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    Foundations of Laser Spectroscopy
    Foundations of Laser Spectroscopy
    Foundations of Laser Spectroscopy
    Ebook547 pages3 hours

    Foundations of Laser Spectroscopy

    By Stig Stenholm

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    One of the first texts to offer a simple presentation of the theoretical foundations of steady-state laser spectroscopy, this volume is geared toward beginning theorists and experimentalists. It assists students in applying theoretical ideas to actual calculations in laser spectroscopy with a systematic series of examples and exercises. Starting at an elementary level, students gradually build up their practical skills with demonstrations of how simplified theoretical models relate to experimentally observable quantities. Detailed derivations offer students the opportunity to work out all results for themselves.
    The first chapter introduces background material on electrodynamics and quantum mechanics, with an emphasis on the density matrix, its equation of motion, and its interpretation. Chapter 2 derives the response of the medium to strong fields. After mastering these two parts, students can proceed to later chapters in any order they wish. Succeeding chapters cover the physical basis of laser operation, applications central to laser spectroscopy, the inclusion of laser fluctuations into the theory, and field quantization. Numerous references, which appear in separate sections, form a concise history of the field and its most noteworthy developments.
    LanguageEnglish
    PublisherDover Publications
    Release dateSep 20, 2012
    ISBN9780486150376
    Foundations of Laser Spectroscopy

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      Foundations of Laser Spectroscopy - Stig Stenholm

      Index

      Interconnections between the Parts of Foundations of Laser Spectroscopy

      CHAPTER 1

      The Components of Theoretical Spectroscopy

      1.1. INTRODUCTION

      The laser is an electromagnetic oscillator that derives its energy directly from the excitation of matter. Often this is found in the internal quantum states of atoms or molecules, and then a quantum description of matter is inevitable. Only recently has it become possible to utilize free electrons for laser operation. The laser provides an optical light source with well-defined phase relationships, and hence it constitutes a straightforward extension of other electromagnetic oscillators. This justifies the use of the term quantum electronics. The light itself can, however, be described in a classical way for most applications.

      Spectroscopy is one field where much progress is due to the introduction of lasers. Even in linear spectroscopy, the laser is the ideal light source, but its large intensity over a narrow frequency range has allowed new effects; nonlinear spectroscopy becomes important. In this book we mainly discuss those features of laser spectroscopy that derive from its nonlinear properties. Many such features are, however, left out. All propagation problems in a nonlinear medium are neglected. Nonlinear mixing of light and its use to create new signals are not included. Such phenomena are discussed in many books on nonlinear optics. Their omission does lead to the exclusion of many techniques of importance to spectroscopy. We do not discuss Raman spectroscopy and its nonlinear generalizations. Some of these are straightforward extensions of the work in Chapter 4.

      To keep our treatment simple, we usually have to restrict the number of levels we choose to introduce. This is most easily justified for atoms; in molecules the density of states is so high that even laser spectroscopy must consider a multitude of levels, The basic theory remains the same, but the treatment becomes so involved that often only numerical work is possible. In solids the situation is even more complex. Thus, in this book, we mostly keep applications to atoms in mind.

      In strong fields the bound states of the electrons break up, and atoms and molecules ionize. This is an interesting phenomenon, which has important applications in particle detection and isotope separation. Such transitions between discrete bound and continuum states are left out of this book. They can be included in low-order time-dependent perturbation theory, but to provide a more complete treatment would lead us too far away from the point of view expressed here.

      This first chapter of the book reviews the basic knowledge used to formulate problems in laser spectroscopy. This includes the classical description of radiation fields and the quantum theory of matter. We devote most of the space to the microscopic description of bound states, including the effects of various phenomenologically introduced processes. The area of physics needed is vast, and many topics are left incomplete. Many statements cannot easily be justified exactly, but their introduction rests on heuristic or pragmatic arguments. The reader should not be afraid if these arguments are hard to grasp at first. If their use does not later provide the understanding desired, the references in the final section of this chapter may provide some illumination.

      The basic applications of the features of this chapter are given in Chapter 2. There the foundations are laid for nonlinear laser physics. After working through that chapter the reader can, more or less at will, read the various application parts independently of each other; to simplify this, a scheme of logical interconnections is provided at the beginning of the book.

      1.2. CLASSICAL DESCRIPTION OF RADIATION FIELDS

      The identity between electromagnetic and optical phenomena was established when Maxwell derived the propagation properties for electromagnetic radiation. Well-defined harmonic vibrations soon became the everyday tool of communication engineering, whereas the source of optical radiation was incandescent bodies. Their light was incoherent and random, and hence optical coherence seemed an elusive and slightly mysterious property. The ordinary radio transmitter, however, emits coherent waves, which is a necessary condition for their reception.

      When the laser appeared, the situation was changed. It could be understood as a classical oscillator emitting coherent light with well-defined phase properties. With a laser one can easily carry out optical diffraction experiments that were very hard to perform with thermal light sources. The rise of holography provides the best-known example.

      Laser research made it manifestly obvious that physical optics could be based entirely on Maxwell’s electrodynamic equations. For large amplitudes of the laser light, a classical description can be used, and a wave equation with well-defined amplitude and phase variables can be applied. The starting point must be Maxwell’s equations

      (1.1)

      (1.2)

      (1.3)

      (1.4)

      where E, D and H, B are the electric and magnetic fields respectively. The density of charges is p, and their current density is j. If we include only free charges into these, the bound neutrals can be taken to have an electric polarization density P giving the relation

      (1.5)

      whereas the magnetic dipole density M can be neglected for nonmagnetic media and hence

      (1.6)

      In nonlinear spectroscopy we can usually take the field-induced polarization P to be the main source of the fields in Maxwell’s equations, and a major part of the physics is concerned with the calculation of P for various cases of interest.

      It is easily seen that Eqs. (1.1) and (1.4) can be satisfied if we introduce the potentials A and ϕ so that

      (1.7a)

      (1.7b)

      The same fields are obtained if the potentials are transformed by an arbitrary gauge function χ(r, t) into

      (1.8)

      (1.9)

      In radiation problems one condition on the gauge is set by choosing auxiliary conditions for the potentials. In nonrelativistic calculations it is often advantageous to require that

      (1.10)

      This is the Coulomb guage.¹ Its advantage is that the potential ϕ is found to , satisfy (1.3) in the form

      (1.11)

      where the charge qi is situated at the position ri. The solution is the well-known function

      (1.12)

      This potential is taken to be entirely static. The remaining part of the equation is then used to describe the radiation fields, which are found to satisfy

      (1.13)

      they are said to be transverse. These transverse fields contain the optical radiation. In the following discussion we use only the Coulomb gauge.

      For spectroscopic purposes it is expedient to separate the formation of neutral bound constituents from the radiation field that induces transitions. The disadvantage is that such a division lacks relativistic invariance and holds only in our chosen frame, the laboratory. Because the binding energies are small, no relativistic energies enter our considerations, and we need to include no relativistic effects anyhow.

      If we combine Eqs. (1.1) and (1.2) with (1.5) and (1.6) we obtain

      (1.14)

      where we have used (1.13) too. This gives the equation for an electromagnetic wave propagating with velocity c = (ε0μ0)-1/2 and driven by the oscillating dipole moment P. For an oscillating point dipole situated at the origin we have

      (1.15)

      we can see the radiation driven by the transverse component of P only and

      (1.16)

      At a large distance R we observe at the time t the radiation

      (1.17)

      must be evaluated at the retarded time t - R/c. Equation (1.17) is the solution of (1.14) with the source (1.15). The magnetic field is given by

      (1.18)

      The energy flux is given by the Poynting vector

      (1.19)

      along the polar axis z. The total radiated power becomes

      (1.20)

      which is the conventional expression for an oscillating dipole.²

      If we want to solve (1.14) in a specific geometry, we introduce the eigenfunctions given by

      (1.21)

      and the boundary conditions appropriate to the system we consider. The functions Un are called the cavity eigenfunctions and can be chosen transverse

      (1.22)

      They will form a convenient basis set for all fields in the space we want to include in our considerations. We expand the electric field as

      (1.23)

      where the amplitude En(t) is determined by the orthogonality of the eigenfunctions Un to be

      (1.24)

      If the eigenfunctions are normalized to one, the denominator can be omitted.

      For the empty cavity, P = 0, Eq. (1.14) gives for En(t) the relation

      (1.25)

      where the angular frequency variable is defined by setting

      (1.26)

      we usually refer to angular frequencies simply as frequencies. The field En(t) oscillates at the frequency Ωn and continues to do so eternally if no damping occurs. If the field has a decay time τ we assume on phenomenological grounds that

      (1.27)

      This leads to the equation

      (1.28)

      When the system is damped only slightly, the oscillator quality factor

      (1.29)

      is large and (1.28) is equivalent with

      (1.30)

      As the loss factor is a small modification only, we assume that its effect can be included into (1.14) phenomenologically by a term of the form (1/τ)(∂/∂t) in all cases. For a strongly damped medium this assumption must be reconsidered.

      In many cases it is immaterial what type of boundaries one assumes for the space to be considered. It is then convenient to choose plane waves normalized in a box of volume V as the solution to (1.21). They are written in the form

      (1.31)

      The vectors ε(λn) are chosen as two orthogonal polarization directions satisfying the transversality condition

      (1.32)

      If we require the box to have periodic boundary conditions, the vectors kn become quantized to the values

      (1.33)

      where the box has the dimensions LxLyLz, and nx, ny, nz are three integers.

      For microwave radiation the box is often a real metallic cavity. For an ideal metal, the field can have no component along the surface, as this would immediately set up a current. Thus the tangential component must disappear on each surface. If the box is rectangular and placed as in Fig.

      Fig. 1.1 This picture shows a cavity resonator with metallic walls. The edges are of lengths Lx, Ly, and Lz respectively, and the eigenmodes consist of vector functions U(r), which are determined by the boundary conditions that the tangential electric field vanishes at the surfaces.

      1.1, the solution is easily found to be

      (1.34)

      (1.35)

      (1.36)

      We find directly that the boundary conditions are satisfied;

      (1.37)

      (1.38)

      and similarly for all other surfaces. The normal components do not disappear because of the choice of solutions. The wave vectors are of the form

      (1.39)

      The transversality condition (1.22) imposes one condition between the components of the vector A = (Ax, Ay, Az) in the form

      (1.40)

      For cavities of more complicated geometries the solution becomes more complicated than (1.34)-(1.36) but the basic idea remains the same.

      If one dimension, Lz say, grows to infinity, no boundary condition is needed in this direction. The solution becomes an exponential exp(ikzz), and the cavity becomes a wave guide. Such structures are widely used in microwave systems.

      The microwave oscillator, the maser, uses metallic cavities as resonance circuits. For optical wavelengths the use of closed metallic cavities is impractical because of the short wavelength. The optical oscillator, the laser, became possible when it was realized that a Fabry-Perot resonator can be used as a cavity.

      To achieve resonance in the plane-parallel Fabry-Perot, Fig. 1.2a, the wave must be in phase after traversing the length of the cavity to and fro, which gives the condition 2kL = 2πn, or with k = 2π/λ the cavity must be an integer number of half-wavelengths

      (1.41)

      and the frequency is given by

      (1.42)

      To avoid diffraction losses one must make the plane-parallel resonator of infinite transverse dimensions. By choosing focusing mirrors the energy can be confined to the interior of the Fabry-Perot. The basic mode acquires a Gaussian distribution in the transverse direction and is focused to a beam waist of radius a inside the cavity; see Fig. 1.2b. Here the transverse intensity distribution is given by

      (1.43)

      There are also higher transverse modes with more complicated dependence on the transverse variables, see examples in Fig. 1.3. For most applications we can assume our discussion to concern the area r a, where the transverse variation (1.43) may be neglected. Then we can choose the cavity modes to be determined by (1.42) in the form

      (1.44)

      with kn = Ωn/c. In addition, two polarization directions are possible. For laser light one polarization is often selected by Brewster windows or polarizers in the apparatus.

      A relaxation rate like that in (1.30) can be introduced by a passive absorbing medium homogeneously distributed in space. Even losses in the metallic boundaries can be approximated by a decay time τ. This includes mirror transmission losses for a laser cavity. In addition, the optical cavity has diffraction losses due to the finite radii of the mirrors. This can also be represented by a time τ. The loss rate depends on the mode structure; the higher transverse modes, especially, have more intensity at the edges and hence experience higher losses. It is often useful to let the loss parameter depend on the mode and set τn. For a more detailed discussion of optical resonance modes see the literature listed in the final section of this chapter.

      Fig. 1.2 (a) The ideal Fabry-Perot resonator with plane mirrors. The wave pattern repeats itself after one round trip over the distance 2L. (b) A real Fabry-Perot usually employs spherical mirrors with a Gaussian transverse intensity distribution. The beam waist a is defined at the most focused cross section.

      Fig. 1.3 The transverse modes of an optical cavity contain one, two or more dark points across the transverse section of the beam.

      1.3. THE QUANTUM DESCRIPTION OF MATTER

      The matter subjected to laser radiation consists of atoms and molecules. Their spectrum consists of a set of bound states having discrete energy eigenvalues. These are the solutions of the time-independent Schrödinger equation with the Coulombic potential (1.12). In addition, there is a continuum of free ionic states, which are not discussed in this treatise; see Fig. 1.4. Electronic transitions between the discrete states of ions can be handled in the same way as those of neutral particles.

      The atomic Hamiltonian is written

      (1.45)

      Fig. 1.4 The bound atomic states become denser toward the ionization limit, above which the states form a continuum.

      In laser spectroscopy we subject this to monochromatic radiation at a few frequencies {Ωi} only. These cause transitions between levels Ei and Ef if the Bohr resonance condition

      (1.46)

      is approximately satisfied. As atomic level spacings usually are unequal, we can achieve resonance with only a few transitions; the rest are well out of resonance. Then it is possible to truncate the Hamiltonian and look at only those levels that participate in the resonance interaction. In laser spectroscopy the matter is often assumed to be a two-, three-, or N-level system. The transitions into the ionized continuum are neglected in this book.

      The time development described by quantum mechanics is unitary, which means that the probability is conserved. If the atomic state is expanded in the energy eigenstates of the Hamiltonian (1.45)

      (1.47)

      unitarity guarantees that

      (1.48)

      for all times. When we work in a truncated part of the state space only, no unitarity applies. If we consider levels 1 and 2 of Fig. 1.5, they may decay spontaneously or by some other mechanism to levels a and b, which remain unobserved. Such a decay is usually exponential and can be described by decay rates added phenomenologically to the Schrödinger equation of the system. From

      (1.49)

      we find that the coefficients Cn(t) satisfy the equations

      (1.50)

      where a relaxation term is added. The coefficients Hnm are the matrix elements of the total Hamiltonian in the representation given by the energy eigenfunctions of the atomic Hamiltonian (1.45)

      (1.51)

      Fig. 1.5 When states 1 and 2 decay with a constant probability per unit time to unobserved states a and b, the decay can be described by an exponential decrease of the probabilities to be in states 1 and 2 with time.

      In (1.50) we have added the decay rates in such a way that they give the solution

      (1.52)

      if no coupling between the states |ϕn〉 occurs. It is obvious that the Eqs. (1.50) do not preserve the probability (1.48). A spontaneous decay between coupled levels, 1 and 2 in Fig. 1.5 say, cannot be described by relaxation terms in the Schrödinger equation.

      For an atom the problem is centrally symmetric, and the angular momentum is a good quantum number. In the position representation the eigenfunctions of the energy then become of the form

      (1.53)

      ²l(l + 1) and lz m. Rn(r) is the radial wave function, which is known analytically for simple cases like, for example, the hydrogen atom. For this the energy depends only on the principal quantum number, whereas for more complicated atoms the degeneracy with respect to l is lifted, but each level has (2l + 1)-fold degeneracy due to the quantum number m. This is lifted by strong external fields, the Stark and Zeeman effects. The spectrum of a simple model atom is shown in Fig. 1.6.

      For real atoms the spectrum is complicated by the occurrence of electronic and nuclear spins, spin-orbit coupling, and other effects. These features make it necessary to specify the levels with care, but do not change the picture of a set of discrete levels acted upon by the radiation field. For molecules the situation becomes even more complicated; vibrational and rotational transitions are to be added.

      The coupling of matter to the electromagnetic field must

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