept of space-time interval, whichcontains a little of both, and is something that all observers can agree upon.In its innitesimal form, the interval ds between two events at (t, x, y, z) and(t + dt, x + dx, y + dy, z + dz) is given byds2= c2dt2dx2dy2dz2. (2.3)One can easily check that the value of ds does not change under this transforma-tion. We say that ds is Lorentz invariant.72.1.2 Relativistic notationIf a theory which is usually formulated in terms of a set of equations is con-sistent with the Principle of Special Relativity we often say that it is Lorentzcovariant. There is a notation which, if followed correctly, automatically ensuresthat equations are Lorentz covariant. Quantities are assembled into 4-vectors(and later on we will encounter 4-tensors) which transform in a simple linearway when one compares their values between observers in dierent states of uni-form motion. Just as one can form a 3-component vector xifrom three spatialcoordinatesxi= (x1, x2, x3) (x, y, z), (2.4)one can dene the 4-vector space-time coordinate for an event byx= (x0, x1, x2, x3) (ct, x, y, z). (2.5)We shall adopt the following conventions when labelling vectors: Greek letters from mid-alphabet for space-time indices (, , , , . . .); Roman letters from mid-alphabet for spatial indices (i, j, k, . . .); bold face will also be used for 3-vectors, e.g. x.2.1.3 Matrix representation of Lorentz transformationsRecall the form of a Lorentz transformation between two frames of referencemoving with velocity v in the x direction relative to one another (2.1). Thislinear transformation can be represented by matrices acting linearly on thecoordinates:xx=3=0x, (2.6)or in dierential formdxdx=3=0dx, (2.7)The matrix has entries =

v/c v/c 1 1

, (2.8)where labels the rows and the columns.In order to save writing a large number of summation signs, we often useEinsteins convention that repeated indices in an expression are summed over au-tomatically without the need for a summation sign. In this convention, Equation(2.7 becomesdxdx= dx, (2.9)82.1.4 Space-time metricIt is convenient in special relativity (and fundamental in general relativity!) todene a matrix g, which is used in the expression for the space-time interval:ds2=3,=0gdxdx. (2.10)We call g the metric. In Special Relativity, the metric always takes a particularconstant value,g =

1 1 1 1

. (2.11)We often call the Minkowski metric. In Einsteins repeated index convemtion,(2.10) can be written asds2= dxdx. (2.12)Note that the position of the indices is important. There is a dierence betweenvectors with superscript indices and those with subscript indices, which shows upin their properties under Lorentz transformations. We may use the metric tensorto dene an innitesimal coordinate with a lower index:dx = dx(2.13)Vectors with subscript indices are termed covariant, to distinguish them fromtheir contravariant partners with superscript indices. They transform oppositelyto covariant 4-vectors, with the inverse of the Lorentz transformation matrix :dx dx

= (1)dx. (2.14)One can check that this is true by noting that the space-time interval can bewritten ds2= dxdx, and then substituting the transformation laws (2.7) and(2.13). One ndsds2ds

2= dx

dx

= (1)dxdx. (2.15)Note that we have not just blindly substituted (2.7): we have the repeated indexfrom a to a . The meaning is still the same: that index is to be summed overthe values 0,1,2,3. However, if we have left the index as equation (2.15) wouldbe ambiguous, as we would not know how to pair o indices in the summations.This is an important rule with index notation: never use repeated indices twiceon the same side of an equation.Continuing with equation (2.15), we note that the indices are paired andcan be summed over. We can see that we are multiplying a matrix by its inverse1and therefore must obtain the identity, which expressed in index notation is(1) = . (2.16)9Here we introduce the Kronecker delta, dened by =

1, = ,0, = . (2.17)Note that the spatial components of a 4-vector with its index lowered (or acovariant vector) have the opposite sign to its counterpart with a raised index (acontravariant vector):x0 = x0, xi = xi. (2.18)Lastly, one can also dene the metric tensor with raised indices as the matrixinverse of the covariant metric tensor= , (2.19)2.1.5 General Lorentz transformations: the Lorentz groupOne can explicitly verify that the transformation law (2.1) leaves the space-timeinterval ds invariant. By choosing space coordinates so that the relative velocityof two inertial frames is along the x direction, it follows that all Lorentz trans-formations leave the interval invariant. Let us see what we can infer about thematrices from this condition. The interval transforms asds2ds2= dxdx= ds2. (2.20)But ds2may be written ds2= dxdx: thus we may infer that = . (2.21)Hence, any matrix which leaves the metric invariant under the transforma-tion (2.21) represents a Lorentz transformation. These matrices form a group oftransformations known as the Lorentz group. When combined with translationsymmetry, x x

= x+ a, with aa constant 4-vector, it forms a largergroup known as the Poincare group.12.1.6 DerivativesIt is through studying derivatives that we discover the utility of the idea of co-variant 4-vectors. Coordinate vectors are naturally contravariant, and it seemsslightly perverse to introduce a covariant version with lowered indices. How-ever, it turns out that there are 4-vectors which are naturally covariant, and theprincipal among these is the partial derivative 4-vector.1A more detailed exploration of the properties of the Poincare group, which is important forunderstanding the concepts of mass and spin in quantum eld theory, can be found in Ryder,Chapter 2, or Peskin and Schroeder, Chapter 3.10The standard partial derivatives with respect to spatial coordinates can becollected together into a 3-vector:i = xi =

x1, x2, x3

. (2.22)The i are the components of the gradient operator in an orthonomal basisei, or = eii. (2.23)The relativistic generalisation is a covariant 4-vector , which is dened as = x =

1ct, xi

. (2.24)The transformation law for may be found from the chain rule:x x

= xx

x (2.25)However, by dierentiatingx= (1)x

, (2.26)which follows from equation (2.6), we nd thatx

= (1)x. (2.27)Hence the 4-vector dierential really is a covariant 4-vector.2.1.7 4-vectors and the scalar productTo recap, a contravariant 4-vector can be dened as any collection of four quan-tities which transform the same way as dxunder a Lorentz transformation.Similarly, a covariant 4-vector can be dened as a collection of four quantitieswhich transform the same way as . One can map any contravariant vectorinto a covariant one with the metric , an operation called lowering the indices,or one can go in the opposite direction with the inverse metric (raising theindices).There are many important quantities than can be assembled into 4-vectors:for example, energy E and momentum p of a particle belong together in a singlemomentum 4-vector p, dened byp= (E/c, p). (2.28)We are used to the interval ds2= dxdxbeing a Lorentz invariant quantity, butit is straightforward to see that one can form a Lorentz invariant quantity fromany pair of 4-vectors a, b. This is the scalar product, writtena b p2= ab = ab= a0b0a1b1a2b2a3b3. (2.29)11We can take the scalar product of the momentum 4-vector with itself and recoverthe relativistic relation between energy, momentum and mass:p2= E2/c2p2c2= m2c2, (2.30)Another invariant can be constructed by contracting the momentum 4-vectorwith the position 4-vectorp x = Et px. (2.31)Another important object is the velocity 4-vector. In Newtonian mechanics, thevelocity is dxi/dt, so it might be thought that dx/dt is the analogous quantity.However, dt is not Lorentz invariant, so the object as a whole does not transformlike dxas a 4-vector should. Instead, we should dierentiate with respect to aLorentz invariant quantity. There is one ready-made for us in the form of thespace-time interval ds, from which we can dene the proper time d = ds/c. The4-velocity of a particle can then be written asv= dxd . (2.32)2.1.8 The dAlembertianThere is a second order dierential operator which students of electromagnetismwill already have come across, called the wave operator or the dAlembertian. Itis sometimes given its own symbol 2:2 = 1c22t2 2. (2.33)This is already Lorentz invariant, for it may also be written as . You mayrecall that it was partly consideration of the theory of electromagnetic waves thatled Einstein to formulate the special theory of relativity: electromagnetism (infree space) is automatically a theory which is consistent with the special theoryof relativity, a fact which becomes obvious when it is written down in terms of4-vectors and tensors. This is not at all obvious if it is written down as Maxwelloriginally did, component by component.2.1.9 Lorentz covarianceThe power of 4-vector formalism is that when one uses it consistently to writedown equations, the equation will keep the same form under a Lorentz trans-formation: that is, it will be Lorentz covariant. For example, we might want arelativistic generalisation of Newtons Second Law, F = dp/dt. We have alreadyseen that dierentiations with respect to time should be changed to dieren-tiations with respect to proper time, so the analogous equation in relativisticdynamics should beF= dpd . (2.34)12The zeroth component of the force 4-vector, F0, is the (proper) rate of change ofthe energy of the particle.A potential source of confusion is whether physical quantities are representedby covariant or contravariant vectors. As a rule, coordinates and momenta nat-urally have their indices up and derivatives naturally have their indices down.2.2 Klein-Gordon equationThe Klein-Gordon equation was originally thought up by Schrodinger, who wanteda relativistic wave equation describing the electron, even before he settled on whatwe now call the Schrodinger equation,i h t = h22m2. (2.35)Klein and Gordon published rst, so they got the credit. We will see laterwhy Schrodinger dropped this equation. One of the ways of understanding theSchrodinger equation is to recall that in quantum mechanics, physical quantitiesare represented by operators:E i h t, p i h, (2.36)(where the double-headed arrow symbol means is represented by). Hence theSchrodinger equation (2.35) represents the non-relativistic equation for the kineticenergy E = p2/2m.As we know, the relativistic relation between energy and momentum is E2=p2+ m2, which seems to suggest that a relativistic version of the Schrodingerequation ought to be2t2 = 2 + m2, (2.37)or, in a manifestly Lorentz covariant form,(2+ m2) = 0. (2.38)(It is traditional to use rather than in this context).The solutions to this equation are(x) = AeiEt+px(2.39)with E = =(p2+ m2), and A an arbitrary complex constant.In fact if is a solution to the Klein-Gordon equation it cannot be interpretedas a wave function as Schrodinger discovered. A wave function is a probabilityamplitude, whose modulus squared is the probability of nding the particle at13a particular position (or with a particular momentum). For example, in non-relativistic quantum mechanics, the probability density = [[2, which is asso-ciated with a probability current j = i( )/2m, in the sense thattogether they make up a probability conservation equation + j = 0, (2.40)an equation which one can check by dierentiating with respect to time andusing the Schrodinger equation.The KG equation also has a conserved density and a 3-vector current, whichare = i2(t t), j = i2( ). (2.41)However, this density cannot be interpreted as a probability density, as it is notpositive denite. This was the reason that Schr odinger chose the non-relativisticform for his equation. This was reasonable for his purposes, but the Klein-Gordonequation makes a comeback later, when we shall see that can be interpreted asa charge density, which is allowed to take both positive and negative values.2.3 Maxwells equations in covariant formThis section recaps some important results, and introduces the formulation ofthe theory in an explicitly Lorentz covariant manner.Firstly, we recall Maxwells equations in free space, writing them down innatural units, in which the permittivity and permeability of free space 0 and 0are both unity:Homogeneous InhomogeneousB = 0 E = tB+ E = 0 tE + B = j(2.42)The homogeneous and inhomogeneous (i.e. having a source term on the righthand side) equations have diering status. The homogeneous equations imply theexistence of potentials and A from which the physically measurable quantitiesE and B can be calculated. These potentials are specied only up to a gaugetransformation, and A A+ , where is an arbitrary functionof space and time. The inhomogeneous equations imply that the source termsmust obey a current conservation equation, + j = 0. To summarise:Homogeneous Inhomogeneouspotentials , A current conservationB = AE = A A A+ + j = 0(2.43)14All these quantities can be assembled into explicitly Lorentz covariant objects.The gauge potentials belong together in a 4-vector potential A= (, A), whilethe charge density and the current density j can be put together into a 4-vectorcurrent density j= (, j). Recall that putting quantities together into 4-vectorsis not just a matter of notation: it means that the quantities transform just likethe space-time coordinates xunder a Lorentz transformation.The electric and magnetic elds E and B also belong together in a Lorentzcovariant object, as they are mixed up by Lorentz transformations (an observermoving through a magnetic eld also sees an electric eld). However, this objectcannot be a 4-vector as there are a total of 6 components of the electric andmagnetic elds, when they are taken together. In fact, the object is an anti-symmetric tensor, the eld strength tensor F, which is dened in terms of thegauge potential 4-vector A:F= AA. (2.44)The Lorentz transformation law for Ffollows from those of and A:FF

= F, (2.45)from which we explicitly see that Fis a covariant tensor of rank 2. Let usexamine the components where = 0 and ranges over spatial indices i:Fi0= iA00Ai= iA0 Ai= i Ai= Ei, (2.46)where we have used the expression for the electric eld in terms of the gaugepotentials in equation (2.43). The magnetic eld is contained in the entrieswhere both and take spatial values:Fij= iAjjAi= iAj+ jAi= ijkBk, (2.47)where we have introduced the Levi-Civita symbol ijk. The Levi-Civita symbolis dened by

ijk =

+1 if i = j = k cyclic,1 if i = j = k anticyclic,0 otherwise.(2.48)Written in matrix form, the eld strength tensor is:F=

0 E1E2E3E10 B3B2E2B30 B1E3B2B10

. (2.49)The reverse relations may be writtenEi= Fi0, Bi= 12

ijkFjk, (2.50)15which can be veried with the help of the identity

ijk

klm = iljmimjl. (2.51)When expressed in terms of 4-vectors and tensors, electromagnetism looks verysimple and beautiful. For example, the covariant expression of the current con-servation equation is simplyj= 0. (2.52)Maxwells equations become:Homogeneous InhomogeneousF+ F+ F= 0, F= j. (2.53)The rst of these expressions looks as if it contains many more equations than the4 of the original homogeneous Maxwell equations. However, the antisymmetry ofthe eld strength tensor (F= F) means that the expression is trivial if anyof the two indices are equal. Thus all the indices , and must take dierentvalues, and the number of ways of choosing three dierent numbers from a setof four is 4C3, which is equal to 4, precisely the number of equations we startedwith.We can use the four-dimensional Levi-Civita tensor to re-express the homo-geneous equations more compactly. This tensor has four indices, and is denedby

=

+1 if = = = , symmetric,1 if = = = , antisymmetric,0 otherwise.(2.54)A symmetric permutation of the indices in one in which an even number ofindices are exchanged, while an antisymmetric permutation is one for which anodd number of indices are exchanged. Thus, for example, 0123 = 1032 = +1,

1023 = 0132 = 1, but 0012 = 0.There is also a version with the indices raised:

=

, (2.55)whose symmetric permutations take the value 1 and antisymmetric ones +1.Using this totally antisymmetric (i.e. antisymmetric on the exchange of anytwo indices) tensor, the homogeneous Maxwell equations become

F= 0. (2.56)One of the problems uses this expression to give you practice with the summationconvention as applied to these tensors.162.3.1 GaugesRecall that the gauge potential Ais specied only up to an arbitrary func-tion of space and time (see equation 2.43). The 4-vector version of a gaugetransformation is writtenAA, (2.57)under which one can check that Fis unchanged. Indeed, using (2.44) one ndsF(A) (A) = F() = F. (2.58)The last step follows because we can take the partial derivatives in any order.In order to use the gauge potential to solve the eld equations, one shouldeliminate this freedom to make gauge transformations, otherwise one can getmisled into thinking that the solutions one obtains are all physically distinct.There are many ways to do this, but some of the most common are as follows.Coulomb gauge This is also called radiation gauge, and is dened byA = 0, (2.59)which implies that the rst of the inhomogeneous equations in (2.42) becomes2 = . (2.60)This gauge is therefore very useful in solving electrostatics problems, and weshall use it later on when studying the Casimir eect. However, it is not so oftenused in relativistic applications, as the gauge condition does not respect Lorentzinvariance.The second of the inhomogeneous Maxwell equations becomes, on using thegauge condition,t(A) 2A = j, (2.61)which we may write as( 2t2 2)A = jT, (2.62)where jT = j + . One can show, using the equation of current conservationand Eq. (2.60), thatjT = 0. (2.63)This of course is necessary for the consistency of Eq. (2.62), because if one takesthe divergence of the left hand side one gets zero as well.In free space, that is, in the absence of charges and currents, we may take = 0, and the equation for the vector potential becomes( 2t2 2)A = 0. (2.64)17One of the important discoveries of the last century was that this equation hasplane wave solutions, which carry energy and momentum: electromagnetic waves.Such a solution may be writtenA(t, x) = aeit+ikx(2.65)where a is a constant 3-vector, and = [k[.The gauge condition (2.59) imposes a condition on the vector a, forA = ika eit+ikx= 0, (2.66)which implieska = 0, (2.67)Hence the gauge potential A is orthogonal to the wave vector k, which removesone of the three apparent degrees of freedom of the eld. Thus the eld oscillatesin directions orthogonal to the direction of propagation: we say that the wavesare transverse.Lorentz gauge This is very often the gauge to choose if one is interested inwave propagation in electromagnetism. It is dened byA= 0, (2.68)and is manifestly Lorentz invariant. In this gauge the equation of motion for thegauge eld is simplied, forF= AA= 2A= j. (2.69)However, an irritating feature of this gauge is that it does not quite specify Afully. One can still make a gauge transformation AA which satisesthe Lorentz gauge condition (2.68), as long as the function satises 2 = 0.Such functions are called harmonic. In classical eld theory this is not too muchof a problem, but in setting up the quantum theory of gauge elds care must betaken.In free space, the eld equation is (2.69)2A ( 2t2 2)A= 0, (2.70)which looks very much like four copies of the Klein-Gordon equation for eldswith zero mass. Plane wave solutions may be writtenA(t, x) = aeikx(2.71)where we recall that k x = k0t kx, and k0= [k[, Again, ais a constant4-vector.18The Lorentz gauge condition (2.68) areA= ikaeikx= 0, (2.72)which implies thatk a = 0. (2.73)Once again the eld is orthogonal to the wave vector k, but this time in the4-vector sense.Once the gauge condition is taken into account there are apparently threedegrees of freedom, one more than in the Coulomb gauge. However, one ofthose is unphysical as a result of the remaining freedom to make harmonic gaugetransformations.3 Lagrangian formulation of eld theory3.1 Lagrangian and Hamiltonian mechanicsThe starting point for classical mechanics is Newtons Second Law of Motion,which states that the rate of change of momentum of a particle is proportionalto the applied force. For a system of P particles, we may write pA = FA, (3.1)where the index A runs from 1 to P. The momentum pA = mAvA, wherevA = xA. If the mass is constant in time, we obtain the formulation mA xA = FA.If the force is conservative, which means that the work done by the forcearound a closed path is zero, then the force can be written as the gradient of apotential energy V :FA = AV (x1, . . . , xP). (3.2)There may also be constraints on the motion: for example, the particles maybe forced to move on a sphere. This reduces the number of degrees of freedom ofthe system. A very common form of constraint is one which relates some or all ofthe coordinates through a function (which may also depend explicitly on time):f(x1, . . . , xP, t) = 0. (3.3)These are known as holonomic constraints. For example, the rst particle may beconstrained to move on a sphere of radius a, or x21 = a2. Thus we lose one degreeof freedon, the radial coordinate of the particle. If there are k such constraints,then the eective number of degrees of freedom is reduced by k: P particlesmoving in 3 dimensions under k constraints have 3P k degrees of freedom.Thus not all of the coordinates are needed to describe the motion of the system,and it is convenient to introduce the notion of generalised coordinates qa which19result from the solution of the constraint equations. In the system of particlesthere will be N = 3P k of them, and one can express the original coordinatesas functionsx1 = x1(q1, . . . , qN, t)... ... (3.4)xP = xP(q1, . . . , qN, t)Generalised coordinates are also useful in systems without constraints: for exam-ple, orbits around central potentials, where it is convenient to use spherical polarcoordinates.3.1.1 Lagrangian MechanicsThere is a later and more sophisticated formulation of classical mechanics due toLagrange. We dene the kinetic and potential energies T and U in terms of thegeneralised coordinates of a system asT = 12ama q2a, U = U(qa, qa) (3.5)(we have allowed for the possibility of a velocity-dependent potential). We thenintroduce a new function called the Lagrangian, dened asL = T U. (3.6)The equations of motion of the system are thenLqa t

L qa

. (3.7)These equations are known as the Euler-Lagrange equations.For example, suppose there is no velocity dependence in the potential. ThenLqa= Uqa, L qa= ma qa, (3.8)so thatt(ma qa) = Uqa. (3.9)Thus we return to Newtons Second Law. The quantities L/ qa can be identiedas a momentum associated with the coordinate qa if the qa are simply Cartesiancoordinates, then they are the ordinary mechanical momenta, but in generalwe refer to them as conjugate or canonical momenta, which are dened as the20derivative of the Lagrangian with respect to the rate of change of the generalisedcoordinates:pa = L qa. (3.10)The Euler-Lagrange equations can be derived from a variational principle,known as Hamiltons Principle or the Principle of Least Action (which is a mis-nomer). The action for a system between times t1 and t2 is dened asI =

t2t1dt L(qa, qa, t). (3.11)The Principle of Least Action states that the action is an extremum for the pathof the motion.To see this, let us consider a small variation in the path, vanishing at t1 andt2, sends qa(t) q

a(t) = qa(t) +qa(t) (see Figure 3.1). The action also changes,I I

= I + I, where to rst order in qa(t),I =

t2t1dt L(qa, qa, t) =

t2t1dt

Lqaqa + L qa qa

. (3.12)Integrating by parts we ndI =

t2t1dt

Lqa t

L qa

qa +L qaqat2t1. (3.13)The second term vanishes, as we are keeping the end points of the path xed,and so if the variation of the action I vanishes we nd thatLqa t

L qa

= 0, (3.14)which are precisely the Euler-Lagrange equations (3.7).Figure 3.1: A path q(t) for a system with one coordinate, and a smallvariation q(t) to that path. Classical paths are ones for which the action(see Equation 3.11) is stationary under such small variations.3.1.2 Hamiltonian MechanicsThere is yet another formulation of classical mechanics due to Hamilton. It mayseem excessive to have so many ways of doing the same thing, but each formu-lation has its own advantages: for example, the Lagrangian and Hamiltonian21formulations are very powerful in bringing out conservation laws in dynamicalsystems, and relating them to symmetries.Recall that associated with each coordinate qa there is also a conjugate mo-mentum pa, dened by Eq. (3.10). We can therefore express the Euler-Lagrangeequations as pa = L/qa. We now dene an important function called theHamiltonian from the Lagrangian:H(pa, qa, t) =apa qaL(qa, qa, t). (3.15)The idea is to replace qa by pa, having solved equation (3.10). The rst of Hamil-tons equations follows from the Euler-Lagrange equations (3.7), after noting thatL/qa = H/qa. It is pa = Hqa. (3.16)The second follows straightforwardly from the denition of the Hamiltonian (3.15)upon partial dierentiation with respect to pa: qa = Hpa. (3.17)If the Lagrangian does not depend explicitly on time, the Hamiltonian is a con-stant of the motion, as the following calculation shows. Recall that H is a functionof qa, pa, and possibly t only. HencedHdt = Ht +a

Hpa pa + Hqa qa

. (3.18)Using Hamiltons equations (3.16,3.17) we see thatdHdt = Ht = Lt . (3.19)Thus if L/t = 0, the Hamiltonian is constant. One can also show thatH = T + U, (3.20)where T and U are the kinetic and potential energies introduced in (3.5). Thusthe Hamiltonian can be identied with the total energy of the system.As a trivial example, consider a set of N free particles with mass m. TheLagrangian isL = 12am q2a, (3.21)from which we can easily derive the Euler-Lagrange equations0 t(m qa) = 0, (3.22)22or that the acceleration of each particle is zero. The conjugate momentum ispa = m qq, and hence the Hamiltonian becomesH = 12ap2am. (3.23)This is of course the familiar non-relativistic expression for the kinetic energy.3.2 Lagrangian mechanics for the real scalar eldA eld is an object which takes a value (e.g. a real number) at every point in space-time. We are probably most familiar with the electric and magnetic vector eldsE(t, x) and B(t, x), which we have seen can be unied into the electromagneticeld strength tensor F(x). The Klein-Gordon eld (t, x), which takes complexvalues, is another example.A very powerful way of classifying elds is to organise them according to theirproperties under Lorentz transformations. The Klein-Gordon eld is an exampleof a scalar eld, which transforms as a scalar (a pure number):(x)

(x

) = (x), (3.24)where x

= x.In eld theory, the dynamical coordinates, the analogues of the qa of classicalLagrangian mechanics, are the values of the elds at every point. A eld thereforehas an uncountably innite number of degrees of freedom, which is the sourceof many of the diculties in eld theory. Sums over the label a are replaced byintegrals over space: for example, the Lagrangian can be expressed as the integralover space of an object called the Lagrangian density L. For a scalar eld, theLagrangian density can be expressed as a function of the eld, its time derivative,and also its spatial derivatives:L =

d3x L(, , , t). (3.25)The appearance of spatial derivatives may at rst sight be puzzling, but arisesnaturally from terms coupling neighbouring space points, in the limit that theseparation goes to zero. The puzzle is perhaps more that higher derivatives dontappear: in practice, they seem not be relevant for most applications in particlephysics.From a Lagrangian we can derive Euler-Lagrange equations, by the applica-tion of Hamiltons Principle to the action for a scalar eld,I =

t2t1Ldt =

t2t1dtd3x L(, , , t). (3.26)23Figure 3.2: A space-time diagram (with the z coordinate suppressed)of the four-dimensional region in which the variation of the eld (x)is non-zero.Consider an arbitrary variation in the eld, (x) (x) + (x), where (x)vanishes at t1 and t2, and outside an arbitrary spatial region R with boundary B(see Figure 3.2). This variation causes a small change in the action:I I

= I + I. (3.27)Taylor expanding the Lagrangian density, we ndI

t2t1dtd3xL + L + L + L , (3.28)whereL =

L(x), L(y), L(z)

. (3.29)We can separate out the variation in the action I, and integrate the last termby parts, to obtainI =

t2t1dtd3xL + L t

L

+L t2t1. (3.30)In this last expression the last term vanishes because of the conditions placed on at t1 and t2. Hence, upon integrating by parts again, we getI =

t2t1dtd3xL L t

L

+

t2t1dt

BdS L.(3.31)Again, the last term vanishes, as we supposed that the variation vanished onthe boundary B. Hence we are left withI =

t2t1dtd3xL L t

L . (3.32)The argument proceeds by noting that not only is arbitrary, but so is theregion R. Hence the condition that the action be stationary reduces toL L tL = 0. (3.33)24This is the Euler-Lagrange equation for a scalar eld.We can put it into a more obviously relativistic form by noting that =(t, i), so that the Euler-Lagrange equation becomesL iL(i) 0L(0) = 0, (3.34)orL L() = 0. (3.35)If there are many elds a we can derive an Euler-Lagrange equation for each ofthem,LaL(a) = 0. (3.36)We are now in a position to ask the question: from what Lagrangian densitydoes the Klein-Gordon equation follow? The Klein-Gordon eld may be split intoits real and imaginary parts, (x) = R(x) +iI(x), each of which obey the sameequation( + m2)R(x) = 0, ( + m2)I(x) = 0. (3.37)Consider rst the following Lagrangian density for the real part,LR = 12RR 12m22R. (3.38)We can immediately deriveLR= m2R, L(iR) = iR, L(0R) = 0. (3.39)Hence,L(R) = (0R, iR), (3.40)and soLRL(R) = m2RR = 0. (3.41)Thus the Lagrangian (3.38) implies the correct equation for the real part of theKlein-Gordon eld: exactly the same argument may be made for the imaginarypart of the eld, so we can write for the total lagrangianL = LR +LL = 12RR 12m22R + 12II 12m22I= 12 12m2. (3.42)25Lastly in this section, we will write down the Hamiltonian for a single realscalar eld. Drawing on our experience with nite numbers of degrees of freedom,it seems natural to dene a conjugate momentum(x) = L (x). (3.43)Thus for our example (3.38), the conjugate momnetum (x) = (x). The Hamil-tonian for the real scalar eld is dened byH =

d3x (x) (x)

d3xL(, i, ). (3.44)If the eld obeys the Klein-Gordon equation, its Hamiltonian isH =

d3x122+ 12()2+ 12m22

. (3.45)We talk of the three terms in the Hamiltonian as kinetic, gradient and potentialterms respectively. One can allow more general functions of in the potentialterm than just a quadratic:H =

d3x122+ 12()2+ V ()

. (3.46)In classical eld theory, V can be any function of , as long as it is bounded below,otherwise the energy of the eld can become arbitrarily negative. Quantum eldtheory is much more restrictive about the terms it allows: it turns out that thetheory does not make sense in four space-time dimensions unless V () is restrictedto be a polynomial of degree no more than four.3.3 Lagrangian mechanics for the electromagnetic eldThe electromagnetic gauge potential A(x) is an example of a 4-vector eld, whichis a eld which transforms like a covariant 4-vector. Its behaviour under Lorentztransformations isA(x) A

(x

) = A(x), (3.47)where x

= x. In this section we will write down a Lagrangian density, andshow that the resulting Euler-Lagrange equations are just the inhomogeneousMaxwell equations. The homogeneous ones will be shown to be an identity,rather than arising from any variational principle. The Lagrangian isL = 14FFjA, (3.48)26where F = A A. The generalised coordinates of this system are thefour components the gauge eld A. We can therefore derive Maxwells equationsby looking at the Euler-Lagrange equations:L(A) LA= 0. (3.49)Let us rst consider the variation with respect to A, which givesLA= jAA= j = j. (3.50)The variation with respect to A is a little more complex, and it will help todene X = A, so that F = XX. HenceL(A) = 14FXF 14FFX,= 14

F 14F () ,= 14(FF) 14(FF)= F. (3.51)Thus we can substitute (3.50) and (3.51) back into the Euler-Lagrange equation(3.49) to obtainF+ j= 0, (3.52)which we recognise as the covariant form of the inhomogeneous Maxwell equa-tions.We now turn to the homogeneous Maxwell equations, and demonstrate thatthey are an identity, which is sometimes known as the Bianchi identity. Thiscan be seen by examining the 4-vector k= 12

F= 0. Substituting theexpression for the eld strength tensor in terms of the gauge potential, we obtaink= 12

(AA). (3.53)By exchanging the labels and in the second term we ndk= 12

A 12

A, (3.54)= A, (3.55)with the last step following by the antisymmetry of the Levi-Civita tensor. Nowwe exchange the labels and , from which we can deduce

A= A. (3.56)27The antisymmetry of the Levi-Civita tensor then shows that

A= A. (3.57)The only number that is equal to its negative is zero, hence12

F= 0. (3.58)Thus we have demonstrated that the Principle of Least Action applied to theLagrangian (3.48), combined with the Bianchi identity (3.58), results in Maxwellsequations.3.4 Noethers theorem and conservation lawsIt is a profound feature of both classical and quantum dynamics that conservationlaws, such as conservation of energy or momentum, follow from symmetries of thedynamical equations. In this section we shall show how this feature is realised inclassical eld theory.A conservation law can be expressed in two equivalent ways. Firstly, and mostobviously, one can say that there is a quantity Q which is constant in time, i.e.dQdt = 0. (3.59)Entirely equivalently, one can also say that there is a 4-vector current j(x) whichsatises a continuity equation,j(x) = 0. (3.60)If we integrate (3.60) over all space, we nd

d3x tj0(x) +

d3x j = 0. (3.61)Dening Q = d3x j0(x), we nd using Gausss Law thatdQdt =

dS j, (3.62)where the surface integral is taken over a sphere at spatial innity. With theassumption that all currents die o at innity, we havedQdt = 0. (3.63)Hence the existence of a 4-vector current satisfying a continuity equation j(x) =0 automatically implies that there is a conserved quantity, which is the integralover all space of the time component of the current.28Now consider a theory with N real elds a(x), with a = 1, . . . , N, and anactionS =

d4x L(a, a). (3.64)Noethers theorem for this theory can be stated as follows.For every transformation a(x) =

a(x

) which leaves the action S invariant,there is a conserved current.The proof of this theorem is quite involved, and we rst need some denitions.Let us rst dene the total variation in the eld,a(x) =

a(x

) a(x). (3.65)Note that this variation can be split into two pieces, the rst due to the fact thatthe functions a are changing:a(x) =

a(x) a(x), (3.66)which denes what we mean by the operator . The remaining piece can beinterpreted as the change in the eld functions due to the change in coordinates,a(x) = a(x

) a(x) a(x)x, (3.67)where x= x

x. Hencea(x) a(x) +

a(x) a(x) + a(x), (3.68)where we have dropped at term of second order in the variation, as

a(x) =a(x) + O(x).One can show straightforwardly from the denition of that(a(x)) = (a(x)), (3.69)but that(a(x)) = (a(x)) a(x)x. (3.70)Hence the partial derivative operator commutes with , but not with .Now consider the eect of the variation in Eq. 3.65 on the action:S =

d4x

L

(x

)

d4x L(x), (3.71)where L

(x

) = L(

a(x

),

a(x

)), and is an arbitrary region in spacetime. Byadding and subtracting L(x) to the rst integrand, we seeS =

d4x

L

(x) +

d4x

L(x)

d4x L(x). (3.72)29We now need to express the integration measure d4x

in terms of d4x, usingthe Jacobian of the coordinate transformation:d4x

= d4x

(x

0, . . . , x

3)(x0, . . . , x3)

= d4x[ det M[, (3.73)where the components of the matrix M are given byM = x

x + x. (3.74)In matrix notation, we may write M = 1 + M, and using the standard matrixidentityln det M = tr ln M, (3.75)we see thatdet M 1 + tr M = 1 + x. (3.76)Substituting into the varation of the action Eq. (3.71), we ndS =

d4x (1 + x)L

(x) +

d4x (1 + x)L(x)

d4x L(x). (3.77)Dropping terms which are second order in the variation we arrive atS =

d4x L(x) +

d4x xL(x) (3.78)Now we can again split the variation in the Lagrangian function into a piecearising from the change in the function itself, and that due to the change ofcoordinates:L(x) = L(x) + L(x)x. (3.79)The change in the Lagrangian function arises because of the change in the func-tional form of the elds, soL(x) = La(x)a(x) + L(a)(x)(a(x)). (3.80)At this point, the commutativity of and stated earlier means thatS =

d4x

La(x)a(x) + L(a)(x)(a(x))+

d4x (L(x)x+L(x)x)

. (3.81)Some simple algebra shows thatS =

d4x

LaL(a)

a(x) +

d4x

L(a)a +L(x)x

.(3.82)30Note that the rst term is precisely the variation in the action we found whenderiving the Euler-Lagrange equation, where we considered variations in the eldsalone, vanishing on the boundary of . Suppose we consider a particular eldconguration (x) satisfying the Euler-Lagrange equation. If we then make aninnitesimal transformation on it which leaves the action invariant, i.e. whichresults in no rst-order change in the action, we must have thatf= 0, (3.83)wheref= L(a)a(x) +Lx, (3.84)or substituting back for (x),f= L(a)a(x)

L(a)a(x) L

x. (3.85)Thus we see the appearance of an innitesimal 4-vector, f, which satises acontinuity equation.3.4.1 Conservation of energy-momentumConsider a translation on the coordinates, with the elds unchanged:x= , a = 0, (3.86)where is a constant 4-vector. This clearly includes both space and time trans-lations. Hence we can writef=

, (3.87)where we have introduced the canonical energy-momentum tensor = L(a)a(x) L. (3.88)Because the translations are arbitrary, translation invariance of the Lagrangianimplies that there must be four continuity relations, = 0. (3.89)The conserved charge associated with these four continuity relations is the total4-momentum,P =

d3x0(x). (3.90)314 Canonical quantisationWe have seen that the Klein-Gordon eld (x) does not have the interpretationas a single-particle wave-function, as it does not have a conserved probabilitycurrent, and it possesses negative energy states which cannot form a Dirac sea.However, as we shall see in this section, the theory of the quantum eld operator(x) is in fact a relativisitic many-particle quantum theory. The process of de-veloping the theory of a quantum eld is sometimes misleadingly called secondquantisation, perhaps because of some idea that one is quantising a wavefunc-tion, itself a quantum object. It should be emphasised at the outset that thescalar eld, real or complex, is not a wavefunction.In this section we go through the quantisation procedure for free elds. Byfree we mean non-interacting: once a state has been set up there are no tran-sitions to any other states. In practice this means that the Lagrangian can bewritten as a quadratic function of the elds. Free elds may sound somewhatirrelevant, but nearly all quantum eld theories have to be treated as small pertur-bations away from free eld theories, so it is an appropriate place to start. Thereare non-trivial eects in free eld theory which come from boundary conditions:one of these eects is the Casimir eect in the free electromagnetic eld.4.1 Quantisation of nite systemsIn this section we shall proceed with quantising the real scalar eld. Firstly, wewill go through the quantisation procedure for a system with a nite numberof degrees of freedom, to show that quantising a eld is not unlike quantising amore familiar system. Suppose the system has N coordinates qa with associatedvelocities qa, and LagrangianL = 12Na=1m q2aU(qa) (4.1)(we are supposing for simplicitys sake that all the masses are the same). Eachcoordinate is also associated with a conjugate momentum pa, dened as thedierential of the Lagrangian with respect to the velocity qa, which in this casehas the value m qa. The Hamiltonian, dened as 12a qapaL, is thenH =Na=1p2a2m + U(qa). (4.2)The procedure of canonical quantisation replaces the classical conjugate variablesqa, pa by operators obeying the following commutation relations[ qa, pb] = iab, [ qa, qb] = 0 = [ pa, pb]. (4.3)32The Hamiltonian is also replaced by an operator H, which generates the timeevolution of the system. We are probably used to thinking about time evolutionas the change in time of the state of the system, governed by the Schrodingerequationi t[(t)` = H[(t)`. (4.4)However, there is another way of thinking about the time evolution of a quantummechanical system known as the Heisenberg picture (to distinguish it from theSchrodinger picture above).In the Heisenberg picture, the states [`H that are time-independent, and thetime evolution of the system is transferred to the operators OH(t), which satisfythe Heisenberg equationsi ddtOH(t) = [OH(t), H]. (4.5)The states and operators in these two pictures can be related by a unitary trans-formation eiHt(a unitary operator is one whose inverse is its Hermitian conju-gate). Thus, if we suppose that the two sets of operators and states coincide att = 0, they are related at all times byOH(t) = eiHtOSeiHt[`H = eiHt[(t)`S. (4.6)This ensures that the matrix elements are the same. To see this, let us considerthe matrix elements between any two states [`H and [

`H in the Heisenbergpicture, and apply these transformations:H'[OH(t)[

`H = S'(t)[eiHteiHtOSeiHteiHt[

(t)`S= S'(t)[OS[

(t)`S. (4.7)This demonstrates that the Heisenberg picture matrix elements are the same asthose in the Schrodinger one, which means there is no dierence in the physicalpredictions one makes.4.2 The real scalar eldWe begin with the Lagrangian densityL = 12 12m22, (4.8)from which we can dene the conjugate momentum(x) = L (x). (4.9)33The Hamiltonian density is then found fromH = (x) (x) L = 122+ 12()2+ 12m22. (4.10)In free quantum eld theory we usually work in the Heisenberg picture, where theelds carry the time dependence, although it is possible to use the Schrodingerpicture. The eld theory is quantised by constructing the equal time canonicalcommutation relations[(t, x), (t, x

)] = i h3(x x

), (4.11)[(t, x), (t, x

)] = 0, (4.12)[(t, x), (t, x

)] = 0, (4.13)where 3(x x

) = (x1 x1)(x2 x2)(x3 x3). One can see the analogybetween the eld commutation relations and (4.3) if one recalls that the spacecoordinate x is like the label a in the system with a nite number of degrees offreedom. The -function acts in the same way for a continuous label x as theKronecker does for a discrete label.So in quantum eld theory, the elds become operators, obeying the Heisen-berg equations(x) = 1i

(x), H

, (x) = 1i

(x), H

, (4.14)where H is the quantum Hamiltonian, constructed by replacing the classical eldsby operators in the original:H =

d3x

12 2+ 12()2+ 1222

. (4.15)It is straightforward to show for a free eld that(x) = (x), (x) = (2m2)(x). (4.16)Hence the eld operator obeys the same equation of motion as the classical eld.It is convenient to expand the eld operator in terms of eigenfunctions of theoperator (2+ m2),(2+ m2)eikx= 2keikx. (4.17)Writing(x) =

d3k(2)3 (fk(x) a(k) + fk(x) a(k)) . (4.18)with fk(x) = Ak(t)eikx, we see that fk(x) obeys the eld equation (2+m2)fk(x) =0 ifAk(t) = ^keikt(4.19)34The normalisation factor ^k is essentially arbitrary, although almost everyoneuses one of two conventions outlined below. The properties of the operators a(k) and a(k) can be deduced from the commutation relations of and . Wetherefore need expressions for a(k) and a(k) in terms of and , where (x) = (x) =

d3k(2)3

ik^k a(k)eikx+ ik^k a(k)eikx

. (4.20)Firstly, we take the Fourier transform of the eld operator and its conjugatemomentum:(k

, t) =

d3x(x)eik

x, (k

, t) =

d3x (x)eik

x, (4.21)and use the relation

d3xei(kk

)x= (2)33(k k

) (4.22)to obtain(k

, t) = ^k a(k

)eik t+^k a(k

)eik t(4.23) (k

, t) = ik ^k a(k

)eik t+ ik ^k a(k

) a(k

)eik t. (4.24)Hence2k^k a(k

) =

d3x(i (x) + k

(x))eik

x. (4.25)By complex conjugation one obtains2k^k

a(k

) =

d3x(i (x) + k

(x))eik

x. (4.26)One can then derive the commutation relation[ a(k), a(k

)] = (2)33(k k

)/2k[^k[2, (4.27)This looks very similar to the commutation relations for simple harmonic oscil-lators. We recall that the harmonic oscillator of unit mass is described by theHamiltonianH = 12p2+ 122x2, (4.28)where is the angular frequency of oscillation. Upon quantisation, the operators a = (i p + x)/2 and a = (i p + x)/2 satisfy the commutation relations[ a, a] = 1. (4.29)This correspondence suggests that one choice for the normalisation factor:^k = 12k, (4.30)35in which case we nd for the complete set of commutation relations[a(k), a(k

)] = (2)33(k k

), (4.31)[a(k), a(k

)] = 0, (4.32)[a(k), a(k

)] = 0. (4.33)The right hand side of the rst equation, with its -fucntion, is the closest wecan get to unity when we are dealing with functions of a continuous variable likea(k). However, we shall adopt a convention which keeps the Fourier expansionsof the eld operators Lorentz invariant,^k = 12k, (4.34)which means that the rst commutation relation becomes[ a(k), a(k

)] = 2k(2)33(k k

). (4.35)It is also quite common to work in a nite volume V , which means that insteadof integrals over k we have sums over allowed wavevectors k = 2(n1, n2, n3)/L,which go over to integrals in the innite volume limit:1Vk

d3k(2)3. (4.36)All wavefunctions are multiplied by a factor 1/V : hence the plane wave expan-sion of the eld operator is written(x) =k12kV

a(k)eikx+ a(k)eikx

. (4.37)The non-trivial equal time canonical commutation relation then becomes[ a(k), a(k

)] = k,k , (4.38)where k,k = n1n

1n2n

2n3n

3.To summarise, we have discovered that the eld theory behaves very muchlike innitely many harmonic oscillators, labelled by their wavenumber k, andwith a dierent frequency k = (k2+ m2). We will see that the occupationnumber n of each oscillator actually corresponds to the number of particles withthat particular momentum in the state considered. The raising and loweringoperators a(k) and a(k) are also called creation and annihilation operators,because they change the number of particles with 3-momentum k in a state.364.3 States of the scalar eld; zero point energyIn this section we will drop the hat notation for operators, unless there could beambiguity, as all elds will be assumed to be operators unless otherwise specied.We saw in the previous section that the scalar eld could be decomposed intoa sum over operators obeying oscillator-like commutation relations, one pair foreach wavevector k. It will therefore be no surprise to see that the states of thequantum eld theory are just towers of oscillator states, one for each k. Theground state is dened as the state annihilated by all the annihilation operators:a(k)[0` = 0, k. (4.39)We can use the creation operators a(k) to construct excited states: for example,the rst excited state in mode k may be written[k` = a(k)[0`. (4.40)The normalisation of these states follows from deciding that the vacuum statehas unit norm, that is '0[0` = 1. Thus'k

[k` = '0[a(k

)a(k)[0` (4.41)= '0[a(k)a(k

)[0` +'0[0`(2)3(k k

). (4.42)The rst term on the right hand side vanishes, and so we have that the rstexcited states satisfy the normalisation condition'k

[k` = (2)3(k k

). (4.43)The corresponding completeness relation is

d3k(2)312k[k`'k[ = 1. (4.44)The states [k` are interpreted in quantum eld theory as single-particle states,an interpretation discussed further below. We could continue applying creationoperators a(k1), a(k2), . . . to make states[k, k1, k2, . . .` = . . . a(k2)a(k1)a(k)[0`. (4.45)The general state could have multiple applications of creation operators with thesame momenta, in which case it is conventional to divide by a symmetry factor,to obtain[nk` =k(ak)nknk! [0`, (4.46)where nk is the number of times the creation operator with label k is applied.The Hilbert space spanned by the set of all states [nk` is called Fock space.37The basis states are interpreted as multiparticle states, which demonstrates oneof the special features of eld theory: it is a quantum mechanical theory ofmany particles. Ordinary non-relativistic quantum mechanics describes only oneparticle at a time.There is an operator which allows us to count the number of particles in agiven state. Let us rst introduce the set of operators n(k) by n(k) = a(k)a(k). (4.47)It is straightforward to see that they return zero when acting on the ground state, n(k)[0` = 0, (4.48)from the denition of the vacuum state (4.39). When acting on 1-particle states,we nd n(k)[k

` = a(k)a(k)a(k

)[0` (4.49)= a(k)

a(k

)a(k) + 2k(2)3(k k

)

[0`, (4.50)where we have used the commutation relations (4.33) to change the order of a(k)and a(k

). Thus, n(k)[k

` = 2k(2)3(k k

)[k`. (4.51)The operators n(k) are used to construct the number operatorN =

d3k(2)312k n(k), (4.52)which tells us the total number of particles in a particular state2. Thus, asexpected, N[0` = 0, meaning that there are no particles in the ground state. Inthe 1-particle state [k

`, we ndNtot[k

` =

d3k(2)3N(k)[k

` =

d3k(2)3(2)3(k k

)[k` = [k

`. (4.53)The state [k

` is indeed an eigenstate of the number operator, with eigenvalue 1 there is 1 particle in state. When applied to the general state (4.46), it can beshown thatN[nk` =

d3k(2)3nk[nk`. (4.54)The energy of the states can be computed from the quantum energy operatorE =

d3x00(x) =

d3x

12 2+ 12()2+ 1222

, (4.55)2if that state has a denite particle number it may be that the state is not an eigenvalueof the number operator38which is immediately seen to be identical to the Hamiltonian. Since the simplestates we have seen so far have been constructed from creation operators a(k)acting on the vacuum, it is more convenient to re-express the Hamiltonian oper-ator H in terms of creation and annihilation operators. Recalling the expansionsof and in terms of a(k) and a(k) in equations (4.18) and (4.20), we ndH =

d3k(2)312kk2 (a(k)a(k) + a(k)a(k)) . (4.56)If we try to change the order of the operators, we run into a problem, forH =

d3k(2)312kk2

2a(k)a(k) + (2)3(0)k

. (4.57)The second term in the brackets has the formally innite term (2)3(0). Thesource of this innity is the fact that we are working in innite spatial volume.Let us revert to a nite cubic region with sides of length L, in which case we canwrite(2)3(0) = limLlimk

k

L/2L/2d3xeix(kk

)= limL

L/2L/2d3x = V, (4.58)where V is a formal volume factor, and the limit of innite volume should strictlybe taken at the end of the calculation of any physical quantity. Thus we caninterpret the quantity multiplied by V as an energy density. However, even thisenergy density is innite, in general. Suppose we try to nd the energy densityof the ground state 0, for whichH[0` =

d3x0[0` =

d3k(2)3ka(k)a(k)[0` + 12

d3k(2)312kk(2)3(0)[0`.(4.59)Hence,0 = 12

d3k(2)3k. (4.60)This integral is divergent. If we integrate the momenta over the range 0 < [k[ ##### Three Physics Paradigms Newtonian, Relativistic, Quantum · PDF fileThree Physics Paradigms – Newtonian, Relativistic, Quantum ... Three Physics Paradigms – Newtonian, Relativistic,
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