Entanglement and Decoherence in a Trapped-Ion

Quantum Register

by

David Kielpinski

B.A., University of Chicago, 1996

A thesis submitted to the

Faculty of the Graduate School of the

University of Colorado in partial fulfillment

of the requirements for the degree of

Doctor of Philosophy

Department of Physics

2001
This thesis entitled:

Entanglement and Decoherence in a Trapped-Ion Quantum Register

written by David Kielpinski

has been approved for the Department of Physics

Dr. David Wineland

Prof. Eric Cornell

Date

The final copy of this thesis has been examined by the signatories, and we find that both the

content and the form meet acceptable presentation standards of scholarly work in the above

mentioned discipline.
iii

Kielpinski, David (Ph.D., Physics)

Entanglement and Decoherence in a Trapped-Ion Quantum Register

Thesis directed by Dr. David Wineland

This thesis describes the construction of a small quantum register using laser-cooled 9Be+

ions in a linear RF trap. Each qubit corresponds to a single trapped ion, with the two qubit

levels being two hyperfine sublevels of the ground electronic state. I outline the basics of ion

trapping and cooling and describe the operation of single-qubit logic gates using Raman transitions. I then review work performed during my tenure. We have realized an entangling gate

and have produced entangled states of up to four ions. This gate enables universal quantum

logic on two qubits in our system. We investigated several applications of this two-qubit register. We demonstrated a violation of a Bell inequality that closes the detector loophole. This

experiment was the first to perform a complete set of Bell correlation measurements on massive

particles. We also demonstrated rotation angle estimation with precision better than the standard quantum limit (SQL) using a two-ion entangled state. As a special case, we performed

Ramsey spectroscopy at precision better than the SQL, opening a possibility for improvement

of atomic clocks. We observed a number of decoherence mechanisms in our register, the most

prominent being magnetic field fluctuations and heating of the ion motion. We propose a way to

eliminate the effects of heating. Finally, we implemented a decoherence-free quantum memory

that was shielded from the effects of magnetic field fluctuations. Encoding one qubit’s worth

of information in the decoherence-free subspace of two ions increased the memory lifetime by a

factor of three under ambient conditions.
Dedication

To my mother, Andrea Kielpinski Sadler, Ph.D.,

my aunt, Eva Parrucci,

my late grandmother, Florence Kielpinski

and especially, my late grandfather

Francis S. Kielpinski

whose affection and encouragement inspired this work
v

Acknowledgements

The credit for this thesis redounds ultimately to the many people who patiently taught

me to do experiments, even when I apparently had bananas in my ears. Special thanks go to

Dave Wineland and Chris Monroe, my two amiable, approachable, and knowledgeable advisors.

Thanks also to Brian King, for answering even the goofiest of first-year grad student questions,

Volker Meyer, for initiating me into the mysteries of German humor, Wayne Itano, for his

amazing ability to communicate with both computers and me, and Cass Sackett, the nicest

supergenius I know. The contributions of my labmates are too numerous to describe here: it

is a pleasure to acknowledge Amit Ben-Kish, Jim Bergquist, Didi Leibfried, Chris Myatt, Mary

Rowe, Quentin Turchette, and Chris Wood. And I fully expect this thesis to pale in comparison

to that of Joe Britton.
Contents

Chapter

1 Introduction 1

2 Trap Apparatus 5

2.1 Linear RF Trap ..................................... 5

2.1.1 Single-Ion Dynamics (Classical) . . ...................... 6

2.1.2 Single-Ion Dynamics (Quantum) . . . . . . . . . . . . . . . . . . . . . . . 10

2.1.3 Multiple-Ion Dynamics . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 13

2.2 Electrode Structures for RF Traps . . . . . . . . . . . . . . . . . . . . . . . . . . 19

2.2.1 Micromachined Linear Trap . . . . . . . . . . . . . . . . . . . . . . . . . . 19

2.2.2 Conventional Paul Trap . . . . . . . . . . . . . . . . . . . . . . . . . . . . 20

2.2.3 Silicon Linear Trap . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 22

2.3 RF Resonators . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 29

2.4 Vacuum System . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 34

2.5 Ion Source . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 35

3 Laser-Ion Interactions 37

3.1 Atomic Structure of 9Be+ . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 37

3.2 Logic State Initialization and Readout . . . . . . . . . . . . . . . . . . . . . . . . 38

3.3 Coherent Coupling of Logic States . . . . . . . . . . . . . . . . . . . . . . . . . . 46

3.3.1 Three-Level Model of Raman Transitions . . . . . . . . . . . . . . . . . . 46
vii

3.3.2 Coupling to the Ion Motion . . . . . . . . . . . . . . . . . . . . . . . . . . 50

3.3.3 Raman Transitions with Multiple Ions . . . . . . . . . . . . . . . . . . . . 53

3.3.4 Cooling to the Ground State . . . . . . . . . . . . . . . . . . . . . . . . . 54

3.3.5 Raman Transitions in 9Be+ . . . . . . . . . . . . . . . . . . . . . . . . . 56

4 Laser Apparatus and Experiment Control 60

4.1 Laser Sources at 626 nm . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 60

4.2 Resonant Doubling to 313 nm . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 66

4.3 UV Beam Control . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 69

4.3.1 BD and RD Beams . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 69

4.3.2 Raman Beams . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 74

4.4 RF Generation and Control . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 78

4.5 Detection Optics . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 80

4.6 Experiment Control and Data Acquisition . . . . . . . . . . . . . . . . . . . . . . 82

5 Multiparticle Entanglement 86

5.1 Theory of the Entangling Gate . . . . . . . . . . . . . . . . . . . . . . . . . . . . 87

5.2 Experimental Results . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 92

6 Applications of Entanglement 103

6.1 Violation of a Bell Inequality . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 103

6.2 Interferometry below the Standard Quantum Limit . . . . . . . . . . . . . . . . . 111

6.2.1 N Views of Ramsey Spectroscopy . . . . . . . . . . . . . . . . . . . . . . 111

6.2.2 Interferometry with Entangled States . . . . . . . . . . . . . . . . . . . . 116

6.2.3 Squeezing Scheme . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 117

6.2.4 Parity Scheme . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 121

6.2.5 Sub-Shot-Noise Ramsey Spectroscopy . . . . . . . . . . . . . . . . . . . . 123
viii

7 Decoherence of the Quantum Register 128

7.1 Dephasing of Internal States . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 128

7.2 Single-Qubit Gate Errors . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 132

7.2.1 Characterization of Gate Errors . . . . . . . . . . . . . . . . . . . . . . . . 132

7.2.2 Spontaneous Emission . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 135

7.2.3 Ion Motion . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 138

7.2.4 Technical Noise . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 139

7.3 Errors in the Entangling Gate . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 140

7.4 Heating of the Ion Motion . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 143

7.4.1 Heating Measurements and Interpretations . . . . . . . . . . . . . . . . . 145

7.4.2 Sympathetic Cooling for Quantum Logic . . . . . . . . . . . . . . . . . . . 150

8 A Decoherence-Free Quantum Memory 160

8.1 Decoherence-Free Subspaces . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 160

8.2 Experimental Implementation of Encoding . . . . . . . . . . . . . . . . . . . . . . 163

8.3 Decoherence of Bare and DFS States . . . . . . . . . . . . . . . . . . . . . . . . . 169

Bibliography 176
Tables

Table

2.1 Scaled positions ui for two, three, and four ions. . . . . . . . . . . . . . . . . . . 14

2.2 Normal mode eigenvalues ξk and eigenvectors ~v(k) for ion motion along ˆz, for two,

three, and four ions. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 17

2.3 Spacing between wafers and thickness of evaporated gold layers for the linear traps. 20

6.1 The correlation values and resulting Bell’s signals for five runs of the experiment.

The phase angles α1, δ1, β2, γ2 are given by Eq. (6.3). The statistical errors are

0.006 and 0.012 for the values of q and B respectively. . . . . . . . . . . . . . . . 106

6.2 Figure of merit ξR for SSN estimation of θ as a function of squeezing angle α.

The loss of precision probably arises from imperfect preparation of the SSS. . . . 122

6.3 The sequence of operations implementing SSN Ramsey spectroscopy. Operations

0 through 2 prepare the nonclassical input state, while operations 3 through 5

perform Ramsey spectroscopy on the input state. We define the zero of RF phase

such that φ1 + φ2 = 0 for pulse 1. . . . . . . . . . . . . . . . . . . . . . . . . . . . 124
x

8.1 The sequence of operations used to demonstrate DFS encoding. Rotations by

an angle θ are written Ucar(θ). “Block” refers to a group of operations that

jointly perform a quantum logic task. A Ramsey block performs a π/2 rotation

on ion 2 while leaving ion 1 alone. The encoding and decoding blocks map the

information in ion 2 into and out of the DFS spanned by |Ψ−i, |Ψ+i. We specify

some intermediate register states for clarity. . . . . . . . . . . . . . . . . . . . . . 165
Figures

Figure

2.1 Electrode structure of a linear RF trap. A common RF voltage is applied to rods

1 and 3, while rods 2 and 4 are held at RF ground. Rods 1 and 3 are held at DC

ground, while the “endcap” segments of rods 2 and 4, labeled with a “+”, are

held at positive DC potential with respect to the middle segments of those rods. 7

2.2 Stability diagram for RF confinement of a charged particle in a linear trap. The

shaded region indicates the values of a and q for which the trapping is stable

along ˆx. ......................................... 9

2.3 A string of four ions in the linear trap. . . . . . . . . . . . . . . . . . . . . . . . . 15

2.4 Schematic of a micromachined linear trap. . . . . . . . . . . . . . . . . . . . . . . 21

2.5 Electrode structure of a ring-and-fork Paul trap. . . . . . . . . . . . . . . . . . . 22

2.6 Schematic of the proposed “quantum CCD” architecture for a large-scale ion-trap

quantum computer. The logic ions, which encode the quantum information, are

held in a storage region. To perform a logic operation, we shuttle the relevant

logic ions into an accumulator region, where they interact with lasers. Auxiliary

ions sympathetically cool the logic ions (see Sec. 7.4.2). . . . . . . . . . . . . . . 24

2.7 Schematic cross-sectional view of the slot sidewalls produced by the anisotropic

etch. The angles shown result from the crystal lattice structure of the silicon. We

measure slot dimensions as the closest distance between sidewalls. . . . . . . . . 26
xii

2.8 Scanning electron micrographs of a typical electrode surface of an alumina-based

trap (top) and a silicon trap piece after the isotropic etch (bottom). Each image

is ∼ 200μm across. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 27

2.9 Schematic of the completed DC electrode structure. Not to scale. (a) Top view.

Dots indicate glass pieces. (b) Side view. Gray shading: Si pieces. No shading:

glass pieces. Dimensions are thicknesses. . . . . . . . . . . . . . . . . . . . . . . . 28

2.10 Schematic cross-section of the planned trap structure. Not to scale. Gray shading:

Si pieces. Dots: glass pieces. Crossbars are partially cut away for clarity. . . . . . 30

2.11 Schematic of a λ/4 coaxial resonator. The antenna couples power from the RF

generator into the resonator. The high RF voltage developed at the other end of

the resonator drives the trap electrodes. . . . . . . . . . . . . . . . . . . . . . . . 31

3.1 Partial energy level diagram of the 9Be+ ion. The dipole-allowed S ↔ |Pi transitions occur near 313 nm and decay with linewidth of 19.4 MHz. The ground-state

hyperfine levels |F = 2, mF = −2i, |F = 1, mF = −1i are used as qubit logic

states and denoted |↑i and |↓i respectively. The hyperfine and magnetic structure

of the P manifold is not shown. . . . . . . . . . . . . . . . . . . . . . . . . . . . . 39

3.2 Histograms of the number of photons detected from a single ion over 1000 shots

of the experiment. (a) Ion prepared in |↑i for each shot. (b) Ion prepared in |↓i

for each shot. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 41

3.3 Histogram of the number of photons detected from a single ion prepared in a

superposition of |↓i and |↑i over 1000 shots. The solid line is the best fit to the

theoretical count distribution. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 42

3.4 Histogram for four ions prepared in a superposition state. The solid line shows

the best fit to a sum of reference histograms. . . . . . . . . . . . . . . . . . . . . 45

3.5 Level diagram for our simplified model of stimulated Raman transitions. Raman

beam 1 couples |1i to P, while beam 2 couples |2i to P. . . . . . . . . . . . . . . 48
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4.1 Jim says, “I love the smell of laser dye in the morning...” . . . . . . . . . . . . . 61

4.2 Optical design of a homemade laser. Intracavity optical elements are not shown.

Values refer to the Raman (detection) laser. R = radius of curvature, T =

transmission, PZT = piezoelectric transducer. Drawing is not to scale. . . . . . . 63

4.3 Optical design of the Wavetrain SC doubling cavity. PZT = piezoelectric transducer, LBO = lithium borate crystal. . . . . . . . . . . . . . . . . . . . . . . . . . 69

4.4 RF/optical schematic for “noiseater” intensity stabilization. PD = photodetector. 70

4.5 Circuit diagram for the servo used in the noiseater. . . . . . . . . . . . . . . . . . 72

4.6 Schematic of the BD beamline from the doubling cavity to the trap. Frequency

shifts of deflected beams are indicated in parentheses, in units of MHz. . . . . . . 73

4.7 Raman beam geometry for our experiment. Beams are labeled by frequency as B

(the bluer beam) or R (the redder beam) and by the port at which they enter as

1, 2, or 3. Thus the bluer beam entering port 1 is denoted B1. Polarizations are

specified in the atomic basis {σ+, σ−, π} selected by the quantization field. The

trap defines the spatial coordinate system, with the trap axis along ˆz. Detection

and optical pumping beams BD and RD are shown for reference. . . . . . . . . . 75

4.8 Schematic of the BR and RR beamlines from the splitter AOM to the trap. Frequency shifts of deflected beams are indicated in parentheses, in units of MHz.

The center frequency is indicated for the double-pass AOM. The switch arrangement for RR is the same as for BR, but is compressed for clarity. . . . . . . . . . 76

4.9 Schematic of the detection optics. L1 is the objective lens, L2 the reimaging lens. 81

5.1 Relevant level scheme for the two-ion version of the entangling gate. The arrows

represent Raman processes driven by pairs of Raman beams. The two pathways

through the intermediate |↓↑i + |↑↓i state exhibit quantum interference. . . . . . 88
xiv

5.2 Experimental flopping curve for p = 1 (dots) with theoretical prediction (line)

with exponential decay of time constant ∼ 30μs and with ions in the motional

ground state. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 95

5.3 Theoretical flopping curves for (a) p = 2 and (b) p = 1.6, calculated for the same

conditions as in Fig. 5.2. The value of p dramatically affects the shape of the

flopping curve. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 96

5.4 Histogram of photon counts obtained after application of the entangling gate. . . 98

5.5 Interference data for (a) two and (b) four ions. Dots are data points, solid lines

are sinusoidal fits with frequencies fixed by Eq. (5.15). The fit amplitudes show

|ρ↓N,↑N | = 0.39 ± 0.01 for two ions and 0.22 ± 0.02 for four ions. . . . . . . . . . . 100

6.1 Histograms of fluorescence signals for the four sets of phases given in Eq. (6.3).

The vertical arrows break up the data into cases of zero, one, or two ions bright

with 98% accuracy. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 107

6.2 Equivalence of Ramsey spectroscopy and Mach-Zehnder interferometry. In the

Schwinger representation of the optical modes, 50/50 beamsplitters are π/2 rotations, while free propagation produces rotations about ˆz. . . . . . . . . . . . . . . 113

6.3 Estimating rotation angles on the Bloch sphere. Rotation of the coordinate system is equivalent to rotation of the atomic spin vector. . . . . . . . . . . . . . . . 115

6.4 Using a CSS (a) and an SSS (b) to estimate the rotation angle θ. The squeezing

of the uncertainty ellipse in (b) permits improved determination of θ. . . . . . . . 118

6.5 Mapping out the uncertainty ellipse of an SSS. The minimum value of ∆Jz(φ)

corresponds to the variance ∆Jsqz along the squeezed direction. . . . . . . . . . . 120

6.6 Measurements of δθ(φ) as a function of the rotation phase φ for the squeezing

scheme with r = 10. The data (dots) follow the ideal error (dashed curve) fairly

well. Error bars for the data are smaller than the marker size. . . . . . . . . . . . 121
xv

6.7 Mean and variance of the parity operator Π in the parity scheme for SSN interferometry. (a) Experimental data, 10 000 shots per point. Points give the mean

value of Π as a function of φ, error bars give the variance. The curve is the best

fit to the mean of Π. (b) Theoretical values for perfect state preparation and

measurement. The curve is the ideal value of the mean as a function of φ, and

error bars give the ideal variance. . . . . . . . . . . . . . . . . . . . . . . . . . . . 124

6.8 Measurement uncertainty δφ derived from the data of Fig. 6.7(a). For certain

values of φ the uncertainty drops below the SQL, demonstrating SSN interferometry.125

6.9 Frequency uncertainty TRδω as a function of TR. The uncertainty drops below

the SQL for TR such that |hΠi| ∼ 0, demonstrating SSN Ramsey interferometry. 127

7.1 A typical heating measurement. The main graph shows nCOM vs. the delay

time t. Insets show the population P↓ remaining in |↓i after the Raman drive.

The data are for linear trap #2 with secular frequency 5 MHz. The data show

Γheat = 12 ± 2 quanta/ms. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 144

7.2 Heating measurements for trap #2 (squares) and trap #3 (triangles) at several

values of ωCOM. Γheat approximately scales as ω−2

COM. . . . . . . . . . . . . . . . . 145

7.3 Schematic of the double Paul trap. The distance between the two rings is 1.7 mm. 146

7.4 Heating measurements for trap PL (circles) and trap PS (squares). (a) First

dataset, showing a scaling Γheat ∝ d(3.8±0.6). (b) Second dataset, showing a

scaling Γheat ∝ d(12±2). . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 147

7.5 Heating measurements for trap P1 over a period of two years. The solid line is a

guide to the eye, but indicates a distinct increase in heating rate with trap age.

The trap remained under vacuum for the entire period shown. . . . . . . . . . . . 149

7.6 Normalized axial mode frequencies as a function of μ for (a) 3, (b) 5, (c) 7, and

(d) 9 ions. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 153
xvi

7.7 Normalized heating rates for the axial modes as a function of μ for (a) 3, (b) 5,

(c) 7, and (d) 9 ions. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 154

7.8 Trap anisotropy at the zigzag instability as a function of μ for 3, 5, 7, and 9 ions.

Arrows indicate the crossover point between two distinct zigzag instabilities. . . 156

7.9 Normalized frequencies of the transverse modes as a function of μ with 2 =

1.120(μ) for (a) 3, (b) 5, (c) 7, and (d) 9 ions. . . . . . . . . . . . . . . . . . . . . 158

7.10 Normalized heating rates of the transverse modes as a function of μ with 2 =

1.120(μ) for (a) 3, (b) 5, (c) 7, and (d) 9 ions. . . . . . . . . . . . . . . . . . . . . 159

8.1 Power spectral density of the B2 intensity as determined from the photodiode

signal. The spectrum is flat to ∼ 100 kHz and drops off rapidly for higher frequencies, as desired. The measurement time of 5 ms limits the spectral resolution. 171

8.2 Decay of the DFS state (circles) and the bare state (crosses) under engineered

collective dephasing. The noise is applied for a fraction of the delay time of about

25 μs between encoding and decoding. Coherence data are normalized to their

values for zero applied noise. The fit lines are exponential decay curves, giving

decay rates of 0.18(1) μs−1 for the bare state and 3.5(5.0) × 10−3 μs−1 for the

DFS state. The bare data point for zero applied noise is excluded from the fit

(see text). . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 171

8.3 Decay of the DFS state (crosses) under application of B2 for up to 100 μs. The

noise is applied for a fraction of the delay time of 200 μs between encoding and

decoding. Coherence data are not normalized. The line is the best fit to a

Gaussian decay curve. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 173
xvii

8.4 Decay of DFS-encoded state (circles) and test state (crosses) under ambient decoherence. We vary the delay time between encoding and decoding to give the

ambient noise a variable time to act. Coherence data are normalized to their values for zero applied noise. The fit lines are exponential decay curves for purposes

of comparison and are not theoretical predictions. The decay rate from the fit is

7.9(1.5)×10−3 μs−1, for the test state and 2.2(0.3)×10−3 μs−1 for the DFS state.

Since the coherence time of the DFS-encoded state is much longer than that of

the test state, we see that the chief source of ambient decoherence is collective

dephasing. . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 175
Chapter 1

Introduction

This thesis describes the construction of a small quantum register using laser-cooled

9Be+ ions in a linear RF trap. The register is composed of elementary two-level systems,

“qubits,” which are analogous to single bits in a classical computer. Each qubit corresponds to

a single trapped ion, with the two qubit levels being two hyperfine sublevels of the ground electronic state. The quantum state of the whole register encodes information, and an appropriate

unitary evolution of the state of the register can perform a computing task.

Here we demonstrate the ability to transform an arbitrary quantum state of two qubits

into any other quantum state. In the language of computing, we can perform universal quantum

logic on two qubits. A universality construction similar to that of classical computer science

guarantees that the operations demonstrated here suffice for universal quantum logic on a large

quantum register as well. In some sense, then, only technical difficulties stand between us and a

full-scale quantum computer. For some tasks, quantum algorithms are much more efficient than

any known classical algorithm, in particular for factoring of large numbers [1]. To factor an Ndigit number, a classical computer requires resources that are exponential in N, but a quantum

computer only requires resources polynomial in N. Because most modern cryptography relies

on the difficulty of factoring large numbers, a full-scale quantum computer could have a large

impact on many areas of technology, Internet commerce being only one example.
2

We used the quantum register to study two central phenomena in quantum information:

entanglement and decoherence. Writing the two states of a qubit as |↓i and |↑i, we see that

the two-qubit state |↓↑i − |↑↓i is entangled: we cannot decompose this state into a product

|Ψ1i|Ψ2i of two single-qubit states. While quantum logic operations on single qubits have been

commonplace for a long time under a variety of names, controlled manipulation of entanglement

has only become possible in the last few years. Entanglement is thought to be closely related to

the efficiency of quantum computers. A unique feature of our experiment is that we can generate

and manipulate entangled states of two qubits on demand, so that we have complete control

over the behavior of the register. No other quantum computing experiment has demonstrated

this ability to date.

If entanglement provides the power of quantum computing, decoherence takes it away

again. Interactions of our quantum register with the environment disturb the register state,

causing transfer of the information in the register to the environment. From the quantum information point of view, decoherence is just this loss of information from the register. A wide

variety of physical processes cause decoherence, and methods of eliminating decoherence will be

essential for large-scale quantum computers.

The outline of the thesis is as follows. Chapters 2, 3, and 4 describe the basics of the

experimental apparatus and the underlying theory. Chapter 2 explains the physical principles

behind ion trapping, from both a classical and a quantum perspective, and the dynamics of

laser-cooled ion crystals in traps. I then describe the construction of our ion traps and the

apparatus needed to operate them. In Chapter 3, I develop the theory of coherent manipulation
3

of a single 9Be+ ion by stimulated Raman transitions and explain how to cool ion crystals to

the ground state of their quantized motion. Chapter 4 describes the laser apparatus used for

initialization, manipulation, and detection of the register state.

The rest of the thesis reviews work done during my tenure as a graduate student. In

Chapter 5, I describe the method we used to create entangled states of the register and present

data on entangled states of two and four qubits. The two-qubit entangling operation, in conjunction with the single-qubit operations discussed in Chapter 3, suffices for universal quantum logic.

Chapter 6 presents two applications of the two-qubit register. We demonstrated a violation of

a Bell inequality with no detector loophole. This experiment was the first to perform a complete set of Bell correlation measurements on massive particles. We also demonstrated rotation

angle estimation with precision better than the standard quantum limit (SQL) using a two-ion

entangled state. As a special case, we performed Ramsey spectroscopy at precision better than

the SQL, opening a possibility for improvement of atomic clocks. Chapter 7 describes the major

decoherence mechanisms affecting the quantum register.1 I present a detailed analysis of gate

errors for the entangling gate. The errors mostly arise from heating of the ion motion, and I

discuss the data on heating gathered so far. We have no definite conclusion on the source of

heating, but I describe a way to eliminate its effects through sympathetic cooling of the ion

crystal. In Chapter 8, I present an experiment on eliminating a particular class of decoherence

processes through encoding of quantum information into a decoherence-free subspace. Rather

than attacking the physical mechanisms creating decoherence, we encoded the information so as

to decouple the register state from the environment. The encoded information was completely

unaffected by the class of decoherence processes we targeted, and survived much longer than the

unencoded information under ambient conditions. The useful properties of the encoding suggest

that some similar encoding will be commonly used in large-scale quantum computing.

1 While some of the work in Chapter 7 predates me, our understanding of decoherence has advanced quite a

bit over the course of my studies. As a general rule, any information not found in Ref. [2] is new work.
4

Since this work was carried out at the National Institute of Standards and Technology,

SI units are obsessively used throughout.
Chapter 2

Trap Apparatus

The experiments described in this thesis used radio-frequency (RF) traps to confine

9Be+ ions under ultra-high vacuum (UHV). In these traps, one applies large RF voltages to

an electrode structure made of conducting material in order to create a quadrupole electric

field with a minimum in free space. For appropriate RF voltages and frequencies, the RF field

induces a ponderomotive potential that confines 9Be+ ions harmonically at the field minimum.

The ions are then well isolated from environmental perturbations, enabling the precise quantum

state control needed for these experiments.

2.1 Linear RF Trap

Linear RF traps were used for almost all the experiments described here. One common

electrode structure for such a trap [3] is shown in Fig. 2.1. Essentially the trap is a quadrupole

mass filter plugged at the ends with static potentials. To operate the trap, one applies RF

voltage to the rods 1 and 3 of Fig. 2.1, while rods 2 and 4 are held at RF ground. The induced

ponderomotive potential confines the ions to the RF nodal line, which lies along the ˆz axis. The

“endcap” segments of rods 2 and 4 are held at a positive DC voltage relative to the middle

segments of those rods, pushing the (positive) ions toward the center of the trap. The ˆx axis

lies on the line connecting rods 1 and 3, while the ˆy axis lies on the line connecting rods 2 and
6

4.

It is difficult to calculate the exact electric potential created by this electrode configuration. However, as we will see, the ions are held at minima of the electric field, so the lowest-order

term in the electric field is the quadrupole term. In our experiments, the amplitude of the ions’

motion is so small compared to the distance between trap electrodes that the higher-order terms

in the multipole expansion are negligible [4]. Thus we can write the electric potential in the

region of interest as a sum of RF and DC quadrupole potentials:

φ(~x, t) = 1

2

x2 − y2

R2 κRFV0 cos ΩT t + κDCU0

1

R2

·

z2 − 1

2

(2DCx2 + (1 − 2DC)y2)

 ̧

(2.1)

where V0, U0 are the applied RF and DC voltages, R is the distance from the RF nodal line

to the electrode surfaces, and κRF,DC and 2DC are geometric factors. The frequency of the RF

field, ΩT , is often called the drive frequency, and is expressed here in rad/s.

2.1.1 Single-Ion Dynamics (Classical)

For the present, we model an ion as a classical charged point particle. The equations of

motion for the ion are separable in the {x, ˆ y, ˆ zˆ} basis. Since there is no RF field along ˆz, the

axial potential is purely harmonic with frequency ωz = peκDCU0/m. The Mathieu equation

d2x

dζ2 + [a + 2q cos 2ζ]x = 0 (2.2)

describes the motion along ˆx, where we use dimensionless variables
7

1

3

2

4

+

+

+

+

yˆ xˆ

zˆ

Figure 2.1: Electrode structure of a linear RF trap. A common RF voltage is applied to rods 1

and 3, while rods 2 and 4 are held at RF ground. Rods 1 and 3 are held at DC ground, while

the “endcap” segments of rods 2 and 4, labeled with a “+”, are held at positive DC potential

with respect to the middle segments of those rods.
8

ζ = ΩT t/2 time (2.3)

q = 2eκRFV0

mΩ2

T R2 RF voltage (2.4)

a = −4e2DCκDCU0

mΩ2

T R2 DC voltage (2.5)

Eq. 2.2 also describes the motion along ˆy for x replaced by y, q replaced by −q, and 2DC replaced

by 1 − 2DC.

Since the Mathieu equation (2.2) is periodic in time, we can apply Floquet’s theorem [5]

to find solutions of the form

x(ζ) ∝ e±iβζ f(ζ) (2.6)

with f(ζ + π) = f(ζ). Stable solutions require β = β(a, q) real. Fig. 2.2 shows the regions of

stability in the a − q plane [6]. In particular, the ion motion is stable for |a| < q2 ¿ 1, and

expanding in these parameters to lowest order, we find

x(ζ) ≈ x0eiβ(ζ+ζ0)

·

1 +

1

2

q sin 2ζ

 ̧

(2.7)

β =

μ

a +

1

2

q2

¶1/2

(2.8)

with ζ0 a constant phase. Rewriting Eq. 2.8 in SI units gives

x(t) ≈ x0ei(ωt+φ)

·

1 +

eκRFV0

mΩT R2 sin ΩT t

 ̧

(2.9)

ω = 1

2

βΩT (2.10)

with φ a constant phase. Eq. (2.9) describes harmonic motion of the ion at the “secular”

frequency ω, modulated by the much smaller “micromotion” term at the drive frequency ΩT .
9

0.2 0.4 0.6 0.8 1 1.2 1.4

0.5

0.5

1

a

stable q

unstable

unstable

Figure 2.2: Stability diagram for RF confinement of a charged particle in a linear trap. The

shaded region indicates the values of a and q for which the trapping is stable along ˆx.

Again ω and ΩT are expressed in rad/s. If we neglect the micromotion and the effects of the

DC field, the ion moves in an effective harmonic potential with secular frequency

ωRF = eκRFV0

√2mΩT R2 (2.11)

Although ωRF formally diverges for small ΩT , so do a and q, so that the approximations leading

to Eq. (2.11) become invalid. If we take ΩT , V0, and U0 to zero in such a way that a and q

remain constant, we find V0 ∝ Ω2

T , so that ωRF actually goes to zero for both large and small

ΩT . To get high secular frequencies, the strategy is to apply as large an RF voltage as possible,

then choose the smallest ΩT such that the ions remain stably trapped. Though Fig. 2.2 suggests that the ion motion remains stable for q ≈ 0.8, our experience was that reliable long-term

confinement of a few ions requires q . 0.3.

Real ion traps are frequently subject to ambient electric fields. These fields arise, for

instance, from charging of insulating material used in trap construction. Usually the amplitude

of ion motion is small enough that we can neglect the spatial variation of the ambient field. Then
10

the ion feels a constant force Fx, which we take along ˆx, and including this force in Eq. (2.2),

we find the new trajectory

x(ζ) ≈ (x0eiβ(ζ+ζ0) + x1)

·

1 +

1

2

q sin 2ζ

 ̧

(2.12)

where x1 = 4Fx/(mω2). The field displaces the ion from the center of the quadrupole potential

and increases the amplitude of micromotion. For Doppler-cooled ions, the amplitude of secular

motion x0 is typically less than 100 nm, so even very small ambient fields (on the order of 100

V/m) can significantly increase the micromotion amplitude.

2.1.2 Single-Ion Dynamics (Quantum)

Quantizing the ion motion along the trap axis ˆz is straightforward. For Doppler-cooled

ions, the anharmonicity of the DC potential is very small and the usual harmonic oscillator quantization holds. The energy eigenstates are equally spaced with level spacing equal to the axial

trap frequency ωz, and the ground state of the axial motion has wavepacket spread p~/(mωz).

The classical dynamics along the transverse directions ˆx, yˆ is rigorously described by the

Mathieu equation (2.2) and so appears difficult to quantize. However, in the usual regime of

small a and q, we can neglect the micromotion to a first approximation and simply treat the potential as harmonic with frequency ω. Quantizing then yields the standard harmonic oscillator

spectrum for the motional energy eigenstates along ˆx and ˆy. In a semiclassical extension of this

model to include the effects of micromotion, one modulates the position operator by adding a

c-number term xμ cos 2ζ, where xμ is the classical micromotion amplitude.
11

While this semiclassical model accounts for the data presented in this thesis, it leaves

important open questions about the full quantum dynamics. For instance, it is not a priori

obvious whether there is a true ground state of the ion motion. Therefore we will briefly outline

the quantum dynamics in the regime of small a and q.

In the Heisenberg picture, the position operator x evolves according to Eq. (2.2) [7] and

we make the ansatz [8]

x(ζ) = r 1

2β [s(ζ)a† + h.c.] (2.13)

s(ζ) = eiβζ ·

1 +

1

2

q sin 2ζ

 ̧

(2.14)

where a is a time-independent operator satisfying the harmonic oscillator commutation relations [a, a†] = 1. Since s(ζ) solves the classical Mathieu equation to lowest order in a and q,

ansatz (2.13) solves the quantum Mathieu equation in the same approximation.

We can rewrite the equation of motion (2.13) using the time evolution operator U as

follows:

x(ζ) = U†x(0)U (2.15)

U = exp μ1

2

r(ζ)

¡

a2 − a†2¢

¶

exp(−iβa†aζ) (2.16)

where r(ζ) = −(q/2) sin 2ζ and we have set ~ = 1. Using i(dU/dζ) = HU, we find the effective

Hamiltonian

H(ζ) = βa†a − 1

2

iq(a2 − a†2) cos 2ζ (2.17)
12

to lowest order in a, q. The first term in this Hamiltonian drives the secular oscillation at

frequency β, while the second term describes squeezing of the ion wavepacket [9].

The evolution operator (2.16) has the form U(ζk) = exp(−iβa†a) for times ζk = kπ/2,

k an integer. If we take snapshots of the dynamics at the times ζk, the evolution appears to be

that of a simple harmonic oscillator. At times ζ 6= ζk, the motional state will be squeezed by

the factor r(ζ). The bandwidth of measurements on the ion motional state is usually much less

than ΩT . In this case, the fractional change in expectation value of an observable O due to the

squeezing is

δsqz(O)=1 −

R π

0 hO(ζ)i dζ

hO(0)i (2.18)

and we find δsqz(x) = 0, δsqz(x2) = O(q2). Thus, to first order in q, we can consider the quantized ion motion to be that of a simple harmonic oscillator of frequency β.

Our quantum treatment has so far assumed that the ion sits at the RF node, so that the

quantum version of Eq. (2.2) describes the evolution of the position operator x. We can readily

include the effect of a uniform force Fx, such as that arising from an ambient electric field, in

our quantum description. In this case x(ζ) evolves according to

d2x

dζ2 + [a + 2q cos 2ζ]x = fx (2.19)

where fx is the normalized force 4Fx/(mΩ2

T ), and we make the ansatz

x(ζ) = r 1

2β [s(ζ)a† + h.c.] + fxβ2

·

1 +

1

2

q sin 2ζ

 ̧

(2.20)

where the first term solves the homogeneous quantum Mathieu equation and the second term
13

is a c-number solution to the inhomogeneous classical Mathieu equation. Thus we can interpret

the ion motion as composed of a quantum state evolving according to Eq. (2.16), but with its

x-coordinate displaced by the classical micromotion of amplitude x1 = 4Fx/(mω2). The semiclassical treatment of micromotion given above therefore neglects only the micromotion-induced

squeezing of the ion wavepacket, which typically affects measurements only in order O(q2).

2.1.3 Multiple-Ion Dynamics

When multiple ions are present in the trap, we must consider the Coulomb repulsion

between ions as well as the ions’ interaction with the harmonic trapping potential. If the ions

are sufficiently cold, the equilibrium positions of the ions are given by minimizing the potential

energy. The possible equilibrium configurations have been extensively studied experimentally

[10, 11, 12] and theoretically [13, 14]. For sufficiently weak axial confinement, the equilibrium

positions all lie on the trap axis x = y = 0, so that the ions line up in a string, as shown in

Fig. 2.3. In this configuration, the ions all have the same (small) amplitude of micromotion. As

we will see in Section 5, this fact is essential to operation of our entangling gate. The string

configuration was used for all the experiments described in this thesis. The positions of the ions

along z are given for the string by minimizing the potential [15, 16, 17]

V (z1,...zN ) = 1

2

mω2

z

X

N

i=1

z2

i +

e2

8π20

X

N

i,j=1

i6=j

1

|zi − zj | (2.21)

where z1,...,zN are the equilibrium positions for a string of N ions and ωz the axial trapping

frequency. Defining normalized positions as ui ≡ zi/` with `3 ≡ e2/(4π20mω2

z ), we find a coupled

set of equations for the ui

ui −X

i−1

j=1

1

(ui − uj )2 + X

N

j=i+1

1

(ui − uj )2 = 0, i = 1 ...N (2.22)
14

Table 2.1: Scaled positions ui for two, three, and four ions.

N ui

2 −(1/2)2/3 (1/2)2/3

3 −(5/4)1/3 0 (5/4)1/3

4 −1.437 −0.454 0.454 −1.437

These equations have analytic solutions only for N ≤ 3. Numerical solutions for N ≤ 10 have

been calculated by Steane [15] and James [16]. For two, three, and four ions, the cases of interest

here, the scaled positions are given in Table 2.1. Here the scaling factor ` = 7.25 ν−2/3 z μm for

9Be+ , with νz the axial trap frequency in MHz (not in rad/s).

The full dynamics for multiple ions are quite complicated, and in general are chaotic

[18, 19, 20] due to the nonlinearity of the Coulomb force. However, Doppler-cooled ions have

a motional amplitude of less than 100 nm and typical ion spacings are several μm, so we can

linearize the Coulomb force. The ions then undergo coupled harmonic oscillations about their

equilibrium positions. The potential is separable in {x, ˆ y, ˆ zˆ} once we linearize the Coulomb force.

We first consider oscillations along ˆz. Writing the displacement of the ith ion as qi(t) = zi(t)−`ui

yields the quadratic Lagrangian [16, 17]

L = m

2

X

N

i=1

q ̇

2

i − 1

2

X

N

i,j=1

∂2V

∂zi∂zj

 ̄

 ̄

 ̄

 ̄

{qi}=0

qiqj (2.23)

= 1

2

mω2

z





X

N

i=1

μdqi

dT ¶2

− X

N

i,j=1

Aij qiqj



 (2.24)
15

200 μm

Figure 2.3: A string of four ions in the linear trap.
16

Aij =







1+2X

N

k=1

k6=i

1

|ui − uk|

3 i = j

−2

1

|ui − uj |

3 i 6= j

(2.25)

where T = ωzt is a normalized time. To find the normal modes, we solve the eigenvalue equation

X

N

j=1

Aijv(k)

j = ξ2

kv(k)

i (2.26)

for the frequencies ξk and (orthonormal) eigenvectors ~v(k) of the N normal modes. In terms of

the physical time t, the frequency of the kth mode is ωzξk. If the kth mode is excited with an

amplitude C, the ith ion moves according to

qi(t) = Re h

Cv(k)

i ei(ξkωzt+φk)

i

(2.27)

where φk is a constant phase. For two, three, and four ions the normal mode eigenvectors and

eigenvalues are given in Table 2.2. The lowest-frequency mode is always the center-of-mass

(COM) mode, in which the ion string moves as a unit, with no relative motion between the ions.

The Coulomb interaction then has no effect on the dynamics of the COM mode, so the COM

frequency is just equal to the single-ion trap frequency ωz. Because of the symmetry of the ion

string about z = 0, the ions’ relative amplitudes of motion in a given mode are either symmetric

or antisymmetric about the center of the string, as seen in Table 2.2.

So far we have only considered the N normal modes of oscillation along the trap axis, but

there are 2N more normal modes of the ion string involving displacement along ˆx, yˆ. We will

treat the motion along ˆx; the analysis is similar for the motion along ˆy. The small displacements

xi are governed by the normalized Lagrangian
17

Table 2.2: Normal mode eigenvalues ξk and eigenvectors ~v(k) for ion motion along ˆz, for two,

three, and four ions.

N ξk ~v(k) Name

2 1

√3

1/

√2 1/√2

−1/

√2 1/

√2

COM

Stretch

3 1

√3

p29/5

1/

√3 1/√3 1/√3

−1/

√201/

√2

1/

√6 −2/

√6 1/√6

COM

Stretch

Egyptian

4 1

√3

2.41

3.05

1/2 1/2 1/2 1/2

−0.674 −0.213 0.213 0.674

1/2 −1/2 −1/2 1/2

−0.213 0.674 −0.674 0.213

COM

Stretch 1

Stretch 2

Stretch 3
18

L = 1

2

X

N

i=1

μdxi

dT ¶2

− 1

2

X

N

i,j=1

Bijxixj (2.28)

Bij =







2

2 − 1

2 − X

N

k=1,6=i

1

|ui − uk|

3 i = j

1

|ui − uj |

3 i 6= j

(2.29)

(2.30)

where ε2 − 1/2 ≡ (ωx/ωz)2 and ωx is the secular frequency along ˆx. This peculiar definition

will come in handy later (see Section 7.4.2). Note that the linearized Coulomb terms |ui −uj |

−3

enter into Bij (Eq. 2.29) differently than for Aij (Eq. 2.25) because the ion string extends along

zˆ. In general, the Coulomb interaction also couples motion along ˆx to motion along ˆz, but

for the string configuration the Coulomb forces are all directed along ˆz when the ions are at

rest, decoupling the ˆx motion from the ˆz motion. This decoupling considerably simplifies our

entangling gate (see Section 5.)

The normal mode eigenvalues ξ(x) and eigenvectors ~v(x;k) can now be found by solving

the counterpart of Eq. (2.26). The frequency of the kth mode in terms of the physical time t

is then ωxξ

(x)

k and the motion of the ions is described by the counterpart of Eq. (2.27). Again

we find a COM mode, but in this case the COM mode is not necessarily the lowest-frequency

mode. For α = (ωz/ωx)2 larger than some N-dependent critical value α0, the ion string breaks

into a “zigzag” in which neighboring ions are displaced from the RF node in opposite directions.

The critical value α0 occurs at the smallest value of α for which det Bij = 0, corresponding to

a motional mode along ˆx with zero frequency, i.e. an unstable mode. The zigzag transition has

been observed in detail by Enzer et al. [12], who found good agreement between experimental

and theoretical values for α0.
19

It is easy to quantize the normal modes, since each mode is just a simple harmonic

oscillator. We consider only the axial modes for simplicity. Writing the operator for small

displacements of the ith ion as xi and the conjugate momentum as pi, we define the annihilation

operator ak for the kth mode in the usual way:

ak =

rNmξkωz

2

X

N

i=1

v(i)

k

μ

xi +

i

Nmξkωz

pi

¶

(2.31)

For our purposes it is sufficient to treat the micromotion classically, as described at the beginning

of Section 2.1.2.

2.2 Electrode Structures for RF Traps

2.2.1 Micromachined Linear Trap

Most of the experiments described here used linear RF traps [21, 22] constructed using

micromachined alumina substrates. Each trap consisted of a pair of substrate wafers bonded

together. The wafers were laser-machined and coated with gold on parts of their surfaces to

form electrode structures. The gold coating consisted of a 25 nm sticking layer of titanium

evaporated directly onto the alumina with a layer of gold 0.25 to 0.75 μm thick evaporated onto

the titanium. A schematic of an assembled trap structure is shown in Fig. 2.4. Each wafer was

1.5 cm long (along the direction of the wide slot), 1.0 cm wide, and 125 μm thick. The single

wide slot on each wafer was 2 mm long and 200 μm wide, while the two thin slots were 20 μm

wide and about 0.5 mm long. The edges of the wide slots were bevelled at 45◦, with the narrow

(200 μm) side nearer the ions. The electrode structures on the two wafers are mirror-images of

each other under reflection about the wide slot. The wafers were bonded together using glass

frit paste. To control the distance between the trap wafers, two alumina spacer wafers were

inserted between the trap wafers, one at each end of the trap wafers.
20

Table 2.3: Spacing between wafers and thickness of evaporated gold layers for the linear traps.

Trap Wafer spacing (μm) Gold thickness (μm)

#1 250 0.25

#2 250 0.75

#3 375 0.75

Three such traps were constructed for the experiments described here. The traps varied

in the spacing between wafers and the thickness of the evaporated gold layers, as detailed in

Table 2.3. In each case, gold wires connected the DC electrodes on the pair of trap wafers to

a nearby RF filter board constructed on thicker alumina wafers (10 mm × 7 mm × 0.2 mm).

Surface-mount resistors (1 kΩ) and capacitors (800 pF) were bonded to the RF board to form

low-pass RC filters (corner frequency 200 kHz). These filters provided the RF grounding for the

DC electrodes. The filters also attenuated externally injected electronic noise at the ion motional frequencies. Such noise can lead to ion heating (Section 7.4). For trap #1, two identical

stages of RC filtering were used, while only one stage was used for traps #2 and #3. Since the

RF electrodes were electrically connected to the λ/4 RF resonator in each case (see Section 2.3),

they were DC grounded.

2.2.2 Conventional Paul Trap

Some of the heating data presented in Section 7.4 was obtained using Paul traps [23, 24]

of the ring-and-fork variety [25]. A schematic of such a trap is shown in Fig. 2.5. Applying

an RF voltage VRF across the electrode structure creates a three-dimensional RF quadrupole

potential φRF ∝ VRF(z2 − αx2 − (1 − α)y2)/R2. Here R is a characteristic length scale for the
21

1'

2'

3'

4'

1

2

3

4

IONS

200 m

Traditional

linear trap:

Our trap:

Figure 2.4: Schematic of a micromachined linear trap.
22

X

Y

Z

fork

ring

Figure 2.5: Electrode structure of a ring-and-fork Paul trap.

size and spacing of the trap electrodes and α is a dimensionless parameter depending on the

ellipticity of the ring electrode (α ≈ 0.4 for a circular ring.) By convention, the ˆz axis is taken

perpendicular to the plane of the ring electrode. The RF potential traps ions by the mechanism

described in Section 2.1.1. To tune the secular frequency, one applies an additional DC voltage

VDC across the electrodes, which adds a DC potential φDC ∝ VDC(z2−αx2−(1−α)y2)/R2. The

DC potential alters the secular frequency as described in Section 2.1.1. Details of the particular

Paul traps used in our experiments are given in Section 7.4; typically the trap dimensions were

on the order of a few hundred μm.

2.2.3 Silicon Linear Trap

During my tenure as a student, we attempted to construct a trap from micromachined

silicon. Such a trap has two attractive characteristics: its electrode surfaces are extremely

smooth, and the fabrication procedure should allow us to make extremely complicated electrode
23

structures. The electrode surfaces for the traps described above were generally quite rough on

the scale of a few microns (see Fig. 2.8). The association of heating with surface effects (Sec. 7.4)

suggests that the heating mechanism might depend on this roughness. With wet-etching techniques, one can obtain nearly atomically smooth surfaces, possibly eliminating the heating.

Easy fabrication techniques for complicated electrode structures will become important

for our experiment in the near future. We plan to scale up the small quantum register described

in this thesis by constructing a large array of interconnected ion traps, as schematically depicted

in Figure 2.6. In this “quantum CCD” model, applying voltages to the DC electrodes shown in

the figure allows us to shuttle ions from place to place in the array and pick out small sets of ions

for quantum logic operations. Fabricating these structures by the techniques of Section 2.2.1

will eventually become impractical. However, the lithographic methods developed for silicon

micromachining already routinely produce much more complicated and delicate structures.

The proposed design uses heavily boron-doped h100i-cut Si wafers as the starting point for

trap fabrication. The wafers were 100 μm thick. We used wafers with resistivity ∼ 1.5 × 10−4Ω

m, so the Si was essentially metallic. We planned to stack three Si layers together, rather than

the two layers used in Section 2.2.1), so that the middle layer carried DC voltage only, while

the outer layers carried the RF voltage. Flat pieces of Corning 7070 glass, a borosilicate glass

with low RF loss, was used as an insulating spacer between the layers. We could attach the

Si rigidly to the glass using the technique of anodic bonding, which is commonly used in the

semiconductor industry. To make a bond, we heated the glass and Si pieces to about 450 ◦

C and applied ∼ 1 kV between the Si and the glass for several minutes. The bond formed

by this process was very strong; bonded pieces would break sooner than be pried apart. It

was essential to use highly polished Si and glass pieces. A specification of 60/40 scratch/dig
24

memory ion

cooling ion (shuttling)

quantum register “accumulator”

quantum

register

Figure 2.6: Schematic of the proposed “quantum CCD” architecture for a large-scale ion-trap

quantum computer. The logic ions, which encode the quantum information, are held in a storage

region. To perform a logic operation, we shuttle the relevant logic ions into an accumulator

region, where they interact with lasers. Auxiliary ions sympathetically cool the logic ions (see

Sec. 7.4.2).
25

for the glass and standard wafer polishing for the Si was sufficient. The pieces also had to be

extremely clean. Standard optics-cleaning procedures with spectroscopic-grade methanol were

ineffective. Instead, we cleaned the pieces by rinsing with VLSI-grade methanol, then VLSIgrade isopropanol, then deionized water (∼ 15MΩ cm resistivity).

To test the RF characteristics of our design, we applied up to 10 W of RF power to a test

piece through the external resonator used for the double Paul trap (see Sec. 2.3). The test piece

consisted of two Si pieces spaced by a piece of Corning 7070 glass. The glass was 175 μm thick,

the thickness used in the full design. The resonance width was the same with the test piece as

without it, implying that little power was dissipated in the test piece. In contrast, replacing the

Corning 7070 glass with the more common Corning 7740 (Pyrex) caused severe broadening of

the resonance and heating of the test piece at RF powers of ∼ 1 W.

To construct the DC and RF electrode layers, we wet-etched slots 200 μm wide and 2

mm long into the wafers and diced the wafers into pieces. The DC pieces measured 1 cm square,

while the RF pieces were 1 mm wide and 8 mm long. We used a KOH-based anisotropic etchant

and etched the wafers from both sides. Figure 2.7 shows a schematic cross-sectional view of

the slot sidewalls produced in this way. The slot dimensions above refer to the closest distance

between the slot sidewalls.

After the first etch, we isotropically etched the wafers with hydrofluoric acid, removing

several μm of Si, to soften any sharp edges left by the anisotropic etch. The resulting electrode

surfaces were smooth on the micron scale. Figure 2.8 shows a scanning electron micrograph of

a typical etched electrode surface.
26

54.74o

54.74o

200 μm Si Si

Figure 2.7: Schematic cross-sectional view of the slot sidewalls produced by the anisotropic

etch. The angles shown result from the crystal lattice structure of the silicon. We measure slot

dimensions as the closest distance between sidewalls.

To produce the separate DC electrodes for axial confinement of the ions, we first bonded

the DC piece to two pieces of 7070 glass, making sure that the glass was at least 2 mm away

from the slot edges. We then used a dicing saw to make two diagonal cuts of width ∼ 40 μm

across the DC piece. The cuts were at about 45◦ to the long axis of the slot and intersected at

the geometric center of the slot. The dicing saw cut completely through the Si, but cut only

slightly into the glass below. The glass was positioned so that the four pieces of Si separated by

the cuts were rigidly supported on the glass. The cut sidewalls were rougher than the etched

surfaces, but fields originating at the sidewalls would likely be shielded from the ions by the

rest of the electrode structure, so we felt that the cut roughness would not induce heating (see

Sec. 7.4). After making the saw cuts, we bonded two more pieces of glass to the other side of

the DC piece, and bonded two “crossbar” pieces of 500 μm thick Si across these glass pieces

for structural support. Without the crossbars, the structure was extremely delicate. Figure 2.9

shows a schematic of the completed DC electrode structure.

This project is currently not finished. We produced four DC electrode structures as outlined above, and etched a large number of RF pieces. The design called for bonding one RF
27

Figure 2.8: Scanning electron micrographs of a typical electrode surface of an alumina-based

trap (top) and a silicon trap piece after the isotropic etch (bottom). Each image is ∼ 200μm

across.
28

DC piece 100 μm

175 μm

175 μm

crossbar 500 μm

crossbar

slot

cut

(a)

(b)

Figure 2.9: Schematic of the completed DC electrode structure. Not to scale. (a) Top view.

Dots indicate glass pieces. (b) Side view. Gray shading: Si pieces. No shading: glass pieces.

Dimensions are thicknesses.
29

piece to each side of a DC structure, with all three slits aligned. By making the RF pieces

small, we hoped to avoid the capacitance associated with large structures, which lowers the

drive frequency and resonator Q (see Sec. 2.3). To make electrical connections to the structure,

we evaporated patches of gold onto the DC and RF pieces in the way described in Section 2.2.1.

We were able to robustly gap-weld 25μm thick gold ribbon to these patches. We also intended

to bond another mask layer to one side of the three-layer structure to avoid plating of Be from

the ovens onto the electrodes (see Sec. 2.5). These mask layers were 1 cm square pieces of Si

wafer with an etched slot 100 μm wide and 1 mm long and were to be bonded to the crossbars

with glass. The center of the mask slot was to be aligned to the center of the RF and DC slots.

Figure 2.10 shows a schematic cross-sectional view of the planned trap structure. Assuming the

Be oven was placed 5 cm from the trap, a 1 mm aperture placed in front of the oven would

suffice to keep atoms passing through the mask slot from hitting the electrode surfaces.

2.3 RF Resonators

Attaining high secular frequencies in RF traps requires the application of large RF voltages at the drive frequency, typically several hundred volts for the traps described here. Since

most commercial RF generators are impedance-matched for a 50 Ω load, it is very inefficient

to drive the trap electrodes directly from such a generator. Instead, we used a λ/4 coaxial RF

resonator [25] as a step-up transformer to match the 50 Ω generator to the trap impedance,

which is usually capacitive on the order of several pF. A schematic of such a resonator is shown

in Fig. 2.11. If no trap is attached to the resonator, the fundamental resonant RF wavelength is

λ = 4L in the ideal case, where L is the length of the resonator; thus the term “λ/4 resonator”.

Attaching the trap lowers the resonant frequency and the Q of the resonator due to trap capacitance and leakage current.
30

DC piece

crossbar

RF piece

RF piece

mask

Figure 2.10: Schematic cross-section of the planned trap structure. Not to scale. Gray shading:

Si pieces. Dots: glass pieces. Crossbars are partially cut away for clarity.
31

outer conductor (ground)

inner conductor

feedthrough

antenna

trap

E

*

B

*

Figure 2.11: Schematic of a λ/4 coaxial resonator. The antenna couples power from the RF

generator into the resonator. The high RF voltage developed at the other end of the resonator

drives the trap electrodes.

To couple power into the resonator, we used a commercial RF generator to drive a loop

antenna mounted at the opposite end of the resonator from the trap. At the antenna end of the

resonator, the center conductor was connected to the outer conductor, so at this end the magnetic field was comparatively strong and the electric field was weak, allowing coupling through

the magnetic field generated by the loop antenna. By adjusting the size and position of the

loop antenna, reflected powers as low as 0.25% of the input power were obtained on resonance.

Effectively, then, the RF generator looks into a 50 Ω load on resonance. At the trap end of

the antenna, the electric field was strong and the magnetic field was weak, providing the large

voltages we required.

The resonators for all the linear traps and all but one of the conventional Paul traps

were constructed of oxygen-free high-conductivity (OFHC) copper and were mounted inside the

vacuum system (see Section 2.4) to avoid breakdowns at the high-voltage end of the resonator.

This arrangement also minimizes the high-voltage demands on the RF vacuum feedthrough.

The feedthrough can be modeled as a capacitor in parallel with a large shunt resistance. If the
32

feedthrough were close to the high-voltage end, we would need to design it so that breakdowns

would not occur. Further, the feedthrough capacitance could significantly lower the resonant

frequency. Finally, more power would be dissipated because of the higher voltages across the

feedthrough. All these effects are negligible in the current arrangement.

The vacuum-mounted resonators used with linear traps #2 and #3 had an inner conductor diameter of 0.250” (6.35 mm), an outer conductor ID of 0.750” (19.05 mm), an outer

conductor OD of 1.000” (25.40 mm), and an inner conductor length of 11.7” (297.2 mm). The

outer conductor was slightly longer than the inner conductor to leave room for the trap. The

resonant frequency was about 250 MHz without the trap and 230 MHz with the trap. These

resonators displayed loaded Q’s of about 600 without the trap and about 500 with the trap.

The resonator used with linear trap #1 differed from the others in that an alumina tube

was inserted between the inner and outer conductors to reduce the resonant frequency. With the

alumina tube inserted, the resonant frequency was 115 MHz without the trap and 113 MHz with

the trap. The loaded Q was about 200 without the trap and 150 with the trap. We had hoped

that the reduction in resonant frequency would increase the secular frequency (see Eq. 2.11),

but because of the lower Q we were unable to maintain the same RF voltage at the trap, so that

we realized no net gain in secular frequency from this maneuver.

For the experiments with the “double” Paul trap (see Section 7.4) we mounted the resonator outside the vacuum system. A two-pin RF feedthrough connected the resonator to the

trap. This tactic enabled us to construct a number of resonators with different resonant frequencies, so that we could vary the drive frequency fairly easily. All these resonators were similar to
33

that shown in Fig. 2.11. The resonator used for the data in Section 7.4 had a resonant frequency

of 148 MHz and a Q of about 230 when attached to the feedthrough.

The RF voltages attainable at the trap were usually limited by electrical breakdown inside

the vacuum chamber. For the linear traps, electrical breakdowns at the trap occurred at several

MV/m, corresponding to several hundred volts of RF. The breakdown path appeared to be the

shortest path joining RF and DC electrodes. Breakdown voltages for a parallel-plate geometry

under comparable conditions occur at fields of about 25 MV/m [26]. We therefore do not expect

to significantly increase the field strength at the trap in the near future. However, breakdowns

also occurred in the body of the resonator, lowering the Q of the resonator and limiting the

trap voltage to yet smaller values. The location of the breakdowns observed is consistent with

the “multipactor” effect [27]. In this process, the RF voltage is such that an electron starting

from the inner conductor arrives at the outer conductor (or vice versa) in half the period of

the RF drive. The electron impact on the conductor creates secondary electrons, which are

again accelerated by the RF, and the resulting avalanche leads to electrical breakdown. This

phenomenon was only observed for the vacuum-mounted resonators and at locations near the

input end of the resonators. In the external resonator, the electrons were stopped by air and

the avalanche could not occur. To reduce the likelihood of breakdowns in the vacuum-mounted

resonators, we initially applied a few mW of RF and slowly increased the RF power to several

watts over a period of a few days. This procedure presumably reduced the secondary electron

emission coefficient as the plasma discharge cleaned the resonator surfaces.
34

2.4 Vacuum System

The vacuum system for the linear traps and all but one of the Paul traps consisted of a

quartz envelope containing the resonator and trap, an ion pump, and a titanium sublimation

pump (TSP). The ion pump had a pumping speed of 22 L/s and was run continuously while the

trap was in use. The titanium sublimation pump (TSP) was run about once a week for a few

minutes, coating the inside of the TSP cartridge with titanium for gettering. Optical-quality

quartz windows were attached to the quartz envelope (using glass frit) for optical access. The

vacuum system was assembled using standard UHV conflat fittings.

After assembly, the vacuum system was roughed out with a commercial turbopump system, then baked in a home-built oven at 350◦ for several days into a large external ion pump.

During this time the pressure was monitored with the internal ion gauge. The pressure decreased

to a steady-state value of ∼ 10−7 − 10−8 bar. The connection to the external ion pump was

then closed off, the internal ion pump was turned on, and the system was baked for another day

or two, until the pressure leveled off again. Typically the pressure decreased by another order

of magnitude during this time. The system was then brought back down to room temperature,

which decreased the pressure to 10−9 −10−10 bar. At this point the TSP was run several times,

eventually yielding pressures on the order of several ×10−11 bar. These pressures are near the

limit of sensitivity for ion gauges, so it was difficult to compare the results from the various

bakes. In practice, we used the ion lifetime to judge pressure; while this hardly provides a quantitative measure of the pressure, the ion lifetime, not the pressure itself, was the quantity of

interest. Typically we would run the TSP if the ion lifetime fell below several hours. Lifetimes

of ten hours were not uncommon.
35

For the “double” Paul trap, the resonator was mounted outside the vacuum system (see

Section 2.3). We replaced the quartz envelope with a commercial stainless steel vacuum chamber

shaped roughly like a hemisphere, which held the trap. This chamber had one large sapphire

window mounted to the flat side of the hemisphere and two small sapphire windows. These

windows exhibited much more birefringence than the quartz windows, presumably because the

absorption edge of sapphire is closer to the laser wavelength of 313 nm. Using quartz windows,

however, would have required glass-to-metal seals several inches long.

Because of the long ion lifetimes and the lack of independent measures for gas partial

pressures in the system, the effects of background gas are difficult to investigate rigorously. The

estimated rate for elastic collisions of trapped ions with background gas is about one per minute

at typical pressures [2]. Since the repetition rate of the experiment was usually at least several

hundred Hz, elastic collisions were not expected to affect our data. The only identifiable trap

“loss” mechanism was the reaction Be+ + H → BeH+ [28], as determined from measurements

of normal-mode frequencies of ion crystals before and after the reaction.

2.5 Ion Source

We produced 9Be+ ions by electron bombardment of neutral Be. Probably only the Be

atoms that were ionized inside the trapping region were actually trapped, since the large RF

fields at the boundary of the trapping region tended to accelerate ions trying to enter the trap.

However, it is difficult to verify this idea. The neutral Be source consisted of Be wire (50 μm

diameter) helically wrapped around a coil of tungsten wire (several mm long, 0.1 mm diameter)

with Be wire winding pitch equal to the Be wire thickness. Resistive heating of the tungsten

wire with current ∼ 1 A caused beryllium to evaporate from the Be wire.
36

Beryllium plating of the trap wafer can create electrical shorts between electrodes. To

prevent this, we masked off the trap wafer with a copper baffle. The baffle had a hole 2.3 mm

wide drilled in it to allow beryllium to reach the trap, and was positioned a few mm away

from the trap. The baffle was electrically isolated from the other trap components and was

electrically connected to the DC feedthrough. This allowed us to apply voltage to the baffle for

compensation of ambient static electric fields (see Section 2.1.1).

The electrons required for ionization of the neutral Be were produced by electron guns

made of pieces of tantalum foil 4 mm long, 1 mm wide, and 12 μm thick. To operate an e-gun,

we biased it to a negative voltage of 50-100 volts relative to the vacuum chamber and passed

several amps of heating current through the 4 mm length of the gun to induce thermionic emission. The biasing ensured that the emitted electrons left the vicinity; otherwise space-charge

effects reduced the emission current. We set the bias voltage to roughly maximize the ionization cross-section for Be. Typically we adjusted the heating current so as to give 20-100 μA of

emission current from the gun, of which a few nA was collected on the trap electrodes.

To load ions into the trap, we operated the Be oven and the electron gun simultaneously.

We could adjust the loading rate by varying the oven current and the electron emission current.

Usually we chose a loading rate of about one ion per minute, so as to easily control the total

number of ions in the trap (recall that our experiments involved no more than four ions.) However, we could increase the loading rate to perhaps ten ions per minute by increasing the oven

current by about 20%.
Chapter 3

Laser-Ion Interactions

We used laser light near 313 nm to manipulate the internal and motional states of the

trapped 9Be+ ions. Light nearly resonant with atomic transitions Doppler-cooled the ions and

provided initialization of the internal state into the |↓i qubit level. Nonresonant light induced

stimulated Raman transitions which coherently couple the internal states to each other and to

the motional state of the ions. These Raman transitions allowed us to cool the ion motion to

the ground state with high efficiency.

3.1 Atomic Structure of 9Be+

The 9Be+ ion has one valence electron, making it structurally similar to the neutral alkali

atoms. The energy levels of 9Be+ relevant to these experiments are the 2s 2S1/2 ground state

and the 2p 2P1/2 , 2P3/2 excited electronic states. An energy-level diagram is shown in Fig. 3.1.

The fine-structure splitting between the 2P1/2 and 2P3/2 states is 197 GHz. The nuclear spin

of 9Be+ is 3/2, inducing hyperfine splitting of 1.25 GHz in the 2S1/2 state and 237 MHz in

the 2P1/2 state [29]. The hyperfine splitting of the 2P3/2 state is less than 1 MHz [30]. The

2S1/2 ↔ 2P1/2 transition lies at 313.395 nm, while the 2S1/2 ↔ 2P3/2 transition lies at 313.266

nm. Both transitions are dipole-allowed with radiative linewidths of 19.4 MHz. Transitions

between hyperfine states in the 2S1/2 manifold, however, have radiative lifetimes on the order

of 1012 s [2], so we neglect these radiative decay processes throughout.
38

We usually used the 2s 2S1/2 |F = 2, mF = −2i and |F = 1, mF = −1i magnetic sublevels

as qubit logic states. These are denoted |↓i and |↑i respectively, in analogy to the two states of

a spin-1/2 particle. Magnetic field coils provided a quantization field of 165 μT directed along

−x/ˆ √2+(ˆy−zˆ)/2, in the coordinate system of Figure 2.1. The Zeeman splitting of the magnetic

sublevels is given by ∆E = gF mF μBB, where B is the magnetic field strength, μB = h · 14.0

kHz/μT is the Bohr magneton, and gF is the Land ́e g-factor with g1 = −1/2, g2 = +1/2.

3.2 Logic State Initialization and Readout

Quantum computation requires the preparation of the computational register in a welldefined input state at the beginning of the computation and the efficient readout of the state

of the register at the end of the computation. Thus, all the experiments described here began

by initializing the ions to the |↓i logic state and ended by detecting the number of ions in the

|↓i state. These tasks were accomplished by two laser beams, called “Blue Doppler” (BD) and

“Red Doppler” (RD), which were nearly resonant with the 2S1/2 |F = 2i ↔ 2P3/2 and 2S1/2

|F = 1i ↔ 2P1/2 atomic transitions, respectively. The Zeeman splitting of the magnetic sublevels was about 24 MHz, comparable to the 19.4 MHz excited-state linewidth. Both beams

propagated along the quantization axis B~ and were polarized σ− (see Fig. 4.7 in Sec. 4.3.2 for

a schematic rendering of the beam geometry.) At the beginning of each shot of an experiment,

we applied the BD and RD beams at the same time for about 20 μs, then left RD on for an

additional few μs. This procedure optically pumped the ions into the |↓i state with a probability

exceeding 99.9%. The extra few μs of RD ensured that the ions would not be pumped into a

dark hyperfine superposition [31] in the event of imperfect BD or RD polarization. BD was

always detuned from the exact |↓i ↔ 2P3/2 resonance by 8 MHz to ensure that the ions were

cooled to the Doppler cooling limit at the beginning of each shot. This detuning is optimal for
39

m = -2

1.25 GHz

m = -1

↓

↑ F = 1{

{ F = 2

2S1/2

2P3/2

2P1/2

197 GHz

λ  313 nm

Figure 3.1: Partial energy level diagram of the 9Be+ ion. The dipole-allowed S ↔ |Pi transitions

occur near 313 nm and decay with linewidth of 19.4 MHz. The ground-state hyperfine levels

|F = 2, mF = −2i, |F = 1, mF = −1i are used as qubit logic states and denoted |↑i and |↓i

respectively. The hyperfine and magnetic structure of the P manifold is not shown.
40

Doppler cooling of 9Be+ [32], as determined from temperature measurements on the cooled ions

(see Sec. 3.3.4.)

We detected the number of ions in |↓i by turning on BD for several hundred μs and

detecting the scattered light on either a multichannel plate imager (MCP) or a photomultiplier

tube (PMT). Assuming perfect σ− polarization, an ion in |↓i can make transitions only to the

|F = 3, mF = −3i hyperfine sublevel of the 2P3/2 state. Dipole selection rules then force it

to decay back to the |F = 2, mF = −2i hyperfine sublevel, i.e. the |↓i state. The number of

photons scattered by a |↓i ion on this cycling transition is limited only by the polarization of the

BD beam. Since the linewidth of the 2S1/2 ↔ 2P3/2 transition is 19.4 MHz and the hyperfine

splitting is 1.25 GHz, the states |↑i and |↓i are spectrally well resolved. Thus we can illuminate

the ion with BD for a long time, scattering many photons if the ion is in |↓i, but scattering

very few if the ion is in |↑i. The detection duration is limited by off-resonant repumping of

|↑i to |↓i by BD, which eventually causes an initially dark ion to scatter many photons. This

technique for high-efficiency internal state discrimination is commonly used in studies of trapped

ions [33, 34, 35].

By performing many repetitions, or shots, of an experiment and binning the results according to the number of photons detected, we can build up a histogram of the photon statistics

for that experiment. Fig. 3.2 shows such a histogram for an ion prepared in |↑i, and another histogram for an ion prepared in |↓i. The two histograms are readily distinguished. We expect the

|↓i histogram to be Poissonian with m the mean number of photons detected. Non-Poissonian

statistics for the |↓i histogram indicated a problem with the experiment, for instance BD intensity fluctuations. The observed ratio of variance to mean for |↓i histograms taken over 1000

shots varied randomly between about 0.95 and 1.05 when the experiment was running smoothly;
41

Number of Photons

0 20 40 0

0.5

1

0 20 40 0

0.1

0.2

Probability

(a) (b)

Figure 3.2: Histograms of the number of photons detected from a single ion over 1000 shots of

the experiment. (a) Ion prepared in |↑i for each shot. (b) Ion prepared in |↓i for each shot.

the nominal value is 1. We discuss the distribution of the |↑i histogram below.

In experiments with one ion, an experiment detected the ion in |↓i on some shots and

in |↑i on others, for instance if we prepared the ion in a superposition state. A histogram

exemplifying this situation is shown in Fig. 3.3. We fit such histograms to a weighted sum of

two reference histograms, one giving the nominal count distribution for an ion in |↓i, the other

for |↑i. In each shot of the experiment, the ion was projected into either the state |↓i or the state

|↑i, in accordance with the quantum measurement postulate, so fitting the histogram was similar

to counting up the number of times that the ion was projected into |↓i or |↑i over the course of

many experiments. On the other hand, if we wanted to read out the ion state on a shot-by-shot

basis, we could set a discriminator at, say, 4 photons detected (for the case of Fig. 3.2). Then

we considered the ion to be in state |↓i if more than 4 photons were detected for a particular

shot. The discriminator method was conceptually simpler but yielded lower signal-to-noise for

the measurement of |↓i and |↑i probabilities than the histogram method, because it did not take

into account the details of the count distributions in the overlap region between the |↓i and |↑i

reference histograms. However, even using the discriminator method, the small overlap of the

reference histograms permitted routine detection efficiencies of 98% in a single shot.
42

0 20 40 60 0

0.1

0.2

0.3

0.4

0.5

Number of Photons

Figure 3.3: Histogram of the number of photons detected from a single ion prepared in a

superposition of |↓i and |↑i over 1000 shots. The solid line is the best fit to the theoretical count

distribution.
43

One limit to our readout technique was imperfect polarization of BD at the ion. To

ensure good polarization, we linearly polarized the BD and RD beams and sent them through a

λ/4 plate mounted on a precision rotation stage just before they pass into the vacuum system.

To optimize the polarization, we monitored the ion fluorescence with BD continuously on and

RD off. If the polarization was imperfect, the ion occasionally decayed from the 2P3/2 state to

a state in the 2S1/2 |F = 1i manifold. The ion then stopped fluorescing until BD off-resonantly

repumped it to the 2S1/2 |F = 2i manifold. We attempted to maximize the fluorescence under

these conditions by adjusting the rotation angle and tilt of the λ/4 plate and by adjusting the

direction of the quantizing magnetic field so that it lies exactly along the propagation direction

of BD. This technique ensured that all but about a part in 105 of the BD intensity was σ−

polarized with respect to the ions. In principle, it is possible to completely eliminate this source

of readout inefficiency.

The off-resonant repumping of |↑i limited our readout efficiency from a more fundamental

standpoint. The scattering rate of BD photons from |↑i is determined solely by the linewidth of

the 2S1/2 ↔ 2P3/2 transition and the ground-state hyperfine splitting. If an ion in |↑i scatters

even one photon, the ion can decay from the 2P3/2 state into the ground |↓i state. At that

point it will begin to fluoresce brightly, so that we might mistakenly think that it was in |↓i

all along. The photon count distribution arising from this process is not trivial to calculate,

since the transition from |↑i to |↓i occurs at a random moment during the detection interval.

A detailed analysis can be found in [28]. In the absence of repumping, we should never detect

photons for an ion in |↑i, but the repumping mechanism gives a probability αp of detecting one

or more photons. If β is the scattering rate of BD photons from the |↑i state, we find [28]

αp ≈ 4β

19πζdetγ

m0 (3.1)

where ζdet is the detection efficiency and Γ = 2π×19.4 MHz is the linewidth of the excited state.
44

The factor m0 is the mean number of photons detected if the ion is in the state |↓i. For 9Be+ we

find αp ≈ 1.35×10−5/ζdet. The detection efficiency for our experiment was ζdet ≈ 8×10−3 (see

Sec. 4.5) and we typically had m0 ≈ 10, so the actual value of αp was ≈ 0.01. Of course, any

stray light entering the detector would also increase the number of detected photons, but the

limit set to spin detection efficiency by αp is fundamental.

In experiments with multiple ions, we collected all the fluorescence from the ion string at

once, making no attempt to spatially resolve the ions. Thus we detected only the number of ions

in |↓i, rather than reading out the ions individually. Typically we collected histograms of photon

count distribution as a function of an experimental parameter. In these cases we constructed

reference histograms by fitting weighted sums of theoretical photon count distributions for zero,

one,... ions in state |↓i to a histogram created by averaging all the data histograms [28]. The

free parameters in the theoretical distributions were the mean number of photons collected per

ion in |↓i and the mean number of photons due to background light. The effect of off-resonant

repumping was included for each reference histogram, using a fixed value of αp. Optical pumping

due to imperfect BD polarization was not included in the fits. The weights in the fits to the data

histograms then gave the probabilities P0, P1,... of finding zero ions in |↓i, one ion in |↓i, etc.

An example histogram for four ions, including a fit to reference histograms, is shown in Fig. 3.4.

The detection efficiency for this scheme decreased as the number of ions increases. For instance,

the probability that an ion in |↑i off-resonantly repumped to |↓i was larger for a larger number

of ions in |↑i. Pumping of an ion out of |↓i due to imperfect polarization of BD was also more

likely if we had many ions in |↓i. Histograms built up over 1000 shots yielded probabilities Pi

with typical uncertainties of ±0.01 for two ions and ±0.02 for four ions. It is also possible to

use the discriminator method for multiple ions, but the signal-to-noise is relatively low unless a

more elaborate readout scheme is used (see Section 6.1).
45

0 50 100 150 200 0

0.01

0.02

0.03

0.04

0.05

Number of Photons

Probability

P0 = 0.07

P1 = 0.24

P2 = 0.41

P3 = 0.25

P4 = 0.03

P0

P1

P2

P3

P4

Figure 3.4: Histogram for four ions prepared in a superposition state. The solid line shows the

best fit to a sum of reference histograms.
46

3.3 Coherent Coupling of Logic States

Quantum computation with the ion string required coherent manipulation of the ion

internal and motional states. Two-photon stimulated Raman transitions provided this coherent

coupling in our experiment. To drive Raman transitions between two hyperfine levels |1i, |2i

in the 2S1/2 manifold, we applied two laser beams at ∼ 313 nm with frequencies ω1, ω2 and

wavevectors *

k1,

*

k2. The electric fields of the beams are then

*

Ej = ˆ2jEj cos(*kj ·

*

x − ωj t + φj ) j = 1, 2 (3.2)

where Ej is the amplitude, ˆ2j is the polarization and φj is the phase of the jth laser field.

The Raman laser frequencies are typically detuned several tens of GHz from the nearest S ↔ P

resonance, so the amplitudes of the excited |Pi states remain small. Writing *

μ for the atomic

dipole moment operator and setting ~ = 1, we find the dipole interaction Hamiltonian

Hdip = −*

μ · (

*

E1 + *E2) (3.3)

= X

j=1,2

Ωj ei(

*

kj ·

*x+φj )

e−iωj t + h.c. (3.4)

where Ωj = −Ej 2ˆj ·

*

μ/2. We use a convention in which Ej and all matrix elements of Ωj are

real.

3.3.1 Three-Level Model of Raman Transitions

In general, each Raman beam will couple both the |1i and |2i states to several sublevels

of the |Pi manifold. For the moment, though, we assume the simplified level diagram shown in

Fig. 3.5. Here Raman beam 1 drives only the |1i ↔ P transition and Raman beam 2 drives only

the |2i ↔ P transition, with ω0 for the |1i ↔ P transition frequency and ω12 for the |1i↔|2i
47

transition frequency. Then we have Ωj = Ωj |Pihj| + h.c. with Ωj ≡ hj|Ωj |Pi, j = 1, 2. The full

Hamiltonian, including the energies of the atomic states, is then

H = −ω0|1ih1| − (ω0 − ω12)|2ih2| + X

j=1,2

Ωj

3

eiφL

j e−iωj t + c.c. ́

(|Pihj| + h.c.) (3.5)

where the operator φL

j ≡ *

kj ·

*

x + φj acts on the ion motional state only.

To get rid of the fast rotation of the interaction term in the Hamiltonian (3.5), we

transform to an interaction picture with respect to the free Hamiltonian −P

j=1,2 ωj |jihj| and

make the rotating-wave approximation [36] to remove terms varying at optical frequencies. We

define ∆1 = ω1 − ω0, ∆2 = ω2 − (ω0 − ω12) the laser detunings. The Hamiltonian then becomes

HI = X

j=1,2

∆j |jihj| + X

j=1,2

Ωj [Vj + h.c.] (3.6)

where we define the transition operators Vj = eiφL

j |Pihj|. In the experimental situation, we have

|Ωj/∆j | ¿ 1, so we can use the method of small rotations [37] to derive an effective Hamiltonian

on the |1i, |2i subspace. This method is equivalent to adiabatically eliminating |Pi [2, 28], but

reveals the physics of the Raman process more clearly. We define the unitary transformation

Usmall = exp



− X

j=1,2

Ωj

∆j

(Vj − h.c.)



 (3.7)

which constitutes a small nonlinear rotation of the Hamiltonian (3.6). Then to lowest order in

the |Ωj/∆j |, we have the effective Hamiltonian H0 = UsmallHIU†

small given by

H0 = X

j=1,2

Ã

∆j +

2Ω2

j

∆j

!

|jihj| − X

j=1,2

2Ω2

j

∆j

|PihP| + Ω1Ω2

μ 1

∆1

+

1

∆2

¶

[VR + h.c.] (3.8)

where we define the Raman transition operator
48

1

2

P

&1

&2

û1

&12

Figure 3.5: Level diagram for our simplified model of stimulated Raman transitions. Raman

beam 1 couples |1i to P, while beam 2 couples |2i to P.
49

VR ≡ −eiφL

|2ih1| (3.9)

φL ≡ φL

1 (

*

x) − φL

2 (

*

x) (3.10)

The operator φL encodes the information about the spatial interference pattern of the Raman

beams: φL = exp[i(

*

∆k·

*

x+φR)], where *∆k = *k1−

*

k2 and φR = φ1−φ2. From Eqs. 3.8 and 3.10

we see that the motional dynamics depend only on *

∆k. The apparently insignificant quantity

φR is one of the most important parameters in our experiments, as we will see in Section 6.

Since the amplitude in |Pi is always small for ∆ much greater than the linewidth, we

neglect the term proportional to |PihP| in Eq. (3.8). The Hamiltonian (3.8) then involves only the

states |1i and |2i; the state P has finally been eliminated. We define the single-photon detuning

∆ ≡ (∆1 + ∆2)/2, the detuning from two-photon resonance δ ≡ ∆1 − ∆2 = (ω1 −ω2)−ω12, and

the Raman Rabi frequency ΩR ≡ 2Ω1Ω2/∆. The term in Ω2

j/∆j describes the AC Stark shift

of the |ji level due to the presence of the Raman light. The detuning from the Stark-shifted

two-photon resonance is then δ0 ≡ δ + 2(Ω2

1 − Ω22)/∆. We write angular-momentum operators

for the effective spin-1/2 system |1i, |2i as

Sz ≡ (|2ih2|−|1ih1|)/2 S+ ≡ |2ih1| S− ≡ |1ih2| (3.11)

with the commutation relations

[S+, S−]=2Sz [Sz, S±] = ±S± (3.12)

The z-axis used for internal states is the quantization axis, NOT the trap axis ˆz of Figure 2.4!

We can recover the vector form of the spin operator *

S with the definitions S± ≡ Sx ±iSy. With

these modifications and definitions, the Hamiltonian (3.8) becomes
50

H0 = −δ0

Sz + ΩR

h

eiφR ei

*

∆k·xS+ + h.c.i

(3.13)

where we have redefined the zero of energy. Transforming Eq. (3.13) to an interaction picture

with respect to the free Hamiltonian −δ0

Sz gives the standard form of the Raman Hamiltonian

[2]

HR = ΩReiφR ei

*

∆k·xe−iδ0t

S+ + h.c. (3.14)

In the case *

∆k ≈ 0 (copropagating Raman beams), the Hamiltonian (3.14) reduces to

Hco ≈ ΩReiφR e−iδ0

t

S+ + h.c. (3.15)

which is completely independent of the ion motion. Assuming that the Raman beams are twophoton resonant with the |1i↔|2i transition (δ0 = 0), the Hamiltonian (3.15) describes Rabi

flopping between |1i and |2i according to





C2(t)

C1(t)



 =





cos Ωt −ieiφR sin Ωt

−ie−iφR sin Ωt cos Ωt









C2(0)

C1(0)





(3.16)

where the atomic wavefunction Ψ = C1(t)|1i + C2(t)|2i. We use this “copropagating carrier”

evolution mainly as a diagnostic tool in the experiment, since it enables us to prepare arbitrary

single-ion qubit states in a way that does not depend on the ion motional state.

3.3.2 Coupling to the Ion Motion

We performed almost all our quantum logic operations using Raman beams with an

angular separation of 90◦ and with *

∆k parallel to the trap axis ˆz. As we will see, this geometry

allows us to drive Raman transitions involving both the internal and motional states of the ions.

Since ω12 is small compared to the |Si↔|Pi transition frequency, we have k ≡ |*

k1|≈|*k2|. Now
51

the dynamics depends on the motion of the ions along the trap axis, so the full Hamiltonian

includes the energy of axial motion. For one ion we find

H = ωza†a + ΩR

h

eiφR ei

√2kze−iδ0

t

S+ + h.c.i(3.17)

where a is the annihilation operator for the quantized motion along ˆz. The Hamiltonian (3.17)

depends only on the motion along ˆz. This fact simplifies the dynamics enormously: for N ions

we need only consider the N axial normal modes rather than the full set of 3N motional modes

(see Sec. 2.1.3.) We can probe all 3N motional modes by applying counterpropagating Raman

beams along an axis skew to all three trap axes.

We continue to consider only one ion. Transforming to an interaction picture with respect

to the free Hamiltonian ωza†a, we find

Hz = ΩRS+eiφR exp £

iη ¡a†eiωzt + ae−iωzt

¢

− iδ0

t

¤

+ h.c.

= ΩRS+eiφR

X∞

m=0

(iη)m

m !

¡

a†eiωzt + ae−iωzt

¢m e−iδ0

t + h.c.

(3.18)

Here we have defined the Lamb-Dicke parameter η ≡ √2kz0, with z0 = (2mionωz)−1/2 the

zero-point wavepacket spread along ˆz. The condition k2hz2i ¿ 1, for which the ion wavepacket

is much smaller than the wavelength of the Raman beam interference pattern, is called the

Lamb-Dicke limit. For each power of a we retain only the lowest-order term in η, giving

Hz = ΩRS+eiφR e−iδ0

t X∞

m=0

(iη)m

m ! [(a†)

meimωzt + ame−imωzt

] + h.c. (3.19)

The Hamiltonian (3.19) exhibits a series of resonances at the two-photon detunings δ0 ≈

mωz. In the Lamb-Dicke limit, the series expansion in Eq. (3.19) converges rapidly. For a

particular value of m, the dynamics on resonance then obeys the approximate Hamiltonian
52

Hm = ΩReiφR (iη)|m|

|m|! S+(a†)

m + h.c. (3.20)

where as usual (a†)−1 = a. This Hamiltonian is similar to the copropagating carrier Hamiltonian

(3.15), except that it changes the motional state of the ion. Writing the ion internal/motional

wavefunction as Ψ = P∞

n=0 C1,n|1, ni + P∞n=0 C2,n|2, ni, we see that while the Hamiltonian

(3.15) couples |1, ni to |2, ni, the new Hamiltonian Hm couples |1, ni to |1, n+mi. The ion then

undergoes Rabi flopping with





C2,n+m(t)

C1,n(t)



 =





cos Ωn+m,nt i|m|−1eiφR sin Ωn+m,nt

i

|m|−1eiφR sin Ωn+m,nt cos Ωn+m,nt









C2,n+m(0)

C1,n(0)





(3.21)

Ωn+m,n ≡ ΩRe−η2/2 η|m|

|m|!

(n<)!

(n>)!

1/2

(3.22)

where n> is the greater of n, n + m. The cases of particular interest for our experiments are

m = 0 (the “carrier”), m = −1 (the “red sideband”), and m = +1 (the “blue sideband”.) For

instance, if the ion starts in the ground state |↓, 0i and we apply Raman light resonant with the

blue sideband transition, it will coherently evolve through a superposition state |↓, 0i + |↑, 1i to

the state |↑, 1i.

1 The intermediate superposition state exhibits entanglement between the spin

and the motion (see Section 5), enabling transfer of quantum information from the spin of one

ion, through a normal mode of motion involving the whole ion crystal, and back into the spin

of another ion [38]. Thus we can use a motional mode of the ion crystal as a “quantum data

bus,” creating an effective interaction between ion spins on demand, even though the ion spins

do not directly interact. The multi-bit logic operations engineered in this way are at the heart

of the quantum computing results presented in this thesis.

1 Here and hereafter, we suppress wavefunction normalization factors unless they are relevant to the point

under discussion.
53

3.3.3 Raman Transitions with Multiple Ions

In the case of N ions, we must consider the coupling to all N axial modes to find the

dynamics under the Raman drive. Instead of Eq. (3.17), we now have the Hamiltonian

H = X

N

k=1

ξkωza†

kak +X

N

j=1

ΩR

h

eiφj ei

√2kzj e−iδ0

t

Sj,+ + h.c.i(3.23)

where ions are indexed by j, normal modes by k. The ak are the normal mode annihilation

operators defined by Eq. (2.31) and φj is the relative phase of the Raman beams φR at ion

j. Defining ηj

k ≡ v(j)

k η/√ξk the Lamb-Dicke parameter of the jth ion in the kth mode and

assuming the Lamb-Dicke limit, we find the analog of Eq. (3.19)

Hz = ΩRe−iδ0

t

X

N

j=1

Sj,+eiφj

X

N

k=1

X∞

m=0

(iηj

k)m

m ! [(a†)

meimξkωzt + ame−imξkωzt

] + h.c. (3.24)

which again exhibits resonances, this time at δ0 ≈ mξkωz for coupling to the kth normal mode.

We define the total angular momentum of the ion crystal in the usual way as

*

J = X

N

j=1

*

S (3.25)

Then the carrier Hamiltonian (m = 0) becomes simply

H0 = ΩRe−iδ0

t

J+ + h.c. (3.26)

Here we have taken the phases φj = 0. We can take φj 6= 0 by making the replacement

Sj,+ → eiφjSj,+ in the following.

In our experiments the modes were always well-resolved from each other, so near the mth

sideband of the kth mode we find the Hamiltonian
54

Hm,k = ΩR

X

N

j=1

(iηj

k)|m|

|m|! Sj,+(a†)

m + h.c. (3.27)

similar to Eq. (3.20). For the mode used for the entangling gate (see Sec. 5), we require v(j)

k

independent of j. In this case the ηj

k are independent of j and we simply write ηk = η/√Nξk.

The Hamiltonian then takes the simple form

Hm,k = ΩR

(iηk)|m|

|m|! Jj,+(a†)

m + h.c. (3.28)

3.3.4 Cooling to the Ground State

At the end of each shot of an experiment, we apply BD for several hundred μs, cooling the

ions to the Doppler limit (about 10 MHz of energy). Usually, though, we use resolved-sideband

Raman cooling [39, 40, 41] to cool the ions still further. For clarity, we first consider Raman

cooling on a single motional mode of a single ion. Suppose the ion is initially in |↓; n = 1i. Driving on the red sideband for a time t = π/Ω1,0 produces |↑; n = 0i. Then, optically pumping to

the state |↓; n = 0i, we see that the motional mode has been cooled by one phonon[41]. Because

the ion is in the Lamb-Dicke limit, the optical pumping is nearly recoilless and is unlikely to

heat the ion [42, 43].

In practice, the ion starts in a thermal state, so the distribution of phonon number states

is broad. Assuming a small probability of occupying states with more than K phonons, we can

cool the whole distribution to the ground state by applying K cycles of Raman drive and optical

pumping. Here we specify the pth Raman pulse duration as tp = π/ΩK−(p−1),K−p, so that the

first Raman pulse transfers all the population in |Ki to |K − 1i, the second pulse transfers

|K − 1i to |K − 2i, etc. In this way the population is sequentially transferred to the ground

state. The Visual C++ program controlling the experiment (see Sec. 4.6) computed the pulse

durations using the inputs ΩR, the mode frequency, and the number of pulses desired. The
55

Raman and optical pumping pulses typically lasted a few μs each. The final temperature of the

mode was approximately equal to the mode heating rate (see Sec. 7.4) times the duration of a

cooling cycle.

We could not use BD in repumping the ion to |↓i. While photon scattering is nearly

recoilless in the Lamb-Dicke limit, the hundreds or thousands of photons scattered on the cycling transition would certainly cause the ion to equilibrate to the Doppler cooling limit [39].

Rather, we must optically pump to |↓i while scattering only a few photons. Scattering a few

RD photons ensures that the ion resides in the 2S1/2 |F = 2i manifold, but does not select the

mF = −2 magnetic state. To pump to mF = −2, we apply RD and the blue optical pumping

(BOP) beam simultaneously. BOP is polarized σ− in the atomic basis and is tuned to the

2S1/2 |F = 2i ↔ 2P1/2 |F = 2i transition. (BD is instead tuned to the 2P3/2 transition.) Dipole

selection rules then keep |↓i dark to BOP, while all other 2S1/2 |F = 2i transitions are bright

to BOP. Hence the cycling transition is never excited by BOP and the ion only scatters a few

photons before arriving at the dark |↓i state. The few photons scattered from BOP and RD

cause almost no heating of the ion during the optical pumping step.

To cool multiple motional modes of an ion crystal, we applied several Raman cooling

cycles to each mode in succession. In this case, however, it was quite difficult to find the pulse

durations corresponding to the procedure above, since we were coupling a single motional mode

to several spins simultaneously. Experimental data and a theoretical treatment for the two-ion

case can be found in [44]. In practice, we found that fudging the input parameters to the Visual

C++ program (see Sec. 4.6) and using the resulting one-ion pulse durations produced cooling

nearly as good as that given by the orthodox procedure. We could cool the stretch modes to the

ground state almost perfectly because of their low heating rate (see Sec. 7.4), but usually about
56

one phonon remained in the COM mode. The precise cooling limit depended on the number of

ions and the trap frequency.

As an alternative to the “pulsed” cooling scheme described above, we tried a “continuous” Raman cooling scheme in which we applied the red sideband drive and the optical pumping

light simultaneously. If the relative intensities of Raman and optical pumping light were correctly adjusted, the average phonon number in the mode being cooled decreased exponentially

with time [39, 45]. In this scheme, one does not need to compute the lengths of cooling pulses,

avoiding considerable headache. The final temperature from continuous cooling was about the

same as that for the pulsed cooling scheme with fudged pulse durations, but continuous cooling

was about a factor of two slower.

3.3.5 Raman Transitions in 9Be+

The three-level model of Raman transitions presented in Sec. 3.3.1 describes the essential features of coherent coupling through two-photon processes. However, real atoms possess

a great deal more than three levels, and the effect of these other levels plays a crucial role in

the engineering of robust quantum logic gates. In the experiments described here, the wavelength of the Raman light was usually set to 626.341 nm, for reasons described in Sections 7.2.2

and 7.3. Since the wavelength of the 2S1/2 ↔ 2P1/2 transition is 626.395 nm and that of the

2S1/2 ↔ 2P3/2 transition is 626.266 nm, both the 2P1/2 and 2P3/2 states participate in the

Raman process. At the same time, the polarizations of the Raman beams affect the two-photon

processes through the hyperfine and magnetic structure of the 2P1/2 and 2P3/2 levels.
57

To understand Raman transitions in the context of the real 9Be+ ion, we return to

the expression (3.3) for the dipole interaction Hamiltonian. For clarity, we consider only

one ion. We index the hyperfine levels of the 2P1/2 manifold by the label F1/2 and write

|F1/2i ≡ 2P1/2 |F,mF i. Similarly, F3/2 indexes the 2P3/2 manifold and FP runs over all indexes

{F1/2, F3/2}. We decompose the laser polarization in the atomic basis ˆ2a ∈ {σ+, σ−, π} defined

by the quantization field and write

Ωj = −Ej

2 h

2S1/2 , F = 2, mF = −2|σ− ·

*

μ|

2P3/2 , F = 3, mF = −3i (3.29)

S(j, FP; ˆ2a) = hj|2ˆa ·

*

μ|FPi

Ωj

(3.30)

so that a laser of amplitude Ej and polarization ˆ2a induces a Rabi frequency of S(j, FP; ˆ2a)Ωj

on the |ji↔FP transition.2 Since only one laser polarization (in the atomic basis) can drive a

given |ji↔FP transition, there is only one choice of ˆ2a for which S(j, FP; ˆ2a) 6= 0; however, it is

notationally convenient to retain the index ˆ2a. The S-factors (3.30) are tabulated in Appendix B

of [28]. We choose a convention in which Ωj and S are real, and as usual ~ = 1. Fixing the zero

of energy at the energy of the 2P1/2 level, writing the fine-structure splitting ωFS = 2π × 197

GHz, and neglecting the hyperfine splittings of the P levels, we find the full Hamiltonian in

Schr ̈odinger picture

H = −ω0|1ih1| − (ω0 − ω↓↑)|2ih2| + ωFS X

F3/2

|F3/2ihF3/2|

+ X

j=1,2

X

FP,2ˆa

Ωj S(j, FP; ˆ2a)

h

(ˆ2

∗

a · 2ˆj )eiφL

j e−iω1t + c.c. ́

(|FPih1| + h.c.] (3.31)

for |1i, |2i in the 2S1/2 manifold. In the experiment, we always ensure that the transition

frequencies between the various 2S1/2 hyperfine states are well-resolved, so that we need only

consider two 2S1/2 sublevels in Eq. (3.31).

2 Recall that the 2S1/2 ↔ 2P1/2 and 2S1/2 ↔ 2P3/2 linewidths are equal.
58

Now going to an interaction picture with respect to the laser frequencies and making

the rotating-wave approximation to remove terms varying at optical frequencies, we find the

generalization of Eq. (3.6)

HI = X

j=1,2

∆j |jihj| + ωFS X

F3/2

|F3/2ihF3/2| + X

j=1,2

X

FP,2ˆa

Ωj S(j, FP; ˆ2a) ((ˆ2

∗

a · 2ˆj )Vj,FP + h.c.)

(3.32)

where Vj,FP = eiφL

j |FPihj|. We again apply the method of small rotations [37], defining the

unitary transformation

Usmall = exp





− X

j=1,2

X

FP,2ˆa

ΩjS(j, FP; ˆ2a)

∆j,k

[(ˆ2

∗

a · 2ˆj )Vj,FP − h.c.]







(3.33)

∆j,k ≡







∆j k = 1/2

∆j − ωFS k = 3/2

(3.34)

Transforming HI → H0 = UsmallHIU†

small and again neglecting terms in |FPihFP|, we find the

effective Hamiltonian

H0 = X

j=1,2



∆j + 2Ω2

j

X

k={1/2,3/2}

R(k)

Stark,j

∆k

|jihj| + ΩRRRaman

ωFS

ωFS − ∆ [VR + h.c.]



 (3.35)

R(k)

Stark,j ≡ X

Fk,2ˆa

(S(j, Fk; ˆ2a))2|2ˆ

∗

a · 2ˆj |

2 (3.36)

RRaman ≡ X

F1/2,2ˆa,2ˆ0

a

S(1, F1/2; ˆ2a)S(2, F1/2; ˆ2

0

a)(ˆ2

∗

a · 2ˆ1)(ˆ2

∗

a0 · 2ˆ2) (3.37)

1

∆k

≡ 1

∆1,k

≈ 1

∆2,k

(3.38)

One can readily verify that RRaman is invariant under the replacement F1/2 → F3/2, though

the individual factors S(j, FP; ˆ2a) are not. The detuning from the Stark-shifted two-photon

resonance becomes
59

δ12 = δ + X

k={1/2,3/2}

Ã2Ω2

1R(k)

Stark,1

∆k

− 2Ω2

2R(k)

Stark,2

∆k

!

(3.39)

and the Rabi frequency of |1i↔|2i transitions becomes

Ω12 = ΩRRRaman

ωFS

ωFS − ∆ (3.40)

= Ω1Ω2RRaman

ωFS

∆(ωFS − ∆) (3.41)

After some manipulation, the Hamiltonian assumes the form

H12 = Ω12eiφR ei

*

∆k·xe−iδ12t

S+ + h.c. (3.42)

identical to the form of the effective Hamiltonian (3.14) derived for the simpler three-level system in Sec. 3.3.1. The results following from Eq. (3.14) in Secs. 3.3.1 and 3.3.2 then hold if we

make the replacements ΩR → Ω12, δ0 → δ12.
Chapter 4

Laser Apparatus and Experiment Control

4.1 Laser Sources at 626 nm

We generated the 313 nm light used in these experiments by resonantly doubling 626 nm

light from stabilized dye lasers. We needed to access the 2S1/2 ↔ 2P1/2 transition at 313.197

nm and the 2S1/2 ↔ 2P3/2 transition at 313.133 nm, and we also needed off-resonant light

at 313.170 nm to generate the Raman beams. These wavelengths are separated by hundreds

of GHz, so accessing multiple wavelengths by modulation from a single laser was not possible.

Therefore we used three dye laser sources operating at 626.395 nm, 626.266 nm, and 626.341 nm

and resonantly doubled these laser beams separately to produce the desired UV wavelengths.

The detection laser (626.266 nm) and the Raman laser (626.341 nm) were home-built to

a design by Jim Bergquist, our resident laser expert (Fig. 4.1). The laser at 626.395 nm was a

Coherent model 699. All three lasers used Kiton Red dye, manufactured by Exciton. Typically

1.5 g of dye was dissolved in ∼ 1 L of ethylene glycol. The dye solution was drawn from a

reservoir and forced through a laminar flow nozzle under an input pressure ∼ 2.75 bar to form

a jet 100 μm thick. This jet provided the laser gain when optically excited by a pump laser.

The dye jet traveled about 5 mm from the nozzle before intersecting the pump beam, and was

oriented at Brewster’s angle for the 626 nm light. To obtain good laser power and linewidth, it
61

Figure 4.1: Jim says, “I love the smell of laser dye in the morning...”

was necessary that the dye jet travel 20 cm past the gain region before striking the catch tube

that returned the dye to the reservoir; otherwise acoustic noise generated at the striking point

traveled back to the gain region, degrading the laser performance. The pump for the Raman

laser was a Coherent Verdi V-5 doubled Nd:YAG laser operating at 532 nm with 5.5 W output

power, yielding ∼ 1 W of 626 nm light from the Raman laser. The pump for the detection

laser and the 699 was a Spectra-Physics Model 2030 argon-ion laser operating at 514 nm with

typically 14–16 W output power. The argon laser light was split two ways to provide 7–8 W to

each dye laser, yielding 600 mW at 626 nm from the detection laser and 300 mW from the 699.

The dye laser powers quoted are for fresh dye. Over time, the dye ages and its gain decreases.

Typically we replaced the dye solution when the laser power dropped to about 2/3 its maximum

value. All three lasers were S-polarized (electric field perpendicular to the optical table.)

The optical design for the homemade lasers is shown in Fig. 4.2. The optical design for

the 699 can be found in [46]. Each laser also had several intracavity elements: an optical diode

to ensure unidirectional operation, a three-plate birefringent filter, a thin etalon, and a thick
62

etalon. The frequency-dependent transmission of the last three elements modulated the cavity

loss at free spectral ranges (FSRs) of 1.9 THz, 200 GHz, and 10 GHz respectively. When the

transmission maxima of the three elements overlapped at a particular frequency, gain competition ensured single-mode operation of the laser [47]. The birefringent filter was manually tuned

with a micrometer. The thin etalon consisted of an optically flat glass plate ∼ 0.5 mm thick,

mounted to a galvo motor. We tuned the thin etalon by adjusting the current to the galvo. The

thick etalon was similar to that found in the 699 [46] and was tuned by adjusting the voltage to

a piezoelectric transducer (PZT) mounted in the thick etalon assembly. The 699 also included

an intracavity Brewster plate to extend the range over which the laser could be tuned without

mode-hops [46]. The thick etalon of each laser was locked to the point of maximum laser power

by modulating the thick etalon voltage at 30 kHz and feeding back the demodulated laser intensity signal to the average thick etalon voltage. The feedback was provided by an integrating

gain stage with bandwidth 1 kHz. This lock maximized the thick etalon transmission at the

frequency of the lasing mode. The three tuning elements allowed us to tune each laser to within

a few hundred MHz of any desired frequency. We could scan each laser about 1 GHz without

mode-hops.

The homemade lasers were frequency-stabilized to a linewidth of . 500 kHz using feedback to the 30 cm mirror, which was mounted on a PZT. The bandwidth of the servo was limited

by the mechanical resonance frequency of the PZT, which was 10 kHz for the detection laser

and 20 kHz for the Raman laser. The error signal was derived by the H ̈ansch-Couillaud method

[48] using a 20 cm confocal reference cavity in each case. The reference cavities were made by

contacting mirrors to the ends of an Invar tube and were enclosed in insulating boxes to minimize thermal drift of the cavity length. One mirror of each cavity was mounted on a PZT for

external adjustment of the cavity resonance. The servos used for locking had proportional and

integral gain, as well as a second stage of integration. Usually we used only the two integration
63

laser 

output

R = , 

T = 2%

pump 

laser

dye jet

R = 10 cm

R = 10 cm

R = 10 cm

R =  (30) cm

90 cm

40 cm

6 (5) cm 6 (7) cm

5 cm

PZT

Figure 4.2: Optical design of a homemade laser. Intracavity optical elements are not shown.

Values refer to the Raman (detection) laser. R = radius of curvature, T = transmission, PZT

= piezoelectric transducer. Drawing is not to scale.
64

stages. The corner frequencies and overall gain were routinely adjusted to give the best noise

reduction. The 699 was frequency-stabilized to a linewidth of about 10 MHz by sidelock to a

reference cavity [46].

Our experiments were insensitive to slow drifts of the Raman laser frequency, as these

drifts amounted to less than hundreds of MHz over a day, much less than the Raman laser

detuning of tens of GHz. Thus it was not necessary to stabilize the Raman laser in absolute

frequency. However, similar drifts of the absolute frequencies of the 699 and detection lasers

would seriously affect our experiments, since we needed the doubled light from these lasers to

be nearly resonant with atomic transitions of linewidth 20 MHz. We therefore used saturatedabsorption iodine spectroscopy [49] to stabilize the 699 and detection lasers relative to iodine

lines. We fed back the resulting error signals to the reference cavity PZTs using integrating

servos with bandwidths of 10–100 Hz.

The beams used for iodine spectroscopy were derived from the main laser beams using

acousto-optic modulators (AOMs), allowing us to offset the laser frequency from the frequency

of the iodine line by up to ∼ 1 GHz. For each laser, we were able to identify an iodine line

whose offset from the desired laser wavelength was within this range. We used line #954 of

the “Iodine Atlas” [50] to lock the 699 and line #961 to lock the detection laser. Our lock-in

detection technique produced a series of dispersive error signals, one for each component of the

iodine line. We locked each laser to the bluemost hyperfine component of its iodine line. The

spectrometer setup and detection electronics are described in detail in [28], although the offset

frequencies given there are NOT those relevant to this thesis. The 699 was offset by a fixed

frequency of -21.5 MHz from the iodine lock point (after implementation of the 617 MHz EOM -

see Secs. 4.2 and 4.3.1), so that the laser wavelength was twice that of the 2S1/2 ↔ 2P1/2 |F = 2i
65

transition. The detection laser was nominally offset by +804.9 MHz from its iodine lock point,

but a double-pass AOM in the iodine spectrometer allowed us to vary the offset by about 10

MHz in either direction. The nominal offset was such that the BD frequency at the trap was 8

MHz red of the cycling transition frequency with the typical offset to provide Doppler cooling

of the ions (see Sec.3.2.) Because of the large slope of the cycling transition scattering rate at

this detuning, any increase in linewidth or drift of the detection laser led to broadening of the

Poissonian distribution of detected photons for ions in |↓i, decreasing our detection efficiency.

Our tolerances for the 699 linewidth and drift were much looser, since efficient optical pumping

only required the ions to scatter a few photons over several μs.

We used a Burleigh model #SA-PLUS-200-B1 spectrum analyzer and a traveling-cart

wavemeter [51, 52] for rough diagnostics of the laser mode structure and wavelength. An arrangement of beamsplitters and mirrors allowed us to monitor any one of the three lasers at any

time. The wavemeter read out the laser wavelength to six significant digits with an update time

< 1 s, allowing us to tune the lasers to within ∼ 1 GHz of the desired frequencies with ease. The

quoted laser wavelengths, e.g., 626.266 nm for the detection laser, are exactly the wavemeter

readings. These wavelengths agree with previously measured values [53]. The wavemeter could

also read out a seventh digit with update time ∼ 10 s, but this feature was rarely used. The

spectrum analyzer had an FSR of 2 GHz and was especially useful for observing mode hops

while tuning the thick etalon. Typically such mode hops changed the frequency of the laser by

the laser cavity FSR of 150 MHz, which was too small to observe on the wavemeter but could be

readily resolved on the spectrum analyzer. Occasionally we wanted to measure the frequency of

the Raman laser to a precision better than that provided by the wavemeter. By coupling both

the Raman laser and the detection laser into the spectrum analyzer simultaneously, we could

obtain the difference of laser frequencies to a precision of ∼ 10 MHz modulo 2 GHz. Combining

this information with the wavemeter reading gave the absolute frequency of the Raman laser to
66

∼ 10 MHz.

4.2 Resonant Doubling to 313 nm

The apparatus described in the last section all sat on a single optical table (the “red

table”). The rest of the apparatus, including the ion trap, sat on a separate table (the “UV

table”), and the 626 nm light from the red table was carried to the UV table by single-mode

optical fibers. We applied a controlled torque to each fiber to cancel the natural fiber birefringence with the induced stress birefringence. This scheme largely decoupled optical adjustments

on the red table from those on the UV table. The light from the 699 passed through a resonant

electro-optic modulator (EOM) operating at 617 MHz before coupling into the fiber. When

this EOM was switched on, the first blue sideband of the 699 light provided BOP light (see

Sec. 3.3.4.) For the experiment described in Section 5, we had not yet implemented the EOM

and generated the BOP beam using an AOM (see Sec. 4.3.1.)

After leaving the fibers, the detection and 699 beams were coupled into a pair of resonant

doubling cavities containing a β-barium borate (BBO) crystal, which generated the UV under

type II critical phase-matching [54]. The fiber birefringence rotated the beam polarizations to

give nearly P-polarized ( *

E parallel to the optical table) inputs to the cavities. The cavities

consisted of three mirrors shared by both beams and two mirrors mounted on PZTs, one mirror

for each beam. This arrangement defined two bowtie cavities with the two resonant beams

circulating through a single BBO crystal. The optical design was that described in [28], except

that the PZT-mounted mirror of the 699 doubling cavity was moved slightly to make the doubling

cavity FSR equal to 617 MHz. This allowed resonant doubling of both the carrier and EOM

sidebands of the 699 light in the same cavity. Each cavity was actively stabilized so that the
67

cavity resonance frequency matched the frequency of the incoming red light. The error signal

was derived by polarization spectroscopy of the cavity itself. The doubling crystal was cut so

that the phase-matched red beam entered the crystal at Brewster’s angle for P-polarization of

the red beam. However, S-polarized red light was partially reflected from the crystal surface,

so that the cavity loss was different for the two polarizations. The resulting error signal for the

fundamental cavity mode was exactly that obtained by the H ̈ansch-Couillaud method [48], and

was fed back to the mirror PZT with a servo system similar to that used for locking the lasers to

the reference cavities (Sec. 4.1). The overall loop bandwidth was again set by the PZT resonant

frequency of ∼ 40 kHz.

The resonant buildup of the red beam greatly enhanced the efficiency of UV generation.

The UV output power PUV for single-pass doubling of a red beam with power PSP is

PUV = ηdblP2

SP (4.1)

where ηdbl ∼ 3 − 5 × 10−4W−1 for BBO. The resonant cavity enhances the power PSP passing

through the crystal by about an order of magnitude over the input power Pred, so that the UV

power is much larger than that obtained without the cavity. We express the overall efficiency

of the doubling process as ηdbl = PUV/P2

red. For the detection and 699 beams, we typically

achieved doubling efficiencies of 0.05 − 0.10 W−1. We then obtained typically 1–2 mW of power

in the BD beam (the doubled detection beam) and several hundred μW in the RD beam (the

doubled 699 beam). The UV light exiting the doubler was S-polarized (electric field parallel to

the optical table) due to the type II phase-matching.

The high UV powers (several mW) provided by the resonant doubling were essential for

our experiments. However, residual noise of the cavity resonance relative to the laser frequency

caused fluctuations of the UV power. The residual frequency noise mostly arose from spectral

components of the laser frequency noise that lay above the lock bandwidth. If this residual noise
68

gave an effective linewidth that was a significant fraction of the doubling cavity bandwidth (∼ 2.5

MHz FWHM for the detection and 699 cavities), the frequency noise of the red beam caused

intensity noise of the doubled beam through Eq. (4.1). The detection laser linewidth was much

smaller than the doubling cavity bandwidth, so this mechanism created little intensity noise

for BD. While the 699 linewidth was sometimes larger than the doubling cavity bandwidth, we

allowed sufficient optical pumping times to make up for any intensity fluctuations in RD.

The Raman light was coupled into an LAS (now Spectra-Physics) Wavetrain SC resonant

cavity doubler containing a lithium borate (LBO) crystal, which was type II critically phasematched for UV generation. The input to the cavity was again P-polarized. The cavity for

this doubler consisted of two mirrors and a prism mounted on a PZT, as shown in Fig. 4.3.

The cavity resonance was servoed to the red beam frequency by the Pound-Drever-Hall method

[55] with feedback of the error signal to the PZT through a servo with proportional, integral,

and derivative (PID) gain. The servo bandwidth was ∼ 6 kHz and was presumably set by the

PZT resonance. We observed doubling efficiencies of ηdbl ∼ 0.15/W for this cavity. Because

the length of the Wavetrain cavity was about a factor of 4 smaller than the lengths of the other

doubling cavities, we believe that the bandwidth was significantly higher than that of the other

cavities. The Raman laser linewidth was less than 1 MHz, so the doubling did not introduce

intensity noise. The UV light from the cavity was S-polarized.
69

PZT

LBO

3.6 cm 3.6 cm

3.6 cm 1.9 cm 0.9 cm

red 

input

red + UV 

output

Figure 4.3: Optical design of the Wavetrain SC doubling cavity. PZT = piezoelectric transducer,

LBO = lithium borate crystal.

4.3 UV Beam Control

4.3.1 BD and RD Beams

On leaving the doubling cavity, the BD beam passed through a noiseater setup using a 80

MHz AOM (Fig. 4.4) to stabilize its intensity. The signal from the photodetector was summed

with an externally controlled setpoint and filtered as shown in the circuit diagram of Fig. 4.5.

The diode clamps on the last amplifier stage help the circuit recover from dye laser “dropouts”,

momentary events in which the red laser power drops to zero. The filtered signal was input to

a mixer in the RF drive setup for the AOM, so that the signal level controlled the RF power to

the AOM. If the laser power exceeded the setpoint, the AOM turned on, deflecting the excess

power into the AOM sideband. An aperture blocked the deflected beam but let the undeflected

beam pass through. Though the noiseater can remove power from the beam, it cannot increase

the beam power. If noise causes the beam power to drop below the setpoint, the servo is ineffective in removing the noise. Hence the setpoint power was typically half the power entering

the noiseater. This setting was found effective in servoing out intensity fluctuations.

The overall noiseater bandwidth is limited by the transit time Ttrans of the acoustic wave
70

stabilized 

beam noise 

beam

input 

beam

detector

servo

RF 

synth.

RF amp

AOM

LO

IF

RF

Figure 4.4: RF/optical schematic for “noiseater” intensity stabilization. PD = photodetector.
71

in the AOM between the AOM’s RF/acoustic transducer and the position of the beam. The

AOMs used in all the UV beamlines were made of quartz, which has a sound velocity of about

6000 m/s. For a beam positioned 6 mm from the transducer, the transit time is thus 1 μs. The

transducer induces stress birefringence in the quartz, so the beam polarization changes unpredictably if the beam is moved much closer than 6 mm to the transducer. In practice, the servo

begins to oscillate slightly at the resonant frequency 1/Ttrans for servo unity-gain frequencies

much higher than 1/(2Ttrans). The resulting noiseater bandwidth is then typically ∼ 50 kHz.

After the noiseater, BD passed through a switch arrangement consisting of a –80 MHz

AOM followed by an aperture. The AOM was arranged so that its first red sideband passed

through the aperture, while the unshifted beam was blocked. Hence BD passed through to the

trap only when the RF driving the switch AOM was turned on. Figure 4.6 shows the entire BD

beamline from the doubling cavity to the trap. We specify the selection of the redshifted beam

by designating the AOM as having –80 MHz RF input; for a blueshifted beam we would write

+80 MHz. A variable RF attenuator controlled the RF power to the switch AOM, allowing us

to vary the BD power easily. With no attenuation, the switch efficiency was about 90%. We

also used an optical neutral-density filter (ND 0.3) mounted on an electrically controlled flipper

mirror for coarse control of the BD power.

In the version of the RD beamline after implementation of the 617 MHz EOM in the 699

beamline, RD passed through an 80 MHz AOM twice, once with a –80 MHz shift and once with

a +80 MHz shift in the UV, to give a net zero shift. This AOM was followed by an aperture

arranged so that RD arrived at the trap only if the RF driving the AOM was turned on. A

variable RF attenuator in the RF drive for this AOM allowed us to control the RD and BOP

power. With no attenuation, the switch efficiency was about 80%. After the switch, RD was
72

Figure 4.5: Circuit diagram for the servo used in the noiseater.
73

switch 

(-80)

to 

trap

noise 

eater

off

on

from 

doubler

Figure 4.6: Schematic of the BD beamline from the doubling cavity to the trap. Frequency

shifts of deflected beams are indicated in parentheses, in units of MHz.

combined with BD on a beamsplitter and the BD, RD beams copropagated through a polarizer and a finely adjustable quarter-wave plate to ensure σ+ polarization in the atomic basis.

The beams then passed through a final focusing lens of focal length 10 cm and into the trap.

Typically the diameter of each beam was about 100 μm at the ion string. Here we roughly

quantify the “beam diameter” as twice the distance between the point of maximum intensity

in the transverse profile and the point at which the intensity drops by a factor of 1/e. The

focusing lens was mounted on an XYZ micrometer stage which allowed repeatable positioning

of the beam at the ion to within a few μm.

In the version of the RD beamline used for the experiment of Section 5, the RD beam

was split with power ratio of roughly 1/2. One of these beams propagated through a –80 MHz

AOM switch like that used for BD and performed the usual function of the RD beam. Again

a variable RF attenuator allowed control of the RD power. In this version, the 699 laser was

offset by +18.5 MHz, rather than –21.5 MHz, from its iodine spectrometer (see Sec. 4.1.) The

other beam passed through a +236 MHz AOM four times and then passed through a +236 MHz

AOM switch. This beam performed the BOP function. The efficiency of BOP generation was

only about 10% after the beamsplitter. The BD, RD, and BOP beams were combined with

beamsplitters as described above.
74

4.3.2 Raman Beams

The geometries and polarizations used for Raman transitions in our experiments are

shown in Fig. 4.7. Beams are labeled by frequency as B (the bluer beam) or R (the redder

beam) and by the port at which they enter as 1, 2, or 3. Thus the bluer beam entering port 1 is

denoted B1. Raman beams B1 and R1 entering port 1 are used together to drive copropagating

carrier transitions between various hyperfine states according to Hamiltonian (Eq. 3.15). B1

and R2 together drive logic transitions between |↓i and |↑i according to Hamiltonian (??). B2

and R1 together drive logic transitions between the |F = 2, mF = 0i and |F = 1, mF = 0i

states. Used by itself, beam B2 is also a convenient way to Stark shift the hyperfine transitions.

Counterpropagating beams B3, R1 drive transitions involving all 3N motional modes and are

used to investigate transverse vibrations of the ion crystal.

We generated the various Raman beams (B1, R1, etc.) from the doubled Raman light

using AOMs. The doubled Raman light was first split into BR (“blue Raman”) and RR (“red

Raman”) beams by a fixed-frequency AOM at 221 MHz. The RR beam was the undiffracted

beam in this AOM, while the BR beam was the first blue sideband of the AOM. Figure 4.8 shows

a schematic of the BR and RR beamlines from the splitter AOM to the trap. BR continued

through another fixed-frequency “offset” AOM at +221 MHz driven by the same synthesizer.

This AOM also served as a noiseater similar to that described for the BD beam, except that the

polarity of the servo was reversed so that the intensity-stabilized beam was the shifted beam

rather than the unshifted beam. RR was sent through a double-pass AOM with nominal frequency –200 MHz (–80 MHz for the experiment of Sec. 5.) We could vary the RF drive frequency

of this AOM by about 15 MHz in either direction, enabling us to control the beat frequency

between BR and RR beams over times of several μs without significantly steering the RR beam.
75

B1 () R1 

(1+/1- )

B2 

(1+/1- R2 )

(1+/1- )

port 2 port 1

port 3

B3 ()

z

ˆ

B

* BD, RD, 

BOP (1-

)

Figure 4.7: Raman beam geometry for our experiment. Beams are labeled by frequency as B

(the bluer beam) or R (the redder beam) and by the port at which they enter as 1, 2, or 3. Thus

the bluer beam entering port 1 is denoted B1. Polarizations are specified in the atomic basis

{σ+, σ−, π} selected by the quantization field. The trap defines the spatial coordinate system,

with the trap axis along ˆz. Detection and optical pumping beams BD and RD are shown for

reference.
76

B1 

switch

(+221)

noise 

eater 

(+221) off

B1

BR

on

B2 

switch

(+221) off

on

B2

RR

double 

pass 

(-200)

entangling 

EOM

switches 

(each 

-221)

R2

shelve 

EOM

R1

Figure 4.8: Schematic of the BR and RR beamlines from the splitter AOM to the trap. Frequency

shifts of deflected beams are indicated in parentheses, in units of MHz. The center frequency is

indicated for the double-pass AOM. The switch arrangement for RR is the same as for BR, but

is compressed for clarity.

After the double-pass, RR was intensity-stabilized by a noiseater like that used for BD. RR then

passed through a resonant EOM with frequency 17.67 MHz. Turning on this EOM provided the

additional Raman frequency for the two-ion entangling gate (see Sec. 5.2.)

The BR beam was split into separate beams B1, B2 using two +221 AOM switches. The

aperture after the first AOM was replaced with a mirror that separated the shifted and unshifted

beams. The shifted beam was sent to the trap. The unshifted beam passed through the second

switch, which was set up in the usual way. Thus the beam passed by the second switch and

the shifted beam from the first switch were both 221 MHz blue of the original BR beam. We

generated R1 and R2 from RR in the same way, except that the AOM switches gave a shift of

–221 MHz. After the switch arrangement, R1 passed through a resonant EOM with frequency

32.67 MHz. Turning on this EOM enabled us to access the “shelving” Raman transitions (see

Sec. 6.1.)
77

R1 and B1 were combined on a polarizing beamsplitter (PBS), B1 being first sent through

a half-wave plate to rotate it to P polarization. The B1 polarization was then parallel to the

quantization field, giving it π polarization in the atomic basis. The R1 polarization was perpendicular to the quantization field, but since the R1 propagation vector was also perpendicular to

the quantization field, the R1 power was constrained to be divided equally between the atomic

σ+ and σ− polarizations. The copropagating B1, R1 beams passed through a 10 cm focusing

lens into the trap. The lens was mounted on an XYZ micrometer stage like that used for the

BD focusing lens. The beam size at the ion was usually about 20 × 40 μm, with the long dimension parallel to the optical table. Residual astigmatism from the doubling cavity caused the

ellipticity of the beam spot.

For the experiments in quantum logic on the 2S1/2 |F = 2, mF = 0i ↔ 2S1/2 |F = 1, mF =

0i transition (Secs. 7.1 and 7.3), R2 and B2 were combined on a PBS as well, with B2 rotated to

P polarization. Since R2 and B2 both propagated parallel to the quantization field, the R2 (B2)

polarization was ˆ2R(B) = σ+ + exp[iφR(B)] σ− in the atomic basis, with φR = φB + π to preserve

the orthogonality of the R2, B2 polarizations. After the PBS, the copropagating R2, B2 beams

were sent through a finely adjustable quarter-wave plate. This waveplate enabled us to control

the relative amplitudes of σ+ and σ− light in the R2 beam, which proved essential in implementing our entangling gate on the |↓i ↔ |↑i transition (see Sec. 5.2.) For the decoherence-free

subspace experiment, the PBS was removed and B2 was sent in nearly parallel to R2 using a

pickoff mirror. In this case, we controlled the R2 and B2 polarizations independently by inserting a separate quarter-wave plate in each beam (see Sec. 8.2.) B2 and R2 again passed through

a 10 cm focusing lens on an XYZ stage before entering the trap.
78

4.4 RF Generation and Control

All the RF used to drive the 80 MHz UV AOMs originated from a single HP 8640 RF

generator set to exactly 80 MHz. The RF was split five ways and passed through isolation

amplifiers to prevent crosstalk between channels. The channel driving the BD noiseater passed

through its servo-controlled mixer, while the channels driving the switch AOMs were input to the

RF switchbox. This switchbox was a collection of Mini-Circuits ZASW-series TTL-controlled

RF switches that translated the TTL logic generated by the experiment control system into

switching of the RF power and hence to switching of the UV beams. The RF switches had very

high isolation (40–60 dB) to minimize optical power leakage through the AOMs when the RF

was nominally off. Each channel was then amplified to ∼ 800 mW power and sent to its AOM.

The 221 MHz signal driving the Raman splitter and BR offset AOMs was generated by

an HP 8662 RF synthesizer. The signal was split into separate channels for each AOM and sent

through isolation amplifiers. Each channel was then amplified to 1–2 W and sent to the AOMs.

For the experiment of Section 5, we drove the 236 MHz AOMs used in the BOP beamline by

amplifying the signal from an HP 8640 RF generator.

Our experiments required that we switch the frequency of the RF driving the RR doublepass AOM in a matter of microseconds. A single RF synthesizer with the desired frequency

stability could not switch frequencies sufficiently quickly, so we used a TTL-controlled RF multiplexer constructed using Mini-Circuits ZYSW-series RF switches to generate the RF for the

double-pass AOM. The inputs to the multiplexer were driven by the ∼ 200 MHz mixing products of several HP 3335 RF synthesizers operating near 50 MHz with a single Fluke 6160 RF

synthesizer operating at 150 MHz. We generated the RF in this way because suitable 200 MHz
79

RF synthesizers would have been prohibitively expensive. After the multiplexer, the RF was

amplified to 1–2 W and sent to the double-pass AOM. All the synthesizers driving AOMs in

the Raman beamlines were phase-locked together, ensuring temporal coherence of the Raman

beams.

We usually determined the frequencies of the carrier and sideband transitions by sweeping

out the ion fluorescence spectrum as a function of two-photon detuning. In this case we used a

single GPIB-controlled HP 8660 RF synthesizer to drive the RR double-pass AOM. The GPIB

interface allowed us to change the synthesizer frequency every few ms. For this application we

collected data over about 100 ms at each frequency, so the dead-time involved in changing the

synthesizer frequency did not affect the experiment duty cycle significantly.

For the experiment of Section 5, we drove the 80 MHz RR double-pass AOM using the HP

3335 RF synthesizers directly as inputs to the RF multiplexer. The output of the multiplexer

was amplified to ∼ 800 mW and sent to the AOM.

The RF signals driving the two resonant EOMs were generated by HP 3335 synthesizers,

amplified to ∼ 100 mW, and sent to the EOM resonant tank circuits. The inductor for each

tank circuit was a coil of 14 gauge copper wire of several μH inductance enclosed in a thick metal

box. The inductor was attached across a pair of banana plugs, the connector type supplied for

this EOM (Linos Photonics PM 25). The EOM typically provided a capacitance of ∼ 30 pF for

the tank circuit; we could adjust the tank resonance by attaching additional pF-scale capacitors

across the banana plugs. The input RF power drove a loop antenna of 14 gauge copper wire,

also enclosed in the box, which magnetically coupled to the tank circuit. We typically attained
80

a Q factor of 50 when the tank circuit was attached to the EOM. Higher Q factors were neither

necessary nor desirable, as they would have limited the EOM resonant bandwidth. The resonant

EOM assemblies were further enclosed in aluminum boxes to minimize RF radiation from the

(snort) banana plugs. How Linos gets away with this banana plug business, I don’t know.

4.5 Detection Optics

Fig 4.9 shows a schematic of the optical system used for detecting the ion fluorescence.

The objective was an f /1 six-element compound lens with a working distance of 40 mm. After passing through the intermediate aperture (600 μm diameter) in the image plane of the

objective, the image was further magnified by a UV doublet (2.5 cm focal length). The fluorescence was detected by either a Photek microchannel plate imaging tube (the “imager”) or a

Hamamatsu H6240-01 photomultiplier tube (PMT). We directed the fluorescence toward one or

the other detection device with a mirror mounted on an electronically controlled flipper. The

imager had quantum efficiency of about 5%, rather lower than the PMT quantum efficiency of

about 20%. However, the imager’s effective spatial resolution of several μm (measured in the

ion object plane) was invaluable during loading, as we could readily distinguish the number of

ions in the trap by lowering the endcap voltage somewhat. We used the PMT to collect all

the data presented in this thesis. Since the PMT had no spatial discrimination, we counted

the total number of photons arriving from the ion crystal rather than the number arriving from

each individual ion. Thus our data allowed us to distinguish only between the cases of zero, one,

two,... ions bright, which did not constitute a full readout of the state of the quantum register.

The experiments reported here did not suffer unduly from this restriction.

During the detection phase, off-resonant repumping and imperfect BD polarization caused
81

L1 (6-element)

aperture

L2 (doublet) 26 mm

40 mm w.d.

ION

baseplate

PMT

imager 

tube

movable

mirror

650 mm

Figure 4.9: Schematic of the detection optics. L1 is the objective lens, L2 the reimaging lens.
82

the spin state of each ion to relax toward equilibrium between the bright |↓i state and the rest

of the 2S1/2 manifold, which was dark. Typically an ion relaxed by 1% after scattering ∼ 104

photons (see Sec. 3.2). Hence the small detection f-number and high PMT quantum efficiency

were essential to our high spin discrimination efficiency.

4.6 Experiment Control and Data Acquisition

Each of our experiments was specified by a sequence of logical states of precisely defined

timing and duration. These logic states controlled the switch AOMs and the EOMs (through

the RF switchbox) and the frequency of the Raman beatnote (through the RF multiplexer).

We specified most of the experiment parameters using a Microsoft Visual C++ program running on a standard PC, which allowed us to reprogram the control sequence in a matter of a

few seconds. The Visual C++ program downloaded the control sequence by GPIB (IEEE 488)

interface to a Hewlett-Packard (HP) 8175 digital pulse generator, which produced independent

pulse sequences on up to 24 TTL channels with 20 ns time resolution. The Visual C++ program

also instructed a Stanford Instruments DG535 pulse generator to send out TTL pulses at the

repetition rate of the experiment. Each pulse triggered the HP 8175, beginning an experiment

cycle.

As described in Section 3.2, each shot of the experiment ended with a detection period in

which BD alone was turned on for several hundred μs. In each detection period, the TTL pulse

stream produced by the PMT was sent to a National Instruments AT-MIO data acquisition

card, which counted the number of TTL pulses for that shot and sent the result to a standard

PC. A National Instruments LabView program running on the PC built up a histogram and

graphically displayed it with about 1 s update time. In many experiments the LabView program
83

stepped the phase of one of the HP 3335 RF synthesizers and built up a histogram at each step.

Each histogram typically consisted of 1000 shots.

We used a different data acquisition procedure for two essential functions of the experiment: finding the spectrum of ion fluorescence as a function of two-photon detuning and

observing Rabi flopping on a particular Raman transition. For each function we could build up

good statistics in 10–100 shots, rather than the 1000 shots used for the histogram data. Hence

we could increase our data rate by stepping the independent variable on a timescale less than

a millisecond, rather than stepping the variable about every 3 ms, as required by the LabView

program. To build up consistent statistics, we had to update the independent variable in “real

time”, i.e., at a time specified to about 1 ms, the duration of a single shot. The frequent CPU

interrupts associated with Microsoft Windows would have resulted in severe update errors, so we

implemented the spectrum and flopping functions with separate PCs running under MS-DOS.

The number of pulses from the PMT was counted over a fixed time, usually about 100 ms, and

was read out to the PC.

Performing the spectrum function required that we apply a Raman pulse of a fixed duration at a fixed time in each shot of the experiment, but that we update the RR double-pass

frequency in real time. The duration and timing of the Raman pulse were specified by the

Visual C++ program through the HP 8175, while the spectrum PC determined the double-pass

frequency. A C program running on the spectrum PC allowed us to specify the spectrum sweep

parameters and graphically displayed the spectrum as it was acquired. The spectrum PC updated the double-pass frequency by GPIB commands to the HP 8660 synthesizer.
84

For the Rabi flopping function, we applied a Raman pulse of a variable duration at a fixed

double-pass frequency and a fixed time in each shot. Here the Visual C++ program specified

the double-pass frequency and the maximum length of time allowed for the pulse, while the

flopping PC determined the pulse duration. A C program running on the flopping PC specified

the sweep parameters and displayed the flopping curve. The flopping PC used a GPIB interface

to instruct another Stanford Instruments DG535 pulse generator to put out TTL pulses of the

desired duration. The TTL outputs of this “flopping” DG535 were connected to a set of logical

OR gates through an array of manual switches. The other inputs to the OR gates were TTL

outputs of the HP 8175. The outputs of the OR gates were input to the RF switchbox. Hence,

when the manual switch for a channel was on, the RF controlled by that channel switched on

when instructed by either the flopping DG535 or the HP8175. The Raman beams to be turned

on during the flopping pulse were specified by these manual switches.

For many of the experiments reported here, we needed to change the trap endcap voltage

multiple times within a single shot of the experiment. The voltage inputs to the trap endcap electrodes originated from the “DC switchbox”, which could switch between two possible

voltage states in a matter of a few microseconds. We controlled the switching either manually

or via a TTL input. The voltages of the two possible switchbox states were set by manual

potentiometers. The decoherence-free subspace experiment (Sec. 8) required switching between

three voltage states, which we implemented by summing the outputs of two identical DC switchboxes. We ensured smooth transitions between voltage states by passing the voltage output of

the switchbox arrangement through a 4-pole Bessel filter with cutoff frequency ∼ 1 MHz. This

filter attenuated high frequencies while preserving the phase relation between low-frequency

pulse components so as not to distort the pulse shape. The TTL inputs to the DC switchboxes

originated from either the DG535s or the HP8175. Directly connecting the TTL control signals

to the switchbox inputs allowed 60 Hz line noise to infiltrate the voltage inputs to the endcap
85

electrodes, degrading the two-ion logic gate (see Sec. 5). To achieve the < 5 mV stability required for the two-ion gate, we isolated the TTL inputs to the switchboxes with optocouplers.

For much the same reason, we electrically isolated the switchbox cases from the racks in which

the switchboxes were mounted. Isolating the endcap voltage from 60 Hz noise proved to be a

huge torment, and a future voltage multiplexing system must take serious precautions against

this problem.
Chapter 5

Multiparticle Entanglement

So far we have described initialization of our quantum register (Secs. 3.2 and 3.3.4),

operations on single qubits (Sec. 3.3.1), and detection of the register state (Sec. 3.2). However,

one more ingredient is needed to perform universal quantum logic: a gate that entangles two

particles [56, 57, 58]. This gate can be a quantum XOR [38, 59], a phase gate [2, 60, 61], or

any of a class of equivalent gates. We implemented the entangling gate of Mølmer and Sørensen

[62, 63] for strings of two and four ions [64], realizing the operations

|↓↓i → |↓↓i + |↑↑i (5.1)

|↓↓↓↓i → |↓↓↓↓i + |↑↑↑↑i

One can use this gate on two ions in conjunction with single-qubit operations to construct an

XOR gate [62], so this gate enables universal quantum logic. Of course, our experiment only

approximates the evolution (5.1). However, measurements on the states produced in the experiment showed that those states were, in fact, entangled. We used the two-ion version of this gate

as our fundamental entangling gate in the quantum logic experiments described in this thesis.
87

5.1 Theory of the Entangling Gate

To apply the Mølmer–Sørensen entangling gate to an ion crystal, one off-resonantly drives

both the red and blue sidebands of a particular motional mode of the crystal. The detunings

∆M from the sidebands are equal and opposite. Each ion is coupled to each sideband with the

same coupling strength, written Ωsb. The relevant level scheme for two ions is shown in Fig.

5.1. We obtain the coupling condition by keeping the intensity of all Raman beams constant

across the ion crystal and by choosing a motional mode in which all ions have equal motional

amplitudes, i.e., |v(k)

i | is independent of the ion index i (see Eq. 2.26.)

The essential features of the entangling gate are intuitively clear in the far-detuned limit

∆M À Ωsb. Considering only two ions for clarity, we see that the intermediate spin state

|↓↑i + |↑↓i is only virtually populated, while the |↓↓i ↔ |↑↑i transition is resonant. The singlet

spin state |↓↑i − |↑↓i does not take part in the dynamics. The situation is reminiscent of

resonant two-photon excitation through a virtual intermediate level [65], as in, e.g., spectroscopy

of the rubidium 5S ↔ 5D transition. However, the “two-photon resonance” here corresponds to

simultaneous excitation of two ions, rather than excitation of a high-lying level of a single atom,

and the off-resonant “single-photon” couplings are actually Raman processes. Exploiting the

analogy to two-photon spectroscopy, we see that the Rabi frequency for the pathway through

the intermediate |n + 1i (|n − 1i) state is proportional to (n + 1)Ω2

sb/∆M (nΩ2sb/∆M) in the

Lamb-Dicke limit. The factors of n come from the harmonic oscillator raising and lowering

operators as in Eq. (3.22). The two paths interfere destructively, so the overall Rabi frequency

for the |↓↓i ↔ |↑↑i transition is proportional to Ω2

sb/∆M. In the Lamb-Dicke limit, the Rabi

frequency is independent of n, so the gate is insensitive to the initial motional state of the ion

crystal! In view of the heating problem in our traps (Sec. 7.4), this feature gives the Mølmer–

Sørensen gate a distinct advantage over the Cirac-Zoller XOR [38], which requires that the ion
88

↓↓

↑↑

↑↓ + ↓↑

n

n+1

n-1

∆M

n

n+1

n-1

n

n+1

n-1

Figure 5.1: Relevant level scheme for the two-ion version of the entangling gate. The arrows

represent Raman processes driven by pairs of Raman beams. The two pathways through the

intermediate |↓↑i + |↑↓i state exhibit quantum interference.
89

crystal start in a Fock state of the motion.

It takes a relatively long time (À 1/Ωsb) to apply the entangling gate in the far-detuned

limit, giving decoherence a long time to act and severely limiting the gate fidelity. However,

as we will see, one can obtain the evolution (Eq. 5.1) even if ∆M is on the order of Ωsb, for

carefully chosen values of ∆M. In the experiment we used this “nonperturbative” gate operation

to speed up the gate, allowing high-fidelity gate operation.

To analyze the nonperturbative regime, we return to the basic interaction Hamiltonian

consisting of the two off-resonant Raman drives. Transforming to an interaction picture with

respect to the free evolution of the ion spin and motion, and assuming the Lamb-Dicke limit,

the Hamiltonian for N ions takes the form [66]

HM = (iηkΩsbJ+a†e−i∆Mt + h.c.) + (iηkΩsbJ+ae−i∆Mt + h.c.) (5.2)

= −ηkΩsbJy[ae−i∆Mt + h.c.] (5.3)

where the pair of Raman beams driving the blue (red) sideband has relative phase φj,B (φj,R)

at ion j and ηk is the Lamb-Dicke parameter for the logic mode (see Eq. 3.28). In Eq. (5.3) we

have omitted the phases φj,B, φj,R, but we can insert them again later by the transformations

Sj,+ → eiφjSj,+ a → e−iδφa (5.4)

φj ≡ (φj,B + φj,R)/2 δφ ≡ (φj,B − φj,R)/2 (5.5)

Here δφ does not depend on j.

The full dynamics under the Hamiltonian (5.3) is given by the evolution operator UM

satisfying the Schr ̈odinger equation i dUM/dt = HMUM. Given an initial state |ψ(t = 0)i, the
90

state at any later time is |ψ(t)i = UM(t)|ψ(0)i. Following [66], we make an ansatz for UM of

form

UM = exp[ik(t)J2

y] exp[h(t)Jya†] exp[−h∗(t)Jya] (5.6)

where k(t), h(t) are c-number functions that govern the dynamics. Substituting this ansatz into

the Schr ̈odinger equation for UM, we find

k(t) = η2

kΩ2sb

∆M

t − i(e−i∆Mt − 1) μηkΩsb

∆M

¶2

(5.7)

h(t) = ηkΩsb

∆M

(ei∆Mt − 1) (5.8)

The first factor in Eq. (5.6) describes the entanglement of the ions through “two-photon” transitions. The other two factors describe the virtual red- and blue-sideband transitions induced by

the Raman beams. From a mathematical standpoint, we can regard the first factor as a kind of

geometric phase factor generated by the noncommutativity of the blue and red sideband drives.

More physically, we can say that the virtual excitation of the motional mode carries quantum

information between the ions.

While the evolution (5.6) creates entanglement for any choice of ∆M, we want to entangle

the ion spins only with each other, not with the motional mode. Thus we want to eliminate the

last two terms of Eq. (5.6), which we can do if h(t) = 0. Moreover, the gate operation (5.1)

corresponds to the unitary operator exp[iπJ2

y/2], so we require k(t) = π/2. These constraints

are satisfied for

∆M = 2ηkΩsb√p t = π

√p

ηkΩsb

(5.9)

with p, j integers. Hence we can operate the gate (5.1) in a time as short as π/(ηkΩsb), if we

choose ∆M properly. This time is on the order of a sideband π-pulse time, so the sideband
91

Rabi frequency sets the clock speed for our quantum register. We require Ωsb . ωz; otherwise

the gate drive also drives off-resonant transitions that spoil the gate. Hence the trap frequency

ultimately limits the clock speed.

We can use the entangling gate to generate the state |↓Ni+|↑Ni with a single laser pulse,

which greatly simplifies the use of entangled states for spectroscopy [67]. In contrast, N two-bit

XORs are required to generate the same state [38]. Moreover, the time required to entangle N

ions with our gate scales as 1/η2

k ∝ N, so our gate is no slower than the usual XOR method.

The simplest initial state for which the entangling gate produces nontrivial evolution is

the two-ion state |ψ(0)i = |↓↓; n = 0i, where n is the number of phonons in the logic mode.

Using Eq. (5.6), we find the normalized state

|ψ(t)i = 1

2 (|↓↓; 0i + |↑↑; 0i) + 12

eik(t)

(|↓↓; 0i − |↑↑; 0i) (5.10)

− 1

2

eik(t) X∞

n=1

(h(t))2n

p(2n)! (|↓↓; ni − |↑↑; ni) (5.11)

− i

2

eik(t) X∞

n=0

(h(t))2n+1

p(2n + 1)!(|↓↑; ni + |↑↓; ni) (5.12)

The time-evolved state divides into a pure spin state, which undergoes no motional evolution,

and a complicated spin/motion entangled state. For h(t) = 0 the spin/motion entanglement

vanishes, leaving residual spin entanglement described by the first two terms. However, in

general we may have a large correlation between spin and motion during gate operation, which

can be affected by motional decoherence (see Sec. 7.4). Half the population stays in the initial

motional state during the gate. Due to the virtual excitation of the logic mode, the rest of the

population is spread over a Poissonian distribution of motional Fock states
92

ρn,n = e−bbn

n! 2 (5.13)

b(t) = |h(t)|

2 = 4 μηkΩsb

∆M

¶2

sin2 ∆Mt (5.14)

For p = 1, b(t) = sin2 ∆Mt. Even in the worst case of p = 1, the spin-motion entangled state

never has motional spread greater than about one quantum. Hence we expect the gate to be

robust under motional decoherence, as discussed in more detail in Sec. 7.3.

5.2 Experimental Results

We first implemented the entangling gate (5.1) on strings of two and four ions held in

linear trap #1. In each case, we had to select a motional mode for logic that coupled equally

to the ions. The obvious choice was the COM mode in each case, but because of the large

heating rate of the COM mode, we selected higher-order modes in both cases. For the two-ion

experiments we used the (axial) stretch mode with frequency √3 ωz, while for four ions we used

the (axial) stretch 2 mode with frequency ≈ p29/5 ωz (see Table 2.2). In general, the only

mode with equal coupling is the COM mode, so the equal coupling of the stretch 2 mode of four

ions was a lucky coincidence for us.

The entangling gate required coupling on two Raman transitions. Rather than using

two pairs of Raman beams to generate the two Raman difference frequencies, we modulated

RR using a resonant EOM at frequency ωEOM (see Sec. 4.3.2). The R2 carrier frequency ωR2

beat with the B1 frequency ωB1 to give a beatnote near the blue sideband, while one of the

EOM-generated sidebands of R2 beat with B1 to give a beatnote near the red sideband. Thus

we set the RR double-pass frequency to give ωent ≡ ωB1 − ωR2 = ω↓↑ + ωlogic − ∆M and the
93

EOM frequency to ωEOM = 2(ωlogic − ∆M), where ωlogic was the frequency of the logic mode.

For two ions, we set ωCOM = 5.0 MHz, so ωlogic = 8.7 MHz. We used ωEOM = 17.67 MHz to

allow for the detuning of the entangling transition from the sidebands. For four ions, we had to

reduce ωCOM to avoid the zigzag instability (Sec. 2.1.3). We chose ωCOM = 3.6 MHz, so that

ωlogic = 8.7 MHz again and we could use the same EOM frequency.

To ensure equal illumination of each ion, we used beam spot sizes at least twice as large

as the extent of the ion crystal. As shown in Sec. 7.3, deviations from this condition affect the

gate fidelity only to second order. To observe loss of fidelity due to unequal illumination, we

had to offset the beams by an amount much larger than the estimated fluctuations in beam

pointing, so we expect this source of error to be small.

To provide equal coupling to each sideband, we set the EOM power so that the intensities

of the EOM sidebands were equal to the intensity of the (depleted) R2 carrier. In this setup,

only the EOM carrier and one EOM sideband were useful for driving the entangling gate. The

rest of the R2 power, which amounted to about half the R2 power before the EOM, contributed

to the spontaneous emission rate (see Sec. 7.2.2) but was wasted for purposes of quantum logic.

Flopping at ωB1−ωR2 = ω↓↑+ωlogic with the EOM on gave the π-pulse time of the blue sideband

as driven by the EOM carrier, while flopping at ωB1−ωR2 = ω↓↑+ωlogic+ 2∆M gave the π-pulse

time of the red sideband as driven by the red EOM sideband. Adjusting the EOM power so

that these times were equal ensured the desired condition. For cold ions, the red sideband signal

was small, so we sometimes flipped the ion spins with a carrier π-pulse before taking the second

flopping curve. To observe loss of fidelity from unequal sideband coupling, we had to change

the EOM power by about a factor of 3 from the optimum, so we expect this source of error to

be small as well.
94

Since we operated the gate in the nonperturbative regime, the behavior of the ions under

the entangling drive depended strongly on ∆M. Typically we set ωdp to detune the Raman

drives symmetrically from their sidebands and tuned ∆M by changing the trap endcap voltage,

preserving the symmetric detuning. We obtained flopping curves for various values of ∆M.

Fig. 5.2 shows experimental (dots) and theoretical (solid) curves for two ions. Here ∆M satisfies

Eq. (6.3) for p = 1, and the theoretical fluorescence is A(2ρ↓↓,↓↓ + ρ↓↑,↓↑ + ρ↑↓,↑↓) from Eq. (5.6)

with A a scaling factor. We multiply the theory curve by an exponential decay with time

constant ∼ 30 μs to account for decoherence during the gate.

Fig. 5.3 shows theoretical flopping curves at p > 1 for the conditions of Fig. 5.2. Here

we parametrize ∆M in terms of Eq. (6.3) by allowing noninteger p. The curves exhibit a slow

oscillation, corresponding to the spin evolution under the first term of Eq. (5.6), with superimposed faster oscillations arising from the last two terms of Eq. (5.6). In Fig. 5.3(a) ∆M satisfies

Eq. (6.3) with p = 2, so the frequency of the fast oscillation is 4 times that of the slow oscillation. In contrast, Fig. 5.2(b) shows a flopping curve for which p = 1.6, so ∆M does not satisfy

Eq. (6.3). The shape of this curve is significantly deformed from those of Figs. 5.2 and 5.3(a).

The flopping curves for four ions at a given value of p were similar to those for two ions.

To operate the entangling gate at p = 1, we tuned ∆M so that the frequency of the fast

oscillation was twice that of the slow oscillation, giving a reasonable approximation to Fig. 5.2.

We then took a flopping curve, fixed the pulse duration at the point of minimum fluorescence,

corresponding to k(t) ≈ π, and varied the endcap voltage (tuning ∆M) to reduce the fluorescence

further. Repeating this last step once or twice was usually sufficient to optimize ∆M. Reducing

the pulse duration by half, we had k(t) ≈ π/2, producing the evolution (5.1).
95

pulse duration [μs]

fluorescence [arb.]

8000

6000

4000

2000

0

0 5 10 15 20 25 30

Figure 5.2: Experimental flopping curve for p = 1 (dots) with theoretical prediction (line) with

exponential decay of time constant ∼ 30μs and with ions in the motional ground state.
96

1 2 3 4

0.2

0.4

0.6

0.8

1

1 2 3 4

0.2

0.4

0.6

0.8

1

(a)

p = 2

(b)

fluorescence [arb.]

p = 1.6

t /(2π∆M )

Figure 5.3: Theoretical flopping curves for (a) p = 2 and (b) p = 1.6, calculated for the same

conditions as in Fig. 5.2. The value of p dramatically affects the shape of the flopping curve.
97

We obtained lineshapes at fixed values of ∆M by sweeping ωdp. The linewidth of the

entangling transition was a few tens of kHz and decreased with increasing ∆M. Because the

R2 power was divided among the sidebands by the EOM, the usable R2 power in either of the

Raman drives was about 1/4 the total R2 power, decreasing the Rabi frequency. The small

linewidth made the entangling gate especially sensitive to perturbations of ω↓↑, as detailed in

Sec. 7.3.

The spin population of the states produced by the evolution (5.1) is nominally equally

divided between |↓Ni and |↑Ni for N ions. Fig. 5.4 shows a histogram of count rates obtained

after application of the entangling gate. As expected, the histogram shows a bimodal distribution of count rate. Fitting the populations for this histogram gave probabilities P0 = 0.43,

P1 = 0.11, P2 = 0.46, where Pn is the probability to find n ions in |↓i. The error in each value

is ±0.01. The populations are quite close to the ideal case P0 = 0.50, P1 = 0.00, P2 = 0.50.

For four ions we typically found P0 ≈ 0.35, P1 ≈ P2 ≈ P3 ≈ 0.10, P4 ≈ 0.35. This distribution

implies that quite a bit of the population resides outside the desired state. However, there are

14 orthogonal states that give one, two, or three ions in |↓i, and only two orthogonal states in

the desired entangled state, so the distribution of population over the 16 orthogonal states is

far from random.

Ideally, the evolution (5.1) produces full coherence between the |↓Ni and |↑Ni states, so

that |ρ↓N,↑N | = 0.5. To measure the coherence produced in the experiment, we “interfered” the

|↓Ni and |↑Ni states using a variant of the method proposed in [67]. After the entangling gate,

we applied a carrier π/2 pulse (Eq. 3.26) with Raman phases φj = φ for all j. Here we have

taken the phases φj = 0 for the entangling pulse. We then detected the number N↓ of ions in
98

0 20 40 60 0

0.1

0.2

0.3

0.4

number of photons

probability

Figure 5.4: Histogram of photon counts obtained after application of the entangling gate.
99

|↓i and compute the parity Π(φ) ≡ (−1)N↓ . Assuming an ideal carrier pulse, we have

Π(φ)=2|ρ↓N,↑N | cos Nφ (5.15)

so we could extract the desired coherence by measuring the parity as a function of phase. To

observe the parity fringes of Eq. (5.15), we had to set the synthesizer frequency for the π/2 pulse

to ωent−ωEOM/2 (in Hz) with accuracy better than ±1 Hz. This condition ensured that the beat

frequency for the π/2 pulse was exactly midway between the two beat frequencies used for the

entangling transition, so that the entangling pulse and the π/2 pulse were mutually coherent.

Any error in setting the frequency of the π/2 pulse caused phase slippage of φ, reducing the

contrast of the parity fringes.

To obtain the interference data, we took histograms of the photon counts at each value

of the phase φ over several oscillations of the sinusoid Eq. (5.15). Averaging all the histograms

together and fitting the result determined the mean count number per ion and the mean number

of background counts. From this information and our knowledge of the repumping parameter

αp (Eq. 3.1) we constructed reference histograms for N↓ = 0, 1,...N. Fitting the individual

histograms to the references gave the probabilities Pn(φ) to find N↓ = n at phase φ, from which

we could trivially infer Π(φ). Fig. 5.5 shows results of interference experiments for two and four

ions. Each curve is fit with a single sinusoid with frequency fixed by Eq. (5.15) to obtain the

amplitude. The presence of other oscillation frequencies in the interference data would imply

coherences involving less than N ions. Since the fits to the data are good, we infer that ρ↓N,↑N is

by far the largest off-diagonal element of the density matrix in each case.1 The data of Fig 5.5

show |ρ↓N,↑N | = 0.39 ± 0.01 for two ions and 0.22 ± 0.02 for four ions. Though these values are

significantly less than the ideal value of 0.50, we will see that the experimental states are still

1 In principle, other coherences could be present, but the interference data would constrain these coherences

to have particular phase relationships that seem physically unlikely.
100

0 90 180 270 360

-1

-0.5

0

0.5

1

0 90 180 270 360 -1

-0.5

0

0.5

1

(a) (b)

Parity

Phase [deg] Phase [deg]

Figure 5.5: Interference data for (a) two and (b) four ions. Dots are data points, solid lines are

sinusoidal fits with frequencies fixed by Eq. (5.15). The fit amplitudes show |ρ↓N,↑N | = 0.39±0.01

for two ions and 0.22 ± 0.02 for four ions.

entangled.

We can easily establish the separability or entanglement of the experimental density

matrix ρ by computing its fidelity F = hψent|ρ|ψenti against the ideal state ψent = (|↓Ni +

|↑Ni)/

√2. To show this, consider a general register state |ψF i that is factorizable by dividing

the register into two subsets X and Y. We can then write [64]

|ψF i = (a|↑ ... ↑i + b|↓ ... ↓i + ...)X ⊗ (c|↑ ... ↑i + d|↓ ... ↓i + ...)Y (5.16)

where |↑ ... ↑iX has all spins in X in the state |↑i. Normalization of the states over X and Y

requires |a|

2 + |b|2 ≤ 1, |c|2 + |d|2 ≤ 1, so we have

2(|ac| + |bd|) ≤ (|a|−|c|)

2 + (|b|−|d|)2 + 2(|ac| + |bd|) (5.17)

= |a|

2 + |b|2 + |c|2 + |d|2 ≤ 2 (5.18)

Writing ρF = |ψF ihψF | and making identifications ρF

↑N,↑N = |ac|

2, etc., we find

(|ac| + |bd|)

2 = ρF

↓N,↓N + ρF↑N,↑N + 2|ρF↓N,↑N | = 2hψent|ρF |ψenti ≤ 1 (5.19)
101

which holds for any factorizable, i.e., non-entangled state. If the density matrix ρ is factorizable

over X,Y, we can write it as a sum of pure states factorizable over X,Y. Hence any factorizable

density matrix satisfies F = hψent|ρ|ψenti ≤ 1/2. This condition is sufficient but not necessary

to show entanglement [68]. By measuring the populations ρ↓N,↓N , ρ↑N,↑N and the coherence

|ρ↓N,↑N |, we established [64]

F = 0.83 ± 0.01 N = 2 (5.20)

F = 0.57 ± 0.02 N = 4 (5.21)

showing that the entangling gate can indeed produce entangled states of two and four ions.

For the two-qubit state, we can approximate the entanglement of formation from the

measured density matrix elements as E ≈ 0.5 [69]. This quantity measures the increase in the

entropy of one qubit produced by tracing over the other qubit. Ideally, tracing over one qubit

leaves the other in a mixture of half |↓i and half |↑i, with no coherence remaining between |↓i

and |↑i. In this case E = 1. The entanglement of formation also gives an upper bound on

the entanglement of distillation ED [70, 71]. Leaving out the technical details, any procedure

designed to distill a perfect EPR state |↓↑i − |↑↓i from several copies of a partially entangled

two-qubit state requires an average of 1/ED copies [72]. The complexity of quantum communications tasks is often quantified in terms of the number of EPR states required, so ED measures

the usefulness of the experimental state ρ. We see that order of two copies of our two-qubit

state ρ are required to distill one EPR state.

There is so far no simple way to characterize the amount of entanglement of a general

four-qubit state. However, the absence of coherences other than ρ↓4,↑4 in the interference data
102

suggests that we can write the experimental four-ion density matrix as [64]

ρ ≈ 0.43|ψentihψent| + 0.57ρdiag (5.22)

where ρdiag has no off-diagonal elements. Thus, not only is the total density matrix entangled,

but any one shot of the experiment has a significant chance of producing the desired state.

The high fidelity of the two-ion entangled state encouraged us to use the evolution (5.6)

with p = 1 and k(t) = π/2 as our basic two-qubit gate. The experiments of chapters 7 and 8

used precisely this gate for simple quantum logic. Reinstating the phases φj , we can write the

normalized gate operator as

Uent(φ1, φ2) = exp[iπJ2

y/2] (5.23)

→ 1 + i

2

h

1 − i

3

ei(φ1+φ2)

|↑↑ih↓↓| + e−i(φ1+φ2)|↓↓ih↑↑| (5.24)

−ei(φ1−φ2)

|↑↓ih↓↑| − e−i(φ1−φ2)|↓↑ih↑↓| ́i (5.25)

In the following, we describe our two-ion logic circuits as products involving Uent(φ1, φ2) and

the normalized carrier evolution

Ucar(θ, φ1, φ2) = O

j=1,2

·

1 · cos

θ

2 − isin θ2 (eiφj |↑ih↓| + e−iφj |↓ih↑|)

 ̧

(5.26)

produced by driving the carrier (Eq. 3.26) on resonance for a time t = θ/ΩR.
Chapter 6

Applications of Entanglement

6.1 Violation of a Bell Inequality

Perhaps the most striking feature of quantum mechanics is its denial of local realism. The

uncertainty principle requires us to believe that measuring the position of an object can change

its momentum, but is still consistent with the basic notion of realism: that the object had some

position or other before the measurement, which is then perturbed by the act of measurement.

At the same time, special relativity insists on locality: the measured position does not depend

on events outside the light-cone of the object. As shown by Einstein, Podolsky, and Rosen [73],

quantum mechanics is incomplete if we assume local realism. Bell [74] proved that quantum

mechanics is actually inconsistent with local realism, deriving an inequality on observables that

is satisfied by all local realistic theories but is violated by quantum mechanics. Many experiments [75, 76, 77, 78, 79, 80] have observed violation of a Bell inequality, so that local realistic

theories are now discounted by the vast majority of physicists. The result of this reasoning is

epistemologically remarkable. One must either accept that “objective reality” is created only

at the moment of measurement, or give up special relativity! Mainstream physics has opted to

retain special relativity, probably a good choice in view of the massive experimental evidence in

favor of quantum field theory. Moreover, the philosophical tradition of Mach and, later, Heisenberg and Bohr, restricts the domain of physics to properties susceptible to measurement, so one

can dispense with realism in a Machian theory. The alternative view of Bohm [81, 82] retains
104

realism at the expense of locality, introducing a “pilot-wave” field that coordinates the behavior

of particles at widely separated points.

Clauser, Horne, Shimony, and Holt (CHSH) [83] proposed an experimental arrangement

to test a form of Bell’s inequality quantitatively. Here a “black box” prepares a pair of particles.

The ith particle (i = 1, 2) enters a measurement apparatus whose properties are parametrized by

an independent classical variable Φi. Each (classical) measurement yields a value Mi(Φi) = ±1

depending on the parameter Φi and the state of the correlated particles. Averaging over many

measurements, one extracts the correlation function q(Φ1, Φ2) = hM1(Φ1)M2(Φ2)i. Then the

CHSH form of the Bell inequality

B(α1, δ1, β2, γ2) ≡ |q(δ1, γ2) − q(α1, γ2)| + |q(δ1, β2) + q(α1, β2)| (6.1)

≤ 2

holds for any values α1, δ1 (resp. β2, γ2) of Φ1 (resp. Φ2) under the assumption of local realism

[83]. The derivation of Eq. (6.1) involves the quantity hM1(α1)M1(δ1)M2(β2)M2(γ2)i. Since the

terms M1(α1), M1(δ1) refer to different arrangements of the experiment for α1 6= δ1, quantum

mechanics forbids us to speak of both terms at once. Hence the product M1(α1)M1(δ1) is not

well-defined in quantum mechanics, allowing violation of Eq. (6.1).

In our experiment, the entangling gate played the role of the CHSH “black box”, preparing

two ions in the state |↓↓i + |↑↑i. To measure the correlation function q(Φ1, Φ2), we applied a

carrier rotation Ucar(θ = π/2, Φ1, Φ2) and built up a histogram of photon counts. Counting the

number of events N0 (N1, N2) with zero (one, two) ions bright, we find
105

q(Φ1, Φ2) = (N0 + N2) − N1

Ntot

(6.2)

To measure B(α1, δ1, β2, γ2), we obtained a value of q for each term in Eq. (6.1). We chose α1,

etc., so as to give the maximum violation of Eq. (6.1) by the quantum-mechanical prediction for

B, which occurs for

α1 = −π/8 δ1 = 3π/8 (6.3)

β2 = −π/8 γ2 = 3π/8

For these angles we have the prediction B = 2√2, contradicting the bound (6.1) set by local

realism.

Independently setting Φ1 and Φ2 for the carrier rotation required positioning each of the

two ions at a specific point in the interference pattern of the Raman beams (modulo the wavelength of the interference pattern). Changing the phase of the RF driving the RR double-pass

AOM shifted the interference pattern by a constant amount, allowing us to control the quantity

φtot = φ1 + φ2. However, the phase difference ∆φ = φ1 − φ2 is set by the separation of the

ions ∆z along ˆz as ∆φ = √2k∆z. To control ∆φ without changing φtot, we varied the endcap

voltage, and therefore the trap strength, so that the ions moved symmetrically about the center

of the trap. Hence the endcap voltage and the RF phase jointly determined the phases Φ1, Φ2.

We repeated the experiment Ntot = 20 000 times for each of the four sets of phases given

by Eq. (6.3). These data are presented as histograms in Fig. 6.1. Shelving the |↑i state (see

below) allowed us to extend the duration of the detection pulse to 1 ms, increasing the number

of photons detected from each bright ion and reducing the error due to photon shot noise. To
106

Table 6.1: The correlation values and resulting Bell’s signals for five runs of the experiment.

The phase angles α1, δ1, β2, γ2 are given by Eq. (6.3). The statistical errors are 0.006 and 0.012

for the values of q and B respectively.

q(α1, β2) q(α1, γ2) q(δ1, β2) q(δ1, γ2) B(α1, δ1, β2, γ2)

0.541 0.539 0.569 -0.573 2.222

0.575 0.570 0.530 -0.600 2.275

0.551 0.634 0.590 -0.487 2.262

0.575 0.561 0.559 -0.551 2.246

0.541 0.596 0.537 -0.571 2.245

determine the values N0, N1, N2 entering Eq. (6.2), we divided the histograms into cases of zero,

one, or two ions bright by setting discriminators at the positions of the arrows in Fig. 6.1. This

procedure gives the same value of q that we would obtain by measuring M1, M2 individually for

each shot of the experiment.

We measured the Bell signal five times to estimate the variation between runs. The five

sets of correlation functions and the resulting Bell signals are shown in Table 6.1. The statistical

error of each Bell signal was ±0.01. Averaging the Bell signals together, we find

B(−π

8

,

3π

8 , −π8

,

3π

8 )=2.25 ± 0.03 (6.4)

an 8σ violation of the Bell inequality (6.1).

The error in the average Bell signal significantly exceeds the statistical error predicted

from Table 6.1. We attribute this discrepancy to drifts of the RF phase between the HP 8660

synthesizer used to drive the entangling gate and the HP 3335 used for the carrier rotation.

Such drifts induce errors in setting the angles α1, etc., for the four measurements of q, but these

errors cannot be estimated from the statistical errors in q. The phase drift was usually constant
107

a1, b2 d1, c2

a1, c2 d1, b2

0 50 100 150 200 0 50 100 150 200

50

40

30

20

10

50

40

30

20

10

Number of Instances

Number of Photons

Figure 6.1: Histograms of fluorescence signals for the four sets of phases given in Eq. (6.3). The

vertical arrows break up the data into cases of zero, one, or two ions bright with 98% accuracy.
108

for a few minutes at a time, but could change magnitude and direction rapidly and randomly.

The drift sometimes amounted to several tens of mrad per minute, while about 15 minutes were

required to make the four measurements of q needed for one measurement of B. Including the

drift in the error analysis for each measurement of B approximately accounts for the statistical

error in the average value of B.

In this experiment, we attempted to falsify local realism by assuming local realism and

arriving at a contradiction, namely B = 2√2 ≤ 2. But Φ1, Φ2 were hard to measure directly,

since they varied rapidly if an ion moves even a few nm. To ensure the logical consistency of

our procedure, we calibrated φtot and ∆φ in terms of macroscopic quantities by using purely

classical reasoning. The RF phase, a macroscopic quantity, directly set φtot by controlling the

(classical) interference pattern of the Raman beams. The endcap voltage, another macroscopic

quantity, set ∆φ indirectly by changing the ion separation. To calibrate ∆φ, we first observed

that applying a pair of pulses with fixed duration and relative phase φ to a single ion caused

the ion fluorescence to vary as C + A cos φ. Here we set φ using the (classical) RF phase, and

the pulse duration and frequency were set to maximize A. (In quantum-mechanical terms, the

pulse duration gave a π/2 rotation of the ion, but our procedure does not rely on this fact.) We

deduced that applying similar pulses to two ions would produce fluorescence

C + A cos φ1 + C + A cos φ2 = 2C + 2A cos

φtot

2

cos

∆φ

2 (6.5)

since the detector is linear. Applying the pulses and using the RF phase to sweep φtot at a

fixed endcap voltage produced a sinusoidal signal of amplitude 2A cos(∆φ/2). Repeating the

experiment for a range of endcap voltages and fitting the amplitudes allowed us to extract ∆φ

as a function of endcap voltage. Since the phase of the sinusoidal signal did not change as a

function of endcap voltage, we could be sure that φtot depended only on the RF phase. Hence

the phases Φ1, Φ2 were completely determined by the classical parameters of RF phase and
109

endcap voltage.

Up to now, all experiments showing violation of Bell’s inequality have had quite low detector efficiencies, so that the number of particle pairs produced by the “black box” far exceeded

the number of events used for calculating the correlation function. To interpret these experiments as ruling out local realism, one assumes that the correlated pairs that are not detected

exhibit the same behavior as those that are detected, the fair sampling hypothesis. However, by

violating this assumption it is possible to construct local realistic theories that reproduce the

experimental data for detector efficiencies below a critical value, the so-called detection loophole

[84, 85]. In these theories, the full set of particle pairs satisfies Bell’s inequality, but the subset

of pairs that is actually detected appears to violate it. While such a conspiracy of detectors

seems quite unlikely, no previous experiment has been able to rule it out.

The experiment described here was the first to demonstrate a violation of Bell’s inequality

without use of the fair sampling hypothesis. Because we could prepare the entangled state on

demand, we could make a measurement on every correlated pair. Thus there was no sampling

involved in our experiment at all. In this situation, quantum mechanics predicts that uncorrelated errors in measuring the two particles simply lower the value of B. We estimate that we

would have observed B = 2.37 in the absence of detection errors.

The overlap of the histograms in Fig. 6.1 produced a detection error rate of 2% in discriminating between the cases of zero, one, and two ions bright. Detection error in our experiments

frequently arises from off-resonant repumping of the |↑i state to |↓i (see Sec. 3.2). This repumping can occur even if the |↑i ion scatters only one BD photon. However, atomic selection rules
110

ensure that an ion initially in the 2S1/2 |F = 1, mF = 1i state must scatter at least three BD

photons to arrive at |↓i. The repumping rate from 2S1/2 |F = 1, mF = 1i to |↓i is therefore

lower than that from |↑i to |↓i. We took advantage of this fact to reduce our detection error.

After the carrier rotation, but before detection, we used two Raman π−pulses to transfer the

population in |↑i to 2S1/2 |F = 2, mF = 0i and thence to 2S1/2 |F = 1, mF = 1i. We implemented this “shelving” sequence with the copropagating beams B1, R1. The Zeeman splitting

between the shelving transitions exceeded the bandwidth of the RR double-pass AOM, so we

used the shelving EOM (see Sec. 4.3.2) to access the shelving transition frequencies.

To falsify local realism using Bell’s inequality, one must ensure that the only correlations

between M1(Φ1) and M2(Φ2) arise from the correlations of the entangled pair. However, if

the measuring devices can communicate during the measurement, i.e., if the measurement processes lie within each others’ lightcones, we can attribute any violation of Bell’s inequality to a

conspiracy between the measuring devices rather than to any violation of local realism by the

entangled particles. Tests of Bell’s inequality using photons have gone to great lengths to close

this “locality loophole” [78, 79, 80]. In our experiment we could not ensure nonlocality because

of the small separation between the ions. However, there is no physical reason to believe that

the measurement outcomes influence each other in this way. We estimated the size of correlations arising from a wide variety of effects, including dipole-dipole interactions between the ions,

interference between the radiation patterns of the ions, etc. [86, 87] In all cases the spurious

correlations are negligible, but strictly speaking we have not closed the locality loophole. Similarly, the photon experiments [78, 79, 80] do not close the detection loophole, though there is

no physical reason to believe in a conspiracy between detectors. In this sense our experiment is

complementary to the photon experiments.
111

This experiment gives the first violation of the Bell inequality for massive particles obtained using a complete set of correlation measurements. A previous experiment using protons

[76] deduced a violation from an incomplete set of measurements using quantum-mechanical

assumptions. Furthermore, our data uses the outcome of every shot of the experiment, so the

violation is obtained without the use of the fair-sampling hypothesis, closing the detector loophole for the first time. However, the “locality loophole” remains open for our data. Since the

detection events on the two ions occur within each other’s lightcones, in principle the detections

could influence each other, leading to spurious correlations.

6.2 Interferometry below the Standard Quantum Limit

6.2.1 N Views of Ramsey Spectroscopy

Frequency measurements using atomic clocks are currently the most accurate measurements of any physical quantity [88, 89, 90]. These clocks measure the transition frequency

between two electronic levels |↓i, |↑i of an atomic sample to derive a stable reference for a local

oscillator (LO). The LO usually interrogates the transition by Ramsey spectroscopy [91]. The

simplest variant of this method consists of three steps: (1) the LO applies a π/2 rotation to the

atoms, (2) the atoms undergo free evolution for a time TR, and (3) the LO again applies a π/2

rotation. In the frame rotating at the LO frequency, the free evolution (step 2) causes the part

of the atomic state in |↑i to acquire a phase TRωR relative to the part of the state in |↓i. Here

ωR is the difference between the LO and the atomic frequencies. After step 3, the average value

of the collective spin operator Jz on N atoms is

hJzi = N(1 − cos TRωR)/2 (6.6)

so that we can measure the frequency difference ωR by measuring the electronic state of the
112

atoms.

While Ramsey spectroscopy is traditionally described as a measurement of the frequency

ωR, this frequency can fluctuate during the interval TR. In this case we find the Ramsey signal

hJzi = N(1 − cos φR)/2 (6.7)

φR ≡

Z TR

0

ωR(t

0

) dt0 (6.8)

so we see that the signal really measures the accumulated relative phase φR between the |↓i and

|↑i states, in the same way that an optical interferometer measures the relative phase between

two optical paths. Now representing the atomic spin state on the Bloch sphere, we see that

the relative phase is just the azimuthal angle of the spin vector [92, 36], so that Ramsey spectroscopy measures the rotation angle of the spin vector around the ˆz axis. Using the Schwinger

spin representation for the two modes of an optical interferometer [93], we see that optical path

lengths correspond to rotations about the ˆz axis, while 50/50 beamsplitters are equivalent to

π/2 pulses [94]. Performing Ramsey spectroscopy is equivalent to operating a Mach-Zehnder

interferometer [95] with Fock-state input to one port, as shown graphically in Fig. 6.2.

While Ramsey spectroscopy measures only the rotation angle φR about the ˆz axis, we can

use atomic spin operators more generally to characterize an arbitrary rotation of our coordinate

system by an angle ζ about an axis ˆu, as depicted in Fig. 6.3. Such a rotation implements the

operator

R(ˆu, ζ) = exp h−iζ*

J · uˆ

i

(6.9)

on the atomic spin system. Since the state of the atomic system |Ψ(ζ)i after the rotation depends
113

accumulate 

phase

rotate

↓ ↓ + ↑ iφ e

input rotate detect

“ “ ↓

“ “ ↑

Figure 6.2: Equivalence of Ramsey spectroscopy and Mach-Zehnder interferometry. In the

Schwinger representation of the optical modes, 50/50 beamsplitters are π/2 rotations, while free

propagation produces rotations about ˆz.
114

on the classical parameter ζ, we can try to estimate ζ by measuring an atomic observable O.

We define the expectation value of O over the state |Ψ(ζ)i as

hOiζ = hΨ(ζ)|O|Ψ(ζ)i (6.10)

Then if a measurement of the operator O yields the result O, we find the estimated value ζ by

inverting the equation hOiζ = O. The uncertainty of the estimate δζ (measured in radians) is

δζ = ∆O

|∂hOiζ /∂ζ| (6.11)

where (∆O)2 gives the variance in repeated measurements of O. We assume that the measurement has negligible technical noise, so

(∆O)

2 ≡ hO2iζ − hOi2

ζ (6.12)

Substituting Eq. (6.12) into Eq. (6.11) gives the quantum-limited uncertainty for an estimate of

ζ by a single measurement.

Ramsey spectroscopy and Mach-Zehnder interferometry both detect a relative phase φR

by measuring the observable Jz on the family of states

|Ψ(φR)i = O

N

i=1

(cos φR |↓i + sin φR |↑i) (6.13)

produced by the final π/2 pulse (or beamsplitter). In both cases, the finite number N of particles

(atoms or photons) involved in the measurement leads to the shot-noise limit for measurements

of φ [96]

δφSQL = ∆Jz

|∂hJzi/∂φ|

= 1

√

N (6.14)
115

zˆ

uˆ

J

*

ζ

J'

*

Figure 6.3: Estimating rotation angles on the Bloch sphere. Rotation of the coordinate system

is equivalent to rotation of the atomic spin vector.
116

The limit (6.14) readily generalizes to estimation of an arbitrary rotation angle ζ by any

family of states |Ψ(ζ)i that satisfy

|Ψ(ζ)i = O

N

i=1

(ai(ζ)|↓i + bi(ζ)|↑i) (6.15)

in the spin representation. The separability condition (6.15) forbids nonclassical correlations between individual atoms in a spectroscopic sample or between individual photons passing through

an interferometer. Since only classical correlations are present in standard spectroscopy or standard interferometry, Eq. (6.14) is usually called the “standard quantum limit” (SQL). This

fundamental noise source limits the stability of some atomic clocks [96, 97].

The separable states that achieve δζ = 1/

√

N for a given axis ˆu are the coherent spin

states (CSS) [98], which are obtained from |↓iN by applying a rotation such that h

*

JiCSS lies in

the plane perpendicular to ˆu. After the rotation by ζ, we measure the observable J⊥ ≡ *

J · nˆ⊥,

where ˆn⊥ is perpendicular to both ˆu and the final CSS spin vector h

*

Ji. Then Eq. (6.11) gives

the SQL for general rotation angle estimation

δζSQL = ∆J⊥

|h

*

Ji| = 1

√

N (6.16)

6.2.2 Interferometry with Entangled States

States that exhibit quantum correlations allow estimates of ζ with smaller uncertainty

than the SQL [94, 99, 100, 101, 102, 103, 104, 67, 9, 105, 106, 107, 108, 109]. Such sub-shotnoise (SSN) interferometry was first demonstrated using squeezed light [100, 101] and later,

using pairs of entangled photons [105]. Recently Orzel et al. [108] have claimed reduction of

the atom number variance in a collection of Bose-Einstein condensates, a step toward realizing

SSN interferometry with matter waves. The analog of SSN interferometry in atomic spin sys-
117

tems uses entangled spin states to improve rotation angle estimation [109]. In particular, using

SSN interferometry to measure φ can lead to improved frequency standards [103, 104, 67, 109].

Though two experiments using atomic spins [106, 107] have reduced the spin noise ∆J⊥ below

the value expected for a CSS, these experiments did not apply the squeezing to interferometry.

Here we use entangled states of two ions to investigate two different schemes for rotationangle estimation with precision better than that allowed by the SQL [109]. In each case we can

estimate ζ better than allowed by the SQL, showing that SSN interferometry is indeed possible.

We quantify the improvement beyond the SQL using the figure of merit [103, 104]

ξR = δζ

δζSQL

= δζ√N (6.17)

A number of other measures of spin correlations are extant in the literature [110, 111, 102] and

have led to a great deal of confusion in interpreting experimental results [106, 107]. In particular,

the definition of [110, 111] gives different correlation values for different CSSs, even though the

SQL is the same for all CSSs! The definition (6.17), on the other hand, provides a natural figure

of merit for SSN interferometry schemes.

6.2.3 Squeezing Scheme

The “squeezing” scheme for SSN interferometry measures the same observable J⊥ as for

CSS interferometry, but uses a spin-squeezed state (SSS) [102] as input to the interferometer.

Fig. 6.4 compares the use of a CSS and an SSS in estimating θ. Both states satisfy the spin

uncertainty relation [94, 102, 7]

∆Jx ∆Jz ≥ |hJyi|/2 (6.18)
118

θ θ

J

*

(a) (b)

J

*

Figure 6.4: Using a CSS (a) and an SSS (b) to estimate the rotation angle θ. The squeezing of

the uncertainty ellipse in (b) permits improved determination of θ.

but while the CSS has ∆Jz = ∆Jy, the SSS has ∆Jz < ∆Jy, allowing improved determination

of the rotation angle θ about ˆy. The SSS takes its name from the squeezing of the “uncertainty

ellipse” defined by the variances of the spin components, which is reminiscent of the squeezing

of an optical field [9].

Our experiment used a two-ion SSS of form

cos α|↓↓i − isin α|↑↑i (6.19)

where α is the “squeezing angle.” We generated the SSS from the initial state |↓↓i by applying the

entangling evolution Eq. (5.6) with k(t)=2π/r, where we write r = π/α. To avoid entangling

the spin with the motion, we required h(t) = 0, leading to the conditions

∆M = ηkΩsbprj t = 2π

ηkΩsb

rj

r (6.20)

with j an integer. These conditions generalize Eq. (6.3) to arbitrary values of r. The usual

entangling gate just has r = 4, j = 1, corresponding to p = 1 in Eq. (6.3). We realized SSS

according to Eq. (6.19) with r = 10, 6, and 5. In each case we found the appropriate detuning
119

∆M by taking flopping curves while varying the endcap voltage. For r = 10 we used a detuning

that yielded a flopping curve with five fast oscillations per slow oscillation and set the pulse

duration equal to the period of the fast oscillations (j = 1). For r = 5 we used the same detuning, but doubled the pulse duration to give j = 2. For r = 6 we used a detuning with three

fast oscillations per slow oscillation and set j = 1. In each case, the detunings correspond to

noninteger p in Eq. (6.3), so these detunings do not allow operation of the usual entangling gate

(Eq. 5.1). However, they are perfectly adequate for generating SSS of form (6.19).

After generating an SSS with the desired value of r, we applied a rotation Ucar(θ ≈

π/2, φ, φ). We then shelved the |↑i state to the 2S1/2 |F = 1, mF = 1i state as described in

Section 6.1 and counted the number of photons detected over 1 ms of illumination with BD. We

performed 10 000 shots of this experiment at each value of φ, recording the number of photons

detected for each shot individually. Building up a histogram from these 10 000 shots allowed us

to set discriminators for that value of φ corresponding to the minimum detection error, as shown

in Fig. 6.1. The detection errors are then the same as discussed in Section 6.1. Comparing the

number of photons counted in a given shot to the discriminators determined the value of Jz for

that shot.

The orientation of the uncertainty ellipse after the final rotation depends on φ, as shown

in Fig. 6.5. Measuring the variance ∆Jz(φ) as a function of φ therefore maps out the uncertainty ellipse, where the minimum value of ∆Jz(φ) corresponds to the variance ∆Jsqz along

the squeezed direction. We can also regard Jz as measuring the rotation angle θ of the final

rotation, as shown in Fig. 6.4. In the absence of technical noise, our procedure determines θ to

a precision
120

J

*

x

ˆ

y

ˆ

sqz ∆J

φ

Figure 6.5: Mapping out the uncertainty ellipse of an SSS. The minimum value of ∆Jz(φ)

corresponds to the variance ∆Jsqz along the squeezed direction.

δθ(φ) = ∆Jz(φ)

|h

*

Ji| (6.21)

ξR(φ) =

√1 − sin 2α sin 2φ

√2 |cos 2α| (6.22)

While δθ attains its theoretical minimum value of 1/2 for α = π/4, we also find |h

*

Ji| = 0.

Any technical noise in the measurement of Jz then induces large errors in estimating θ, so that

we cannot achieve the optimum δθ in practice. We demonstrated SSN interferometry with the

α = π/4 state |↓↓i + |↑↑i using the “parity” scheme described below.

Fig. 6.6 shows measurements of δθ(φ) for the squeezing scheme with r = 10. The data

follow the theoretical prediction of Eq. (6.22) fairly well. Though the measured minimum

uncertainty δθmin = 0.65 ± 0.01 is considerably larger than the ideal minimum uncertainty

δθideal = 0.56 for this value of r, it is still considerably smaller than the SQL uncertainty

δθSQL = 1/

√2. We have ξR = 0.92 ± 0.01, clearly demonstrating SSN measurement of θ.
121

δθ [rad]

1.1

1.0

0.9

0.8

0.7

0.6

1.6

1.4

1.2

1.0

0.8

SQL

φ [rad]

-π/4 0 π/4 π/2 3π/4

ξ

R

Figure 6.6: Measurements of δθ(φ) as a function of the rotation phase φ for the squeezing scheme

with r = 10. The data (dots) follow the ideal error (dashed curve) fairly well. Error bars for

the data are smaller than the marker size.

We measured δθ(φ) as a function of φ for all three cases (r = 10, 6, and 5). Table 6.2

compares the measured minimum values of δθ to the prediction of Eq. (6.22). The figure of

merit ξR decreases with increasing squeezing angle α (Eq. 6.19). We believe that the loss of

precision arises from imperfect preparation of the SSS.

6.2.4 Parity Scheme

The “parity” scheme uses the maximally entangled state |↓iN +|↑iN encountered in Section (5) to estimate the angle φ. In this case |h

*

Ji| = 0, so we cannot define J⊥ and the CSS

method breaks down. Instead we perform a π/2 rotation at phase φ and measure the expectation

value of the parity operator Π = S1,zS2,z, just as in Section 5.2. Ideally this scheme attains the

Heisenberg limit δφ = 1/N set by the uncertainty relation (6.18). For large N this uncertainty is

a vast improvement over the SQL δφ = 1/

√

N, allowing corresponding improvements in atomic

clocks [103, 104, 67]. Our implementation of this scheme with two ions shows δφR < 1/

√2,

a clear proof of principle for this method. A variant of this scheme using pairs of entangled
122

Table 6.2: Figure of merit ξR for SSN estimation of θ as a function of squeezing angle α. The

loss of precision probably arises from imperfect preparation of the SSS.

α Measured ξR Ideal ξR

π/10 = 18◦ 0.92 0.79

π/6 = 30◦ 0.96 0.73

π/5 = 36◦ 1.40 0.72

photons previously demonstrated SSN estimation of θ [105].

The parity scheme is closely related to the dual Fock-state scheme considered in optical

and matter-wave interferometry [112, 113, 114, 115, 105, 116]. In the spin representation, the

state propagating through the interferometer is again the state |↓iN + |↑iN , but the observable

is the variance V = J2

z − hJzi2 rather than Jz. For two particles, V = (1 + Π)/2 − hJzi2, so the

dual Fock-state scheme and the parity scheme are equivalent. Our data, like that of [105], can

thus be seen as testing both schemes.

We measured the number of photon counts at each value of φ over 10 000 shots and

extracted Jz for each shot, just as described for the squeezing method. Fig. 6.7(a) shows the

observed mean hΠi and variance (∆Π)2 of the parity as a function of φ. For comparison,

Fig. 6.7(b) shows the expected mean and variance for a perfect input state |↓↓i + |↑↑i and

ideal rotation and detection. The imperfect preparation of the entangled state significantly increases the variance for |hΠi| ∼ 1, but is clearly dominated by the quantum noise (Eq. 6.12) for

|hΠi| ∼ 0. We can directly derive the uncertainty δφ from the data using Eq. (6.11). Fig. 6.8

shows these derived values of δφ as a function of φ. While the technical noise increases δφ well

above the SQL for |hΠi| ∼ 1, δφ drops significantly below the SQL for |hΠi| ∼ 0. The minimum

value of δφ is 0.59 ± 0.01 < 1/

√2, giving a squeezing factor ξR = 0.83 ± 0.01.
123

6.2.5 Sub-Shot-Noise Ramsey Spectroscopy

The parity method clearly improves our estimate of φ, so it should also allow improved

determination of ωR in atomic clocks [103, 104, 67]. To demonstrate this result directly, we

performed Ramsey spectroscopy with a nonclassical input state. Table 6.9 gives the sequence of

operations used for this experiment. Here the state after operation 4 is

|Ψiout = sin(2TR∆ω)

h

e−i(α+φfluct)|↓↓i − ei(α+φfluct)|↑↑ii− cos(2TR∆ω) [|↓↑i + |↑↓i] (6.23)

so that hΠi = sin2(2TR∆ωR). Operations 0 through 2 prepare the desired input state |↓↑i+|↑↓i,

while operations 3 through 5 perform Ramsey spectroscopy on the input state. We scanned the

accumulated phase TRωR by detuning operations 3 and 5 from the carrier frequency by about

10 kHz and scanning the free evolution time TR. Our experiment implements the Heisenberglimited Ramsey scheme of [67], for which the state between the Ramsey π/2 pulses must be the

state |↓iN + eiβ|↑iN for some phase β.

The sensitivity of Ramsey spectroscopy scales as the time TR between the Ramsey π/2

pulses, so atomic clocks usually use the longest feasible TR [88, 89]. In our case, heating affects

carrier pulses performed with the B1, R2 beams after a few tens of μs (Sec. 7.4), effectively limiting TR. We therefore used the copropagating B1, R1 beams for operations 3 and 4. However, we

could only perform the entangling gate using the B1, R2 beams (Sec. 5.2). The phase difference

φ1−φ2 for any operation using B1, R1 is zero, since *

∆k = 0 for these beams. The method of [67]

requires the state after operation 3 to be |↓↓i + eiβ|↑↑i, so the state before operation 3 must be

|↓↑i+|↑↓i, forcing φ1−φ2 = 0 for operation 2 as well. The ions therefore had to be separated by
124

-1.0

-0.5

0.0

0.5

1.0

-1.0

-0.5

0.0

0.5

1.0 (b)

−π/2 −π/4 0 π/4

φ [rad]

Ä3Ô

(a)

Figure 6.7: Mean and variance of the parity operator Π in the parity scheme for SSN interferometry. (a) Experimental data, 10 000 shots per point. Points give the mean value of Π as

a function of φ, error bars give the variance. The curve is the best fit to the mean of Π. (b)

Theoretical values for perfect state preparation and measurement. The curve is the ideal value

of the mean as a function of φ, and error bars give the ideal variance.

Table 6.3: The sequence of operations implementing SSN Ramsey spectroscopy. Operations 0

through 2 prepare the nonclassical input state, while operations 3 through 5 perform Ramsey

spectroscopy on the input state. We define the zero of RF phase such that φ1 +φ2 = 0 for pulse

1.

Number Operation Beams Used φ1 + φ2 φ1 − φ2 State After Operation

0 initialize – – – |↓↓i

1 Uent B1, R2 0 φλ |↓↓i − i|↑↑i

2 Ucar(θ = π/2) B1, R2 −π/4 0 |↓↑i + |↑↓i

3 Ucar(θ = π/2) B1, R1 φfluct 0 |↓↓i + e2iφfluct |↑↑i

4 wait – – – |↓↓i + e2i(φfluct−TR∆ω)

|↑↑i

5 Ucar(θ = π/2) B1, R1 φfluct 0 |Ψouti

6 detect – – – –
125

1.0

0.9

0.8

0.7

0.6

0.5

δφ [rad]

φ [rad]

-π/2

SQL

Heisenberg Limit

-π/4 0 π/4

Figure 6.8: Measurement uncertainty δφ derived from the data of Fig. 6.7(a). For certain values

of φ the uncertainty drops below the SQL, demonstrating SSN interferometry.

a distance s0 = 2πn/(

√2k) for operation 2, where n is an integer. Because of the fixed resonant

frequency of the entangling EOM (Sec. 4.4), we could not satisfy this condition for operation 1.

We therefore changed the endcap voltage between operations 1 and 2 to produce φ1 − φ2 = 0

for operation 2. Minimizing P1 for the state after operation 3 ensured this condition.

The relative optical phase between R1 and R2 was not stabilized for this experiment,

so that even if the Ramsey pulses were not detuned from the logic pulses, random changes in

the relative path length produced a fluctuating phase φfluct, as shown in Table 6.9. We had

previously characterized this phase using simple Ramsey spectroscopy on a single ion. When

we applied one π/2 pulse with the B1, R1 beams and the other with the B1, R2 beams, we

observed the usual Ramsey fringes (Eq. 6.6) for measurement times of a few tenths of a second.

At longer measurement times the phase of the fringes shifted randomly and suddenly, indicating

an abrupt change of the relative path length. The cause of the fluctuations remains unknown.

Since we usually needed 1 s or more to achieve the high signal-to-noise ratio characteristic of our

experiment, we could not perform experiments that required phase coherence between B1, R2
126

operations and B1, R1 operations. Surprisingly, the experiment described in Table 6.9 does not

require coherence between the operations using R2 and those using R1. While the final state

|Ψouti does indeed vary with φfluct, the populations depend only on the accumulated phase

TRωR because of the particular form of the input state |↓↑i + |↑↓i. Other input states generally

give rise to a signal dependent on φfluct, limiting measurement time to less than 1 s.

We collected parity data at several values of TR in the way described previously. Using

Eq. (6.11), we obtained the measurement uncertainty TRδωR as a function of TR. The data

are shown in Fig. 6.9. Since scanning TR is equivalent to scanning the phase φ in the previous

parity experiment, the data of Fig. 6.9 are similar to that of Fig. 6.8. Again, the technical noise

increases the uncertainty far above the SQL for values of TR corresponding to |hΠi| ∼ 1, but

the uncertainty drops well below the SQL for TR such that |hΠi| ∼ 0. The maximum gain in

precision is a factor of 1.14 ± 0.01, corresponding to ξR = 0.88 ± 0.01. Here we are comparing

the precision obtained with our imperfect state preparation and measurement to the precision

obtained for an experiment on a CSS of two ions with ideal preparation and measurement.

The SSN Ramsey scheme [67] for two ions is thus more precise than any concievable frequency

measurement on two unentangled ions. The value of ξR measured here is consistent with that

measured for the previous parity experiment, since the Ramsey π/2 pulses add some extra noise

to the SSN Ramsey experiment.
127

1.7

1.5

1.3

1.1

0.9

0.7

0.5

6 8 10 12 14 16 18

Heisenberg Limit

T

R ∆ω

R [rad]

SQL

TR [μs]

Figure 6.9: Frequency uncertainty TRδω as a function of TR. The uncertainty drops below the

SQL for TR such that |hΠi| ∼ 0, demonstrating SSN Ramsey interferometry.
Chapter 7

Decoherence of the Quantum Register

7.1 Dephasing of Internal States

A qubit in one of the basis states |↓i, |↑i is practically unaffected by its environment in

our experiments. The radiative lifetime of our qubit states is long (À 1012 s), so the qubit levels

|↓i, |↑i do not decay. Stray radiation coupling the basis states must have a frequency near the

hyperfine splitting of 1.25 GHz. Typical ambient noise power at this frequency is so small as

to be negligible. Quasistatic rotation of the quantization field is also negligible, since ambient

fluctuating field strengths are much less than the quantization field strength.

Notwithstanding these facts, a qubit in a superposition state |ψ(φ)i = |↓i + eiφ|↑i is

readily susceptible to changes in its phase φ. In Section 6.2.1, we showed how to exploit this

effect to stabilize the frequency of an LO to an atomic transition, realizing an atomic clock.

Ideally, the LO and atomic phases are synchronized at all times, so that in a frame rotating at

the LO frequency the state |ψ(φ)i remains the same as time passes. If the relative phase drifts

by an unknown amount φd, though, we might presume that we have the state |ψ(φ)i when in

fact we have |ψ(φ+φd)i. Now suppose that the state |ψ(φ)i is the state after the first π/2 pulse

of a Ramsey experiment and φd is characterized by a Gaussian distribution of width σφ over the

free evolution time TR. The second π/2 pulse rotates the density matrix ρ associated with the
129

state after the first π/2 pulse and the subsequent measurement of Sz gives the coherence ρ↓,↑

as

hSziφ = |ρ↓,↑|sin φ (7.1)

= 1

2

Z

sin (φ + φd)

Ã

dφd

√2πσφ

exp "− φ2

d

2σ2

φ

#! (7.2)

= 1

2 exp[−2σ2

φ] sin φ (7.3)

so that the coherence is reduced by a factor exp[−2σ2

φ] from its ideal value. This type of decoherence, known as dephasing, occurs whenever an unknown phase is added to a superposition

state.

To evaluate σφ, we must consider the physical process inducing dephasing. If the process

acts for a time τ and has a spectral density Sφ(ω) for the phase fluctuation φd, the WienerKhintchine theorem yields [117]

σ2

φ = 2 Z ∞

0

Sφ(ω)

sin ωτ

ωτ

dω (7.4)

Frequently we can approximate S(ω) by white noise with a high-frequency cutoff ωlp, so that

S(ω = 1/τ ) σφ |ρ↓,↑|

τ . 1/ωlp S0 ∝ τ ∝ exp−(τ/Tlp)2

τ & 1/ωlp 0 ∝ √τ ∝ exp−τ/T2

(7.5)

In the white-noise case, ωlp → ∞ and we have an exponential decay for all times, with decay

rate set by the dephasing time T2. In the other limit we have Gaussian decay with dephasing

time Tlp.
130

We can describe dephasing as implementing transformations |↓ij → |↓ij , |↑ij → eiφd,j |↑ij ,

where j indexes the qubits and the φd,j are (possibly correlated) random variables. From this

standpoint it is easy to characterize dephasing of a general density matrix of N qubits. Using

the product basis {|ki} over the single-qubit states {|↓i, |↑i}, we can write the dephased density

matrix ρ0 in terms of the original density matrix ρ as

ρ0

k,k0 = hk0

|ρ|ki (7.6)

→ exp



i

X

N

j=1

φd,j (hSz,j ik − hSz,j i

0

k)



ρk,k0 (7.7)

∼ exp[iφd(hJzik − hJzik0 ]ρk,k0 if φd,j ≈ φd for all j (7.8)

For white noise, we find that the dephasing time for a general state of N ions is no faster than

T2/N (and can be much slower: see Chap. 8). This result contrasts strongly with similar results

on dephasing of harmonic oscillators. In the latter case, the dephasing time between Fock states

|ni and |n0

i scales as |n − n0|

2 [118, 119, 120]. Apparently, large superposition states made of

qubits are much less susceptible to decoherence than large superpositions made with a harmonic

oscillator.

In our experiment, dephasing during free evolution of the register mostly arose from

fluctuating magnetic fields. To first order, these fields did not change the direction of the quantization field, but they did change its magnitude: the Zeeman effect shifted ω↓↑ by about 21

GHz/T. The Raman beatnote was derived from a high-stability RF source and has negligible

phase noise.1 We actively stabilized the magnetic field using a pair of coils perpendicular to,

and centered on, the path of BD. The coils were placed 10 cm on each side of the trap. Fluctuating magnetic fields induced currents on one coil, which were picked up, amplified, and driven

through the other coil. This feed-forward arrangement reduced the magnetic field fluctuation

1 In contrast to the atomic-clock situation, here the LO was stable and the atomic transition fluctuated.
131

by about a factor of two. The 1/e time for decay of the coherence was about 150 μs with this

arrangement.

Almost all the magnetic field noise comes from AC power line noise, so its spectrum is

concentrated at 60 Hz and its harmonics. The analysis leading to Eq. 7.1 is therefore invalid,

since φd is correlated from shot to shot. If we assume that the fluctuations are sinusoidal with

a period much longer than the duration of a single shot, we find φd = βTR sin φn over one shot,

where β measures the amplitude of the frequency fluctuations and φn is the fluctuation phase

for that shot. The analog of Eq. (7.1) is then

hSziφ =

Z 2π

0

sin (φ + β sin φn)

μ dφn √2π

¶

(7.9)

= 1

2

J0(2βTR) sin φ (7.10)

where J0 is a Bessel function. The coherence (Eq. 7.9) exhibits collapses and revivals as βTR

increases. For fixed β, the Bessel function asymptotically decays as T −1/2

R , rather than decaying

exponentially. However, for short times the coherence decays as a Gaussian, as in Eq. (7.5).

In real life, the magnetic field noise is neither monochromatic nor Gaussian, but the coherence

does in fact exhibit collapses and revivals.

We performed a Ramsey experiment with a single ion on the transition between the

2S1/2 |F = 2, mF = 0i and |F = 1, mF = 0i hyperfine states. Here the transition frequency

has only a quadratic Zeeman shift at zero field. At our field strength of 165 μT we estimate

a Zeeman shift of 1.2 kHz/μT, as compared to the 21 kHz/μT Zeeman shift of the |↓i ↔ |↑i

transition. After initializing the ion in |↓i, shelving pulses using B1 and R1 transferred the ion to

|↑i, then to |F = 2, mF = 0i. From Eq. (3.37), driving the |F = 2, mF = 0i↔|F = 1, mF = 0i
132

transition requires circular polarization of both Raman beams in the atomic basis, so we used

the B2, R2 beams to perform the Ramsey π/2 pulses. After the π/2 pulses, we transferred the

population in |F = 2, mF = 0i back to |↓i and measured the fluorescence. These measurements

showed that the coherence lasted for a 1/e time of about 1 ms, rather than the 150 μs found for

the |↓i ↔ |↓i transition.

7.2 Single-Qubit Gate Errors

The results on Raman transitions presented in Chapter 3 assume that the ion state evolves

coherently under the Raman Hamiltonian (3.42). In reality, many mechanisms create decoherence during Raman transitions, causing errors in our quantum logic gates. The fundamental

source of error is spontaneous emission from the small state amplitude in the P manifold, which

has an inevitable lower bound depending on the atomic structure. Since the logic Hamiltonian

depends on the ions’ motional state, any thermal motion of the ions can lead to gate errors (see

Sec. 7.4 below.) In addition, technical noise in the Raman beams can create gate errors, though

in principle such errors can be very small.

7.2.1 Characterization of Gate Errors

In Section 7.1, we saw that dephasing is the only decoherence mechanism affecting one of

our qubits under ambient conditions. However, during a logic gate, processes like spontaneous

Raman scattering can change the populations of internal states as well as their coherences,

causing decay of the populations toward some equilibrium distribution. To quantify the relative

importance of dephasing and population decay, we consider the decoherence of a single qubit

arising from a rotation applied to that qubit. From Eq. (7.1), a phase error φd transforms
133

ρk,k0 → exp[−2σ2

φ(hSzik − hSzik0 )]ρk,k0 (7.11)

where Sz acts on the rotated ion. To model exponential decay of the populations by a factor γr

toward an equilibrium distribution Peq

k over the product basis states, we make the transformation

ρk,k → e−γr ρk,k + (1 − e−γr )ρeq

k,k (7.12)

We can measure the error induced by a decoherence process using the fidelity F = Tr[ρρ0

]

as

2 = 1 − F = 1 − Tr[ρρ0

] (7.13)

where ρ → ρ0 under the decoherence. We take ρ to represent an intermediate state in a large

quantum computation. In such a state, the populations are about equally distributed over the

possible measurement basis states, so we take ρk,k ∼ 1/2N for any product basis state |ki.

Assuming maximum coherence of the register state, we find ρk,k0 ∼ 1/2N for any |ki, |k0

i as

well. Then the error due to dephasing is

2 = X

k,k0

ρk,k0 (1 − exp[−2σφ(hSzik − hSzi

0

k)])ρk0,k (7.14)

∼ 1 − 1

2

e−σ2

φ Tr£

ρ2¤

∼ 1 − 1

2

e−σ2

φ (7.15)

since hSzik − hSzi0

k = 1 for about half the possible choices of |ki, |k0

i. On the other hand, the

error due to population decay is

2 = X

k

ρk,k(1 − e−γr )(ρk,k + Peq

k ) (7.16)

∼ 1 − e−γr

2N

X

k

μ 1

2N + Peq

k

¶

(7.17)

. 1 − e−γr

2N (7.18)
134

Any process that causes population decay also causes dephasing, so we see that as the quantum

register becomes large (N → ∞) we can ignore relaxation errors.

Noise processes in a single-qubit gate introduce fluctuations in the phase or the polar angle of the rotation operator Ucar (Eq. 5.26). Solving the optical Bloch equations with dephasing

[36] shows that the presence of noise causes decay of the inversion hSzi(t) and coherence hS+i.

To observe the effects of noise in the experiment, we typically performed Rabi flopping on the

carrier resonance of a single ion and observed the decay envelope of the Rabi oscillations. We

could then characterize the noise by fitting the observed decay to the model based on the Bloch

equations.

For perfect rotations, the inversion and coherence oscillate at the Rabi frequency ΩR, so

the duration t = θ/ΩR of the Raman pulse sets the timescale of interest for the noise. Assuming a

flat noise spectrum with a high-frequency cutoff ωlp, we find two limiting cases, as in Section 7.1:

the white noise regime ωlp & ΩR and the slow noise regime ωlp . ΩR. White noise in θ and

φ have roughly the same effect, so we model both by including white-noise dephasing in the

optical Bloch equations. Starting from |↓i and applying a carrier pulse of duration t, we find in

this model

hS+i = e−t/T2 sin ΩRt (7.19)

hSzi = −1

2

e−t/T2 cos ΩRt (7.20)

so that dephasing during a logic gate causes decay of the ion populations (∝ hSzi) as well as

decay of the coherence (∝ hS+i).
135

In the experiment, the effects of slow noise in θ dominate the slow noise in φ. To model

this noise, we average the noiseless solution to the Bloch equations over a Gaussian distribution

in θ centered about the desired value θ. We find

hS+i = e−(t/Tlp)2

sin ΩRt (7.21)

hSzi = −1

2

e−(t/Tlp)2

cos ΩRt (7.22)

where we have exploited the similarity to Eq. (7.5) to define an effective dephasing time Tlp.

In both models, the inversion and the coherence decay in the same way, so we can use

flopping curves on one ion to extract the effective dephasing behavior due to any source of gate

error. Hence we can predict the effect of gate errors on a large register from the behavior of

a single ion. Since the decay of coherence under gate errors is similar to that under ambient

noise (Sec. 7.1), we can describe both gate errors and ambient noise in terms of the associated

white-noise dephasing time T2 and slow-noise dephasing time Tlp. Adding the contributions

from several (small) sources of noise gives the overall dephasing times as

1

T2

= X

j

1

T2,j

1

T2

lp

= X

j

1

T2

lp,j

(7.23)

where j labels the individual noise sources.

7.2.2 Spontaneous Emission

Although the Raman laser detuning is always much larger than the linewidth Γ = 2π×19.4

MHz of the S ↔ P transition, a small part of the atomic population remains in the P state
136

throughout a Raman transition. This population can spontaneously decay, emitting a photon

with a random phase and adding a correlated random phase to the hyperfine coherence ρ12 in

the process. Since we cannot recover the emitted photon phase, a spontaneous scattering event

essentially destroys the hyperfine coherence. Assuming that the Raman beam intensities are

constant across the ion crystal, the probability that a spontaneous scattering event will occur

at a particular ion is [9]

Pse = γseT (7.24)

γse ≡ X

j=1,2

X

k={1/2,3/2}

Ω2

jΓ

R(k)

Stark,j

∆2

k

(7.25)

for a Raman pulse of length T. If no emission event occurs, the density matrix after the logic

gate is the ideal density matrix ρI . However, if an event occurs at ion i, all coherences involving

ion i vanish. A little algebra shows that spontaneous scattering from the Raman beams effectively induces white-noise dephasing with T2 = 1/γse.

The probability of dephasing during a 2π carrier pulse is Pse,2π(∆) = 2πγse/|Ω12|, where

Ω12 is given by Eq. (3.40.) This quantity is a convenient measure of the spontaneous emission

error per gate. We have

Pse,2π = 2πΓ

RRamanΩ1Ω2

 ̄

 ̄

 ̄

 ̄

∆

μ

1 − ∆

ωFS ¶ ̄

 ̄

 ̄

 ̄

X

j=1,2

X

k={1/2,3/2}

Ω2

j

R(k)

Stark,j

∆2

k

(7.26)

For the beams B1, R2 used to drive logic gates on the |↓i ↔ |↑i transition, we have ˆ2B1 = π,

2ˆR2 = σ+ + σ−. Then writing |ΩB1| = r|ΩR2|, we find

Pse,2π(∆, r) ∝

 ̄

 ̄

 ̄

 ̄

∆

μ

1 − ∆

ωFS ¶ ̄

 ̄

 ̄

 ̄

μ2r2 + 5

3r

1

∆2 +

4r2 + 8

3r

1

(ωFS − ∆)2

¶

(7.27)

which exhibits a broad minimum near 313.160 nm and r = 1.5 with Pse ≈ 6 × 10−3. For the

experimental Raman wavelength 313.341 nm, and allowing for day-to-day variations of r from
137

0.5 to 2, the value of Pse,2π is larger than the minimum by no more than a factor of 2.

The relatively large gate error Pse,2π limits us to a few tens of quantum logic operations.

We cannot reduce this source of error through technical improvements. However, from Eq. (7.27),

we see that Pse,2π scales as 1/ωFS. While the 9Be+ ion has a relatively small ωFS ≈ 2π × 197

GHz, heavier ions can have much larger values of ωFS, e.g., tens of THz for Cd+. For this reason,

such ions may be more promising qubits than 9Be+ for large-scale quantum computation.

Finally, we note that an ion can change hyperfine state on spontaneously scattering a

photon from the Raman beams. Both Rayleigh and Raman scattering can occur. The transition

rate from the initial state |ii to the final state |fi, with i 6= f, is just

T(i, f) = X

j=1,2

X

k={1/2,3/2}

X

Fk

X

2ˆa

Ω2

jΓS(i, Fk; ˆ2a)S(f, Fk; ˆ2

0

a)(ˆ2

∗

a · 2ˆj )

∆2

k

(7.28)

from which we can deduce the population decay under spontaneous emission. This process can

remove an ion from the computational basis {|↓i, |↑i} entirely. The relaxation is a useful diagnostic, as we could deduce the single-photon Rabi frequency of a Raman beam by applying that

beam alone and observing the decay of the ion populations. During quantum logic operations,

however, the population decay from this effect was much smaller than the population decay from

dephasing.
138

7.2.3 Ion Motion

The Rabi frequency of logic transitions (3.22) is a function of the Fock state of the ion

motion. Assuming that the motional state does not change much over the course of a Raman

pulse, the evolution under the logic Hamiltonian is no longer sinusoidal but becomes a sum of

Rabi flopping curves of form (3.21) weighted by the populations of the motional Fock states.

This indirect effect of the ion motion on the carrier transition is analogous to the Debye-Waller

(DW) effect encountered in condensed-matter physics [121].

Here we will treat only the decoherence of the carrier transition. We will discuss decoherence of the two-ion gate in Sec. 5. For one ion, the Rabi frequency (3.22) reduces to

Ωn ≡ Ωn,n = ΩRe−η2/2(1 − nη2). If the ions are not perfectly cold, n fluctuates about its

average value n, so that θ = Ωnt effectively fluctuates from shot to shot, as in the slow-noise

model. Expanding about the average value Ω=ΩRe−η2/2(1 − nη2) gives

Tlp,DW =

"

X

n

Pn(Ωn − Ω)2

#−1/2

(7.29)

For multiple ions, the gate error from the DW effect depends on the fluctuations in all N

axial modes. Writing Tlp,j the error due to the jth mode, we find the total decay time Tlp,DW

from Eq. (7.23). If the ions are in the Lamb-Dicke limit and each of the N axial modes is in a

thermal state of average excitation np, we find [2]

ΩRTlp,DW =

"

X

N

p=1

3

v(p)

j ηp

 ́4

np(np + 1)#−1/2(7.30)

η2

p ≡

√2k

(2NmBeξpωz)1/2 (7.31)

where we define ηp the Lamb-Dicke parameter of the pth mode of N ions and the quantities

v(p)

j , ξp are defined by Eq. (2.26). For a given set of motional-state parameters, we see that
139

ΩRTlp,DW is constant, so the total gate error is proportional to the rotation angle θ.

7.2.4 Technical Noise

Intensity fluctuations of the Raman light during a logic gate also lead to uncontrolled

fluctuations of the Rabi frequency with time. If the bandwidth of the intensity noise is much

less than the inverse of the Raman pulse time, as is the case in our experiments, we can use

the slow-noise model of the gate error. Since ΩR ∝ √I1I2, where Ij is the intensity of the jth

Raman beam, we find

ΩRTlp = 2

(σ1 + σ2) (7.32)

where σj is the relative intensity noise (RIN) of the jth beam. Again we see that the total gate

error is proportional to θ. Contributions to σj over times longer than the averaging time are

neglected here, so the “decoherence” here depends on the length of the quantum computation

being attempted. In other words, if you want to run a computation ten seconds long, the Raman

beam intensity only needs to be stable for ten seconds. Eq. (7.32) is only an approximation to

the gate error if the noise bandwidth is smaller than the average time between gates, as the gate

errors can be correlated in this case.

Though the intensity noise cannot be reduced below the shot-noise limit, the resulting

lower limit on gate error can be very small [2]. Typically we achieved about 10 Rabi flops for

a gate error < 50% over an averaging time of ∼ 30 s using intensity-stabilized Raman beams

(Secs. 4.3.1 and 4.3.2). Assuming σ1 ≈ σ2, Eq. (7.32) gives σ1 ≈ 10−2 over the frequency range

between 1/30 Hz and ΩR ∼ 100 kHz. Each Raman beam has a power ∼ 1 mW, so the shot-noise
140

RIN over this frequency band is ∼ 10−5. It is technically feasible to approach the shot-noise

RIN, so this source of gate error can be greatly reduced.

Fluctuations in the relative optical phase of the Raman beams can induce fluctuations

in the ion phases. The difference frequency between the Raman beams is generated by highstability RF synthesizers, so phase noise from this source is negligible. The motion of the ion

relative to the trap effectively causes fluctuations in the phase, as described in the last section.

Phase errors can also arise from motion of the trap relative to the laser beams, which produces

the same effect as phase noise of the RF synthesizer. However, we do not need to know the

relative phases between gates performed in different shots of the experiment, so fluctuations

slower than ∼ 1 ms do not affect our results. We damped out vibrations of the trap relative to

the resonator using springs inside the vacuum chamber. Mechanical vibrations of the resonator

should be slower than 1 kHz. In any event, we see little evidence of phase fluctuations unless we

require coherence between logic gates performed with different beam geometries (see Sec. 6.2.5).

However, mechanical vibrations may have an effect on long quantum computations.

7.3 Errors in the Entangling Gate

All the errors described in the previous section for single-qubit rotations also affect the

entangling gate, although to different extents. The T2 from spontaneous emission remains the

same. Since the entangling gate is slower than a π/2 rotation by a factor of 2/ηlogic, the total

gate error due to spontaneous emission increases by a factor of 2/ηlogic. On the other hand, the

error due to laser intensity noise is just proportional to the rotation angle θ ≡ ΩMt = π/2 of

the entangling gate, so the error in the entangling gate from this source is the same as the error

in a π/2 carrier rotation.
141

An evaluation of the effects of ion motion on the entangling gate can be found in [66].

Since the entangling gate acts directly on the logic mode, the motion in this mode affects the gate

differently from the motion in the spectator modes. In thermal equilibrium, we can characterize

the motion by the average excitation nCOM of the center-of-mass mode. The Debye-Waller effect

of the spectator modes gives an upper limit to the error contribution

2DW,ent ≤ π2N(N + 1)

8 η4

COMnCOM(1.2nCOM + 1.4) (7.33)

Heating of the logic mode during gate operation can cause errors in the nonperturbative

regime by destroying the spin/motion entanglement essential to the gate. Assuming a constant

heating rate on the logic mode, a master-equation treatment [66] predicts a contribution to the

gate error

2heat,ent = πN

4ηlogicΩR

dnCOM

dt (7.34)

which is just proportional to the number of quanta gained during a sideband π-pulse time. Our

experiments used a logic mode with essentially zero heating rate (see Sec. 7.4), so this source of

error was negligible.

An additional source of technical noise comes into play for the entangling gate. The

Stark shift of ω↓↑ (Eq. 3.36) depends on the Raman beam power, so intensity noise of the beams

causes a fluctuating Stark shift. These fluctuations are no greater than a few tens of kHz RMS,

so they are small compared to the carrier Rabi frequency ΩR and have no observable effect on

single-qubit rotations. However, the linewidth of the entangling transition is only ∼ 10 kHz, so

these fluctuations have a large effect on the entangling gate.
142

The Stark shift varies with the wavelength of the Raman beams in much the same way

as the spontaneous emission rate (Sec. 7.2.2). The ratio κStark of Stark shift to carrier Rabi

frequency conveniently quantifies the sensitivity of a single-qubit rotation to the Stark shift.

The Stark shift was given in Eq. (3.39). Taking the usual Raman beam polarizations ˆ2B1 = π,

2ˆR2 = σ+ +σ− on the |↓i ↔ |↑i transition, we find from Eq. (3.39) that the Stark shift is exactly

zero. However, Eq. (3.39) fails to account for the small corrections to the detunings ∆k arising

from the hyperfine splitting ωHF. To lowest order in ωH/∆k, we find the Stark shift

δStark = ωHF

Ω2

B1 + Ω2R2

3

· 1

∆2 +

2

(ωFS − ∆)2

 ̧

(7.35)

and writing |ΩB1| = r|ΩR2| gives

κStark = 1 + r2

r

∆

μ

1 − ∆

ωFS ¶ · 1∆2 +

2

(ωFS − ∆)2

 ̧

(7.36)

The sensitivity to Stark shift is minimized for r = 1 with the Raman wavelength of 313.341 nm

used in the experiment. In practice, we needed to carefully adjust the polarization of R2 using

a λ/4 plate to obtain the minimum κStark. The response of the entangling gate to a detuning

is difficult to calculate, so we did not quantify the noise contribution from this mechanism.

However, we did not observe broadening of the entangling transition beyond the 10 kHz Rabi

linewidth for κStark near its minimum, suggesting that the remaining noise did not induce much

gate error.

We also implemented the entangling gate on the field-independent |F = 2, mF = 0i↔|F =

1, mF = 0i transition. Here we used the B2, R1 beams as described in Section 4.3.2. The fidelity

was about the same as that obtained on the |↓i ↔ |↑i transition. The |↓i ↔ |↑i transition is

about a factor of 20 more sensitive to magnetic field fluctuations than the field-independent

transition, as described in Section 7.1. The Stark shift of the field-independent transition is also
143

independent of polarization. Hence the experiments on the field-independent transition ruled

out polarization and magnetic field fluctuations as sources of error in the entangling gate.

We were never able to reduce the error in the two-ion gate below ∼ 0.08. The fundamental error limit due to spontaneous emission is 0.02. The initial temperature of the COM

mode was at least n = 0.50, because of the limit to Raman cooling imposed by the ion heating

(Secs. 3.3.4 and 7.4). Eq. (7.33) then predicts an error of ∼ 0.06 from the Debye-Waller effect.

The error from these two effects accounts for most of the observed error, so we believe that other

technical noise did not contribute much error. The four-ion gate error of 0.43 probably arises

almost completely from the Debye-Waller effect. From Eq. (7.33) we see that the Debye-Waller

error scales as N2 for fixed nCOM. It was difficult to measure the temperature of the four-ion

crystal, but nCOM was certainly higher for four ions than for two. Reducing the heating rate

and thus the initial temperature should enable entanglement of more than four ions.

7.4 Heating of the Ion Motion

In all ion traps described in this thesis, we observed a heating of the COM motional

mode. This heating caused the COM motional state to relax toward a thermal distribution with

average occupation number nCOM over motional Fock states. Since nCOM increased linearly

under the heating, we quantify the heating rate as Γheat = n ̇ COM. To measure nCOM at a given

time, we typically took sideband spectra for a single ion. Defining Prsb (Pbsb) the maximum

population in |↑i when driving the red (blue) sideband, we have [2, 122]

nCOM = Prsb

Pbsb − Prsb

(7.37)

which is valid for any drive duration, even if the ion is not in the Lamb-Dicke limit. To obtain

high signal-to-noise, we typically set the drive duration to maximize Pbsb. We extracted Γheat

by Raman cooling the ion near the ground state and measuring nCOM after a variable delay.
144

4

3

2

1

0

<n>

0.00 0.05 0.10 0.15 0.20

t [ms]

rsb

bsb

rsb

bsb

∆ω

P↓

∆ω

P↓

0

1

0

1

Figure 7.1: A typical heating measurement. The main graph shows nCOM vs. the delay time t.

Insets show the population P↓ remaining in |↓i after the Raman drive. The data are for linear

trap #2 with secular frequency 5 MHz. The data show Γheat = 12 ± 2 quanta/ms.

Figure 7.1 shows a typical example of a heating measurement.

Since ions are electrically charged, the obvious candidate for a source of heating is a

fluctuating electric field. If this field originates at the trap electrodes, we expect the field to be

spatially uniform over the ion crystal, so that the field couples only to the COM mode of the

crystal. Heating measurements with a two-ion crystal [28] confirm that the COM heating rate is

a factor of several hundred larger than the heating rate of the axial stretch mode. We measure

the fluctuating field in terms of a spectral density SE(ω) (in SI units) normalized so that the

RMS field strength is (2π)−1 R ∞

0 SE(ω) dω. For a single ion of mass m, the heating rate can be

found from SE as [123]

Γheat = 104e2

4m~ωCOM

SE(ωCOM) (7.38)

The ion only responds to driving fields near ωCOM. Assuming the electric field originates at

the electrodes and that the fluctuating potential is spatially uniform over an electrode, a typ-
145

1

10

heating rate [quanta/ms]

2 3 4 5 6 7 8 9

10

trap secular frequency, ωm /2π [MHz]

 trap 5

 trap 6

Figure 7.2: Heating measurements for trap #2 (squares) and trap #3 (triangles) at several

values of ωCOM. Γheat approximately scales as ω−2

COM.

ical heating rate Γheat ∼ 1ms−1 corresponds to a fluctuating potential noise power density

∼ 1(nV)2/Hz near ωCOM. Even a very small fluctuation is enough to give rise to the observed

heating.

7.4.1 Heating Measurements and Interpretations

We measured Γheat for a wide variety of traps, as described in [122]. The heating rate

was typically 1−30ms−1 for all traps. These included Paul traps constructed from both Be and

molybdenum, and the three linear traps, which had gold electrodes, so the heating appears to

be largely independent of trap design. Figure 7.2 shows heating data for linear traps #2 and

#3 at several values of ωCOM. From Eq. (7.38) and the approximate ω−2

COM scaling of the data,

we find SE ∼ 1/ωCOM. The difference in scale between the two data sets is probably due to the

increase in wafer spacing from trap #2 to trap #3 (see Sec. 2.2.1).

Figure 7.2 suggests that the heating rate depends on the size of the trap. However, we
146

ring 

electrode

trap 3b trap 3a

fork electrode

Figure 7.3: Schematic of the double Paul trap. The distance between the two rings is 1.7 mm.

could not confidently compare data from different traps because of the many factors that could

not be controlled from trap to trap. To give only one example, it is extremely difficult to investigate the surface chemistry at the electrodes in situ. To investigate the size dependence in a

controlled way, we constructed a “double” ring-and-fork Paul trap with two rings, as shown in

Figure 7.3. By this means we hoped to compare the two traps formed by the two rings under

similar conditions. We refer to the larger (smaller) trap as trap PL (PS). The distance d from the

ion to the nearest electrode surface was 395 μm for PL, 175 μm for PS. Figure 7.4 shows Γheat

as a function of ωCOM for two datasets taken with these traps. Between the two datasets, the

trap was removed from the vacuum apparatus, cleaned as described in [122], and replaced. The

first dataset yields a scaling Γheat ∝ d(3.8±0.6), while the second yields the remarkably strong

scaling Γheat ∝ d(12±2).

The heating rate also tended to increase as a trap aged. The first trap used for quantum

logic experiments in our group was a molybdenum Paul trap (trap P1) with d = 170μm. Figure 7.5 shows heating rate measurements for trap P1 over a period of almost two years. The trap

was under vacuum continuously over this time, and the only known change inside the vacuum

apparatus was the plating of Be onto the electrodes from loading. Here the solid line is meant
147

0.1

1

10

100

heating rate [quanta/ms]

1 2 3 4 5 6 7 8 910

(a) small trap (3a)

 big trap (3b)

0.001

0.01

0.1

1

10

100

heating rate [quanta/ms]

1 2 3 4 5 6 7 8 910

trap secular frequency, ωm /2π [MHz]

(b) small trap (3a)

 big trap (3b)

Figure 7.4: Heating measurements for trap PL (circles) and trap PS (squares). (a) First dataset,

showing a scaling Γheat ∝ d(3.8±0.6). (b) Second dataset, showing a scaling Γheat ∝ d(12±2).
148

solely as a guide to the eye, but it certainly appears as though the heating rate increased by

about a factor of four over this time.

The heating data obtained so far do not clearly indicate the source of heating. However,

we have ruled out a great many possible sources of heating, including injection of noise from the

RF and DC sources used to operate the trap, background gas collisions, and RF heating [2]. One

theoretical explanation, which is absolutely not borne out by experiment, is thermal (Johnson)

noise [124, 125]. Modeling the thermal noise as a potential on the electrodes originating from a

lumped resistor gives the expression [122]

Γheat = e2kBT R

m~ωCOMd2 (7.39)

where we take T = 300 K. Apart from the fact that a realistic value of R gives Γheat ∼ 1s−1

rather than 1ms−1, this model predicts a scaling Γheat ∝ d−2. The data presented in Figures

7.2 and 7.4 are both inconsistent with this scaling.

Among people who have made experimental studies of heating, a popular model of heating assumes a random distribution of fluctuating charges at the electrode surface. The potentials

due to the fluctuations at different points add incoherently, leading to a scaling Γheat ∝ d−4,

which is consistent with the data in Figures 7.2 and 7.4(a). Unfortunately, we have made little

progress in identifying the source of these fluctuations. One possible mechanism involves the

adsorption of gas onto patches of the electrode. The gas charges up, and movement of the

patches gives rise to charge fluctuations. However, it is hard to imagine the motion giving rise

to significant noise power at ωCOM ∼ 5 MHz. Another intriguing possibility is the formation of

charge traps by impurities and crystal defects at the electrode surface [126]. Electrons tunnel

in and out of these traps over a large range of timescales, giving rise to a 1/f noise spectrum.

Noise from this source has been detected in other systems for frequencies up to 100 GHz [127].

The 1/f spectral dependence is also consistent with the observed scaling SE ∝ 1/ωCOM.
149

0

0.5

1

1.5

2

2.5

3

3.5

4

4.5

5

11/94

1/953/955/957/959/95

11/95

1/963/965/967/969/96

date

heating rate [quanta/ms]

Figure 7.5: Heating measurements for trap P1 over a period of two years. The solid line is

a guide to the eye, but indicates a distinct increase in heating rate with trap age. The trap

remained under vacuum for the entire period shown.
150

7.4.2 Sympathetic Cooling for Quantum Logic

Heating is currently the largest source of error in our entangling gate through the DebyeWaller effect. Moreover, the temperature of the ion crystal increases as a computation proceeds,

so that the error contribution increases over the value of 0.06 quoted above. Section 8.2 furnishes a striking example of this effect. Apparently we must actively cool the ions to perform a

long computation. However, the cooling mechanism will cause decoherence if it couples to the

register.

We have proposed a scheme for sympathetic cooling [128] of the quantum register by

laser-cooling only the center ion in a string with an odd number of ions [17]. Here the COM motion of the whole crystal is cooled through coupling to the center ion. If the ions are sufficiently

spatially separated, one can simply focus the cooling beams so that they affect only the center

ion. By using a different species for the cooling ion, the cooling transition can be separated from

any transition of the logic ions by many THz, relaxing the focusing requirement. In either case,

the cooling affects only the internal states of the center ion, leaving all other internal coherences

intact. If the logic operations use a mode in which the center ion remains at rest, the motional

coherences in that mode are also unaffected by the cooling. In this case the logic mode will not

directly couple to the uniform electric fields causing the heating and the decoherence will arise

from the Debye-Waller effect (see Secs. 7.2.3 and 7.3).

If we use different species for logic and cooling ions, the dynamics of the crystal are

no longer governed by the theory of Section 2.1.3, but change according to the mass ratio
151

μ = mcool/mlogic of the two species. The potential energy (Eq. 2.21) remains the same, since

mω2

z depends only on the electric field strength. The equilibrium positions ui (Eq. 2.22) are

then independent of μ. To solve for the normal modes, we define a normalized time T = ωzt

and normalized motional amplitudes Qi(t) as Qi = qi, i 6= nc, Qnc = qnc

√μ. Here nc is the

index of the center ion. The motion is then governed by the Lagrangian [17]

L = 1

2

X

N

i=1

μdQi

dT ¶2

− 1

2

X

N

i,j=1

A0

ijQiQj (7.40)

A0

ij =







Aij i, j 6= nc

Aij/

√μ i or j = nc, i 6= j

Aij/μ i = j = nc

(7.41)

Aij =







1+2PN

k=1,6=i

1

|ui−uk|3 i = j

−2 1

|ui−uk|3 i 6= j

(7.42)

where i and j index the ions and nc is the index of the center ion.

Figure 7.6 shows the axial mode frequencies calculated from the Lagrangian Eq. (7.40)

for 3, 5, 7, and 9 ions. The modes are numbered in order of increasing frequency (at μ = 1),

and are normalized to the trap frequency ωz for a single logic ion. In each case, the lowest-lying

mode has all ions moving in the same direction and is just the COM mode for μ = 1. For μ 6= 1,

however, all the odd-numbered (“hot”) modes participate in the center-of-mass motion of the

crystal. Hence all these modes couple to the spatially uniform heating field. The even-numbered

(“cold”) modes correspond to the (N −1)/2 modes for which the center ion does not move. Their

frequencies are independent of μ, and they do not participate in the COM motion, so they do

not heat.
152

Cold modes are suitable for quantum logic, since they do not heat and they are not

directly affected by the sympathetic cooling. Figure 7.6 shows that each cold mode is nearly

degenerate with a hot mode for μ À 1. To perform our entangling gate, we need to resolve the

logic mode from other normal modes, so small frequency splittings limit the speed of our logic

gates. The regime μ . 1 is more convenient for quantum logic, since mode #2 is well-resolved

from all other modes in this case.

We can predict the mode heating rates Γk for the kth hot mode of the two-species crystal

by modifying Eq. (7.38). Normalizing to the COM heating rate for a one-species crystal with

the same number of ions and the same trap frequency, we find [17]

γk ≡ Γk

Γheat

= 1

Nξk





v(k) nc

√μ

+ X

N

j=1,6=nc

v(k)

j



 (7.43)

assuming that the spectral density SE(ω) is constant over the frequency range of the normal

modes, i.e., SE(ωz) = SE(ζkωz). Figure 7.7 shows the calculated heating rates for the hot modes

of 3, 5, 7, and 9 ions. Since the heating rate is inversely proportional to the mode frequency, the

variation of heating rate in Figure 7.7 is partly due to the variation of mode frequency with μ.

We now consider the effect of species substitution on the transverse dynamics. The

transverse potential (Eq. 2.11) depends on the mass, so the trap potential is different for the

center ion than for the others. If we include the effects of the static field, the transverse potential

becomes

ωx = ωRFr

1 − ωz

ωRF

2

(7.44)
153

    





    





    





    

































 









 

 

μ

ζN

ζN

μ

D E

F G

Figure 7.6: Normalized axial mode frequencies as a function of μ for (a) 3, (b) 5, (c) 7, and (d)

9 ions.
154

    







    









    









    

















 

 

D E

F G













 

















KHDWLQJ UDWH KHDWLQJ UDWH

μ μ

Figure 7.7: Normalized heating rates for the axial modes as a function of μ for (a) 3, (b) 5, (c)

7, and (d) 9 ions.
155

under the condition ωx = ωy. Using the same normalizations of time and amplitude as for

Eq. (7.40), we have the Lagrangian

L = 1

2

X

N

i=1

μdxi

dT ¶2

− 1

2

X

N

i,j=1

B0

ijxixj (7.45)

B0

ij =







Bij i, j 6= nc

Bij/

√μ i or j = nc, i 6= j

Bij/μ i = j = nc

(7.46)

Bij =







ε2 − 1

2 − PN

k=1,6=i

1

|ui−uk|3 i = j, j 6= nc

ε2

μ − 12 − PN

k=1,6=i

1

|ui−uk|3 i = j = nc

1

|ui−uj |3 i 6= j

(7.47)

where ε was defined after Eq. (2.28).

For sufficiently small ε, the ion string breaks into a zigzag pattern as described in Section 2.1.3. We can find the critical value εs(μ) for this transition by solving det B0

(ε, μ) = 0.

Figure 7.8 shows the critical anisotropy as a function of μ for 3, 5, 7, and 9 ions. As discussed in

[17], either of two different modes becomes unstable, depending on the value of μ. The arrows

in Fig. 7.8 show the position of the crossover between the two modes.

Figure 7.9 shows the frequencies of the transverse modes, normalized to ωz, for 3, 5, 7, and

9 ions. We choose ε = 1.1εs(μ); at higher values of ε the mode spacings decrease rapidly. Again

we find cold (even-numbered) and hot (odd-numbered) modes. The large frequency splitting

required to use a cold transverse mode for quantum logic only occurs near the zigzag instability

and for μ ∼ 1, as shown by Figure 7.9. The effect of perturbations under these conditions is
156

μ

    

εV









1 

1 

1 

1 

Figure 7.8: Trap anisotropy at the zigzag instability as a function of μ for 3, 5, 7, and 9 ions.

Arrows indicate the crossover point between two distinct zigzag instabilities.
157

probably large, so the axial modes are better candidates for logic than the transverse modes in

this scheme.

We calculated the heating rates for the transverse modes in the same way as for the axial

modes. Again, all hot modes couple to the COM motion. Figure 7.10 shows the results for 3, 5,

7, and 9 ions.

The analysis above indicates that, all other things being equal, it is best to use a cooling

ion that is not much heavier than the logic ions, and that the lowest-lying cold axial mode is

most suitable for quantum logic. For technical reasons, however, we plan to implement sympathetic cooling of 9Be+ with 24Mg+. For three ions, an axial frequency of 10 MHz then yields a

spacing of 1.6 MHz between the cold axial breathing mode and its nearest neighbor. This splitting is large compared with our sideband Rabi frequencies (∼ 50 kHz), so sympathetic cooling

is compatible with quantum logic even in this nonideal case.
158

    





    





    





    





















μ

ζN

ζN

μ

D E

F G



   





  

  





















 







Figure 7.9: Normalized frequencies of the transverse modes as a function of μ with 2 = 1.120(μ)

for (a) 3, (b) 5, (c) 7, and (d) 9 ions.
159

    









    









    









    















D 



 

 E













 

















F G

KHDWLQJ UDWH KHDWLQJ UDWH

μ μ

Figure 7.10: Normalized heating rates of the transverse modes as a function of μ with 2 =

1.120(μ) for (a) 3, (b) 5, (c) 7, and (d) 9 ions.
Chapter 8

A Decoherence-Free Quantum Memory

8.1 Decoherence-Free Subspaces

Large-scale quantum computers will require robust long-term storage of quantum information [129]. As we saw in Chapter 6, a multitude of decoherence mechanisms stand in the

way of a reliable quantum memory. In a classical computer, one usually evades memory errors

by storing backup copies of information. However, the destructive nature of quantum measurement makes it impossible to reliably copy quantum information [130], so we must resort to

other methods. One approach to improving quantum memory is to try to reduce the sources of

noise on physical grounds, using magnetic shielding or “field-independent” transitions to avoid

magnetic dephasing (Sec. 7.1) or reducing the technical noise of a laser to improve gate fidelity

(Sec. 7.2.4). However, one can also use techniques based on the properties of quantum information to improve a quantum memory. These fall into two broad categories: decoherence-free

subspace (DFS) encoding [131, 132, 133, 134, 135] and quantum error correcting codes (QECC)

[136, 137, 138, 15, 139].

QECC’s are the quantum analog of the error-correcting codes developed in classical information theory. They usually encode one qubit’s worth of information (a “logical qubit”) into

two basis states of a small quantum register (the “physical qubits”). The decoherence mecha-
161

nism is assumed to cause an error on one of the physical qubits, altering the logical qubit as

well. A properly selected partial measurement of the state of the physical qubits specifies the

type of error that has occurred, without revealing any information on the state of the logical

qubit. Applying suitable quantum logic then reverses the effect of the error. One can construct

a wide variety of QECCs [140, 141, 142, 139] and many of these permit fault-tolerant QC, allowing perfect error correction even if the correction process is itself imperfect [143, 144, 139].

QECC-based fault tolerance comes at a high cost: for a large-scale quantum computation, we

need & 10 physical qubits per logical qubit [145].

Collective decoherence processes, which have the same effect on each qubit, are expected

to cause decay of many proposed quantum memories [146, 147, 148, 149, 131, 132, 133]. In fact,

the most prominent decoherence mechanism in our experiment is the collective dephasing caused

by fluctuating magnetic fields, discussed in Section 7.1. Since the fields are spatially uniform

across the ion string, the Hamiltonian for this process is

HB ∝ Bz(t)

X

N

i=1

Si,z = gJz (8.1)

where Bz is the magnetic field component along the quantization axis (not along the trap axis

zˆ!) Since we do not know Bz(t), we cannot predict the evolution of an arbitrary superposition

of register states.

One can exploit the equal coupling of Eq. (8.1) to find subspaces of the full register

Hilbert space that are unaffected by collective decoherence. Since the noise Hamiltonian HB

depends only on Jz, a superposition of two eigenstates of Jz with the same eigenvalue does not

evolve under HB. Modeling the evolution under HB by the transformation |↑i → eiφB |↑i, we

see that an arbitrary superposition of the two-ion states |↓↑i, |↑↓i evolves as
162

a|↓↑i + b|↑↓i → eiφB (a|↓↑i + b|↑↓i) = a|↓↑i + b|↑↓i |a|

2 + |b|2 = 1 (8.2)

Since the overall phase has no physical relevance, the superposition does not evolve under HB.

On the other hand, if the coupling were not the same for the two qubits, an unknown relative

phase would develop between |↓↑i and |↑↓i and a superposition of |↓↑i and |↑↓i would decohere.

The space spanned by |↓↑i and |↑↓i is the simplest example of a decoherence-free subspace [131, 132, 133]. Since this space contains two states, and the amplitudes a, b of the states

are invariant under HB, we can encode one qubit’s worth of information into this DFS. Below

we experimentally demonstrate encoding of a qubit into this DFS and show that the encoded

qubit resists collective dephasing. Although we have discussed only the coupling to magnetic

field so far, encoding into the DFS protects against all forms of collective dephasing, i.e. any

noise Hamiltonian of form (8.1). In particular, our data show the resistance of the DFS to

collective dephasing originating from either a fluctuating magnetic field or a fluctuating uniform

Stark shift.

The DFS discussed above only protects information from collective dephasing. A general

collective process has the Hamiltonian

Hcoll = g(t)

X

N

i=1

nˆ(t) ·

*

Si = g(t)ˆn(t) ·

*

J (8.3)

where both g(t) and ˆn(t) fluctuate. In this case the evolution under Hcoll depends only on the

value of *

J. In particular, a state |ΨSi with *J|ΨSi = 0 is invariant under Hcoll. Such states

are called singlet states, and the space of singlet states is the most general DFS for quantum

registers made of qubits. For two qubits, there is only one such state, namely |↓↑i − |↑↓i. How-
163

ever, the number of singlet states grows rapidly with the number of qubits and approaches the

number of register basis states in the limit of large N [131]. We therefore anticipate little or

no overhead associated with the singlet DFS encoding for large registers. Of course, the singlet

DFS encoding protects against collective dephasing as a special case.

Generic DFSs have been shown to support fault-tolerant QC [134, 135]. Moreover, a

generic DFS is robust to small perturbations that break the collective symmetry; transitions

between the DFS and the rest of the Hilbert space are suppressed to first order [134]. As a

robust, low-overhead, QECC-compatible way to remove the ubiquitous collective decoherence,

DFSs will most likely be intrinsic to future quantum computing architectures. Logic gates on

DFS-encoded qubits have been proposed in the context of cavity quantum electrodynamics [150]

and solid-state quantum logic schemes [135]. Also, a recent experiment has observed resistance

of the singlet state of two photons to collective noise [151], but a DFS of at least two states is

required to encode a qubit. Here we encode a qubit into a DFS for the first time.

8.2 Experimental Implementation of Encoding

Our encoding method reversibly maps an arbitrary state of one ion onto the DFS spanned

by the states |Ψ±i = |↓↑i ± |↑↓i, which is the same as the DFS over |↓↑i, |↑↓i discussed above.

The sequence of operations used to demonstrate encoding is given in Table 8.1. Here a “block”

denotes a group of operations that jointly perform a quantum logic task. The first Ramsey

block prepared the bare state |Ψbare(α)i = |↓i ⊗ (|↓i − ieiα|↑i by rotating one ion while leaving

the other ion alone. We refer to the ion rotated by the first Ramsey block as ion 2. In the

bare state, the quantum information was carried by ion 2 alone. The encoding block mapped

|Ψbare(α)i to the DFS-encoded state |ΨDFS(α)i = |Ψ−i + eiα|Ψ+i, in which the information
164

was carried jointly in the DFS states of the two ions. We allowed the encoded state to evolve

freely for a short time (usually ∼ 20μs), then decoded it by performing the inverse of the unitary

evolution applied by the encoding block. Decoding nominally returned the register to the bare

state |Ψbare(α)i. Finally, the second Ramsey block rotated ion 2 alone about an axis specified

by β.

Ignoring the encoding and decoding blocks for a moment, we see that the two Ramsey

blocks just implemented Ramsey spectroscopy on ion 2. Sweeping β and detecting hSzi gave

rise to the usual Ramsey fringes (Eq. 6.6), but with a constant offset in hSzi, since ion 1 always

remained in the state |↓i. The amplitude of the Ramsey fringes nominally measured the coherence remaining in the state of ion 2 just before the second Ramsey block. Since the encoding

and decoding blocks together amounted to just the identity operator, we could measure the

coherence of the encoded state from the amplitude of the Ramsey fringes.

To perform a Ramsey block, we individually addressed the ions using the RF phase and

trap voltage. In a classical picture of spin, the first rotation (e.g., step 1 in Table 8.1) took ↓↓ to

&&. The second rotation (e.g., step 2) reversed the sense of rotation on ion 1, while keeping it

the same on ion 2, so the second rotation took && to ↓→. This procedure rotated ion 2 alone,

without changing the state of ion 1. The phase α was just set by the RF phase. More generally,

two rotations by θ at these ion phases generated the state cos 2θ|↓i − ieiα sin 2θ|↑i.

In practice, we optimized the Ramsey block by using θ = π/2, which nominally performed

a π-pulse on ion 2, and maximizing the probability P1 to have one ion bright. We used endcap

voltages V1, V2 for the two rotations so as to change φ1 by π. V1 was the same endcap voltage
165

Table 8.1: The sequence of operations used to demonstrate DFS encoding. Rotations by an angle

θ are written Ucar(θ). “Block” refers to a group of operations that jointly perform a quantum

logic task. A Ramsey block performs a π/2 rotation on ion 2 while leaving ion 1 alone. The

encoding and decoding blocks map the information in ion 2 into and out of the DFS spanned

by |Ψ−i, |Ψ+i. We specify some intermediate register states for clarity.

Number Block Operation φ1 φ2 State After Operation

0 Initialize – – – |↓↓i

1 Ramsey Ucar(π/4) α α –

2 Ucar(π/4) α + π α |↓i ⊗ (|↓i − ieiα|↑i)

3 Encode U−1

ent 0 0(|↓↓i + i|↑↑i) − eiα|Ψ+i

4 Ucar(π/2) π/2 0 |Ψ−i + eiα|Ψ+i

5 Wait – – – –

6 Decode Ucar(π/2) 3π/2 π (|↓↓i + i|↑↑i) − eiα|Ψ+i

7 Uent 0 0 |↓i ⊗ (|↓i − ieiα|↑i)

8 Ramsey Ucar(π/4) α + β α + β –

9 Ucar(π/4) α + β + π α + β |↓i ⊗ (sin β

2 |↓i − ieiβ cos β2 |↑i)

10 Detect – – – –
166

used for the entangling operations. We set V2 from the calibration curve obtained for the Bell

inequality experiment (Sec. 6.1). We then varied the differential voltage Vshim between the trap

endcaps so as to produce a uniform field along ˆz at the ion. For some value of this voltage,

ion 2 moved by an integer number of wavelengths of the Raman beam pattern between the first

and second rotations, so φ2 remained the same for the two rotations, giving P1 ∼ 0.98. After

tweaking V2 and Vshim to maximize P1, we reduced the carrier pulse lengths to half their former

values, corresponding to θ ≈ π/4. We then tweaked the pulse lengths slightly to make P1 = P2,

as expected for the state |↓i ⊗ (|↓i + eiα|↑i. The resulting populations were approximately

P0 = P1 = 0.48, P2 = 0.04. Although the carrier pulses were less than 1 μs long, we had to

allow ∼ 5μs settling time after changing the trap voltage; otherwise the ion phases fluctuated

during the second rotation, reducing the fidelity. The total time required to perform the Ramsey

block was about 15 μs.

To encode the bare state into the DFS, we first applied the entangling gate three times

(step 3 of Table 8.1) by tripling the usual entangling pulse duration, nominally obtaining the

evolution U−1

ent (see Eq. (5.6). This operation mapped the part of the bare state in |↑i into the

DFS state |Ψ+i. The phases of the subsequent π/2 pulse (step 4) were set so that this pulse

had no effect on |Ψ+i, but mapped the state |↓↓i+i|↑↑i resulting from the entangling operation

into the orthogonal DFS state |Ψ−i.

We set the phases for step 4 by changing the endcap voltage from V1 to a third value

V3 and using the appropriate RF phase. We determined the value of V3 roughly from the calibration curve. Using a π-pulse on ion 2 to generate |↓↑i, then applying the encoding block

nominally produced the DFS state |Ψ+i, with P1 = 1. An incorrect value of V3 reduced P1,

so we optimized V3 by encoding |↓↑i and maximizing P1. Here P1 was independent of the RF
167

phase, so we were able to set V3 and the RF phase separately. To set the RF phase, we applied

the encoding block to the initial state |↓↓i, nominally mapping it to |Ψ−i, and maximized P1

with respect to the RF phase. Since the two bare basis states |↓↓i, |↓↑i were mapped to the DFS

basis states |Ψ−i, |Ψ+i, the superposition principle ensured that an arbitrary bare state would

be mapped to a DFS state. Applying U −1 ent took about 30 μs. Adding in the time to change the

trap voltage and apply the rotation, we found that the encoding block required about 35 μs.

After waiting a minimum of ∼ 25μs, we decoded the DFS state by applying the operator

inverse of the encoding block. First we applied another π/2 pulse with the rotation axes of both

ions reversed, corresponding to a change in the RF phase by π. The trap voltage was kept at V3

throughout steps 4–6 to save settling time. We then changed the trap voltage to V1 and applied

the usual entangling gate. We had to apply the entangling gate three times to obtain its inverse,

so the error of the inverse gate was three times that of the usual gate. The gate error from the

DW effect increased due to heating, and this error would have tripled if we had used the usual

gate in the encoding and the inverse gate in the decoding. The decoding block took 15 μs to

complete. After decoding, the second Ramsey block and the detection produced the Ramsey

fringes that constituted our signal.

We built up a histogram at each step of β from 1000 shots of the experiment. These data

were analyzed as in the experiment of Chapter 5. We constructed reference histograms from the

data averaged over β and used these references to deduce the value of Pi at each step of β. Ion

2 typically scattered about 10% more photons in |↓i than ion 1 did, presumably due to unequal

illumination by BD. We therefore allowed the mean of the Poissonian for one ion bright to vary

independently of the mean for two ions bright. The extra free parameter did not increase the

error in determining Pi. We fit the data on Pi(β) to sinusoids of amplitude P ̃i. The error in P ̃i
168

was typically 0.05–0.08.

To determine the fidelity of the encoding/decoding sequence, we compared data taken

with the encoding and decoding (“DFS data”) to data taken without these operations (“bare

data”). Dead time was substituted for the encoding/decoding operations in the latter case, so

that the time between Ramsey blocks was the same in both cases. Ideally P ̃0 = 0, but in practice

P0(β) oscillated appreciably in the data. For the bare data we found P ̃0 . 0.04, showing that

the presumed individual rotation of ion 2 actually affected ion 1 as well. The two rotations

composing a Ramsey block were therefore identical up to a gate error of ∼ 0.02. For the DFS

data we found P ̃0 . 0.10. P1 and P2 oscillated π out of phase with P0 in both cases and we

found P ̃2 = P ̃0 + P ̃1 within error. We characterized each dataset by the coherence C = P ̃2. For

the bare data we found Cbare = 0.69±0.08, while for the DFS data we found CDFS = 0.43±0.05.

The loss of contrast of the bare data mostly comes from the decoherence of the bare state over

the ∼ 100μs delay between the two Ramsey blocks, as well as from the gate error, which we

estimate as ∼ 0.04 for each Ramsey block. We expect the population decay to be smaller than

the decay of the coherence, so we can set a lower limit for the encoding/decoding fidelity as

Fencode & CDFS/Cbare = 0.62. The errors in encoding and decoding mostly come from the DW

effect on the entangling operations.

We set the synthesizer frequencies used for the Ramsey blocks to the exact resonance

frequency of the carrier transition, which we measured using a normal Ramsey experiment on

both ions. Here TR (∼ 100μs) was much longer than the total time required to drive the Ramsey pulses (∼ 1μs), so the measured value of ω↓↑ did not suffer a significant Stark shift from

the Raman beams. On the other hand, we measured the RR double-pass frequency ωent used

for the entangling transition (step 3) by taking a spectrum for a fixed pulse duration, so the
169

measured frequency experienced a relatively large Stark shift of ∼ 7 kHz. We set the synthesizer

frequency of the π/2 pulse (step 4) to ωent − ωEOM/2 (in Hz) with accuracy better than ±1 Hz,

just as described in Section 5.2, to ensure mutual coherence of steps 3 and 4. We used the same

frequencies for steps 3 and 7, and the same frequencies for steps 4 and 6.

The operations in the Ramsey blocks (steps 1, 2, 8, and 9) were thus all mutually coherent, as were the encoding and decoding sequences (steps 3, 4, 6, and 7). However, the

encoding/decoding blocks were not coherent with the Ramsey blocks, since the synthesizer frequencies between the two blocks differed by the Stark shift (∼ 7 kHz). In a frame rotating at

the beat frequency of step 4, the bare state phase α precesses at ∼ 7 kHz, so the data were

effectively averaged over the bare state phase. The persistence of the coherence under these

conditions shows that the encoding procedure works even if we do not know the bare state

phase. It is then no stretch of credulity to suppose that the encoding procedure works for an

arbitrary superposition of |↓i and |↑i.

8.3 Decoherence of Bare and DFS States

To study the effects of decoherence on the bare and DFS states, we applied an engineered

noisy environment for a time tnoise during step 5. We left some dead time after tnoise so that

step 5 always lasted a fixed time of 25μs. To apply the noise, we turned on B2 with a randomly

varying intensity for the time tnoise. As described in Section 4.3.2, we directed B2 into the trap

with a pickoff mirror so that we could vary the B2 and R2 polarizations independently. While

the R2 polarization was set to minimize the Stark shift from R2, the B2 polarization was set

to maximize the B2 Stark shift. We measured the Stark shift of B2 by turning on B2 for a

short time between the two pulses of a Ramsey experiment and observing the resulting fringe
170

shift. Nominally, B2 equally illuminated the ions, so they experienced the same Stark shift. The

fluctuating intensity of B2 therefore introduces a pure collective dephasing of the ion string.

We applied the random intensity modulation to B2 by modulating the amplitude of the

RF power driving the B2 switch AOM. The modulation signal originated from the white-noise

output of a Stanford Instruments DS345 signal generator. The noise was low-pass filtered with

cutoff frequency 100 kHz and attenuation of 6 dB/octave using a Stanford Instruments SR560

preamplifier. We used the resulting signal to induce amplitude modulation of the RF signal

driving the B2 switch before the RF signal was sent to its amplifier. We characterized the resulting intensity noise of B2 using a photodiode of bandwidth > 500 kHz. Figure 8.1 shows the

power spectrum of a typical photodiode signal. We see that the spectrum is fairly flat below 100

kHz and drops off rapidly for higher frequencies. Since the photodiode bandwidth was much

larger than the noise bandwidth, we can completely characterize the noise from the photodiode

signal.

While the bare state should rapidly decohere under the noisy environment, the DFS state

should resist the decoherence. Figure 8.2 shows the coherence remaining in the bare and DFS

states as a function of the noise duration tnoise. The coherence data are normalized to their

values for zero applied noise. We extracted the decay rates for the bare and DFS states from

the exponential fit lines in Fig. 8.2, obtaining a rate of 0.18(1) μs−1 for the bare state and

3.5(5.0) × 10−3 μs−1 for the DFS state. This result confirms our expectations; clearly collective

dephasing occured, since the test state decayed, but the DFS state resisted the dephasing.

For white-noise intensity fluctuations of the Stark-shifting beam, we expect C to decay
171

2

3

4

5

6

10

2

3

4

5

6

103

2 4 6 8

104

2 4 6 8

105

2

Frequency [Hz]

Noise Power [arb.]

Figure 8.1: Power spectral density of the B2 intensity as determined from the photodiode signal.

The spectrum is flat to ∼ 100 kHz and drops off rapidly for higher frequencies, as desired. The

measurement time of 5 ms limits the spectral resolution.

0.1

2

3

4

5

6

1

 Normalized Coherence

0 5 10 15 20

 Noise Duration [μs]

Figure 8.2: Decay of the DFS state (circles) and the bare state (crosses) under engineered

collective dephasing. The noise is applied for a fraction of the delay time of about 25 μs between

encoding and decoding. Coherence data are normalized to their values for zero applied noise.

The fit lines are exponential decay curves, giving decay rates of 0.18(1) μs−1 for the bare state

and 3.5(5.0)×10−3μs−1 for the DFS state. The bare data point for zero applied noise is excluded

from the fit (see text).
172

exponentially for the bare state, as shown by the fit line. Since the intensity fluctuations of the

noise beam had ∼ 100 kHz bandwidth, Eq. (7.5) predicts a crossover between Gaussian and

exponential decay of the bare state for tnoise of a few μs. This causes the anomalously low decay

rate between tnoise = 0 and 2.5μs in Figure 8.2. We were able to fit all the bare data in Figure 8.2 by numerically evaluating Eq. (7.4) using the observed noise spectrum from Figure 8.1.

However, we also obtained a good fit to a simple exponential by simply excluding the bare data

point with tnoise = 0. The decay rate from this simpler fit is that quoted above.

To investigate the eventual decay of the DFS state, we increased the delay time to about

200 μs and applied decoherence for up to 100 μs during the delay time. The results are shown

in Figure 8.3. The line is the best fit of the data to a Gaussian decay curve. The DFS state

coherence dropped by 50% for 100 μs applied noise relative to its value for the same delay time

and no applied noise. We believe that this decay was due to small departures from equal illumination by B2. In this case B2 applies a small, random differential phase to the two ions, modeled

by the transformation |↑i → eiφdiff |↑i on ion 2 alone. Taking φdiff constant for a moment and

averaging over α, we find C = e−φ2

diff/2, so the contrast drops even if the differential phase shift

is not noisy. The phase φdiff should increase linearly with tnoise, so this model reproduces the

approximate Gaussian decay observed in Figure 8.3.

We also measured the decay of the bare and DFS states under ambient conditions in our

laboratory. Here we left a variable delay time (step 5) between encoding and decoding to give

the ambient noise a variable time to act. Figure 8.4 shows the coherence as a function of the

delay time. The coherence data for this case are normalized in the same way as the data with

applied noise. We empirically found a roughly exponential decay, shown by the fit lines, for both

bare and DFS states. The fit lines give a decay rate of 7.9(1.5) × 10−3 μs−1 for the test state
173

0.30

0.28

0.26

0.24

0.22

0.20

0.18

0.16

0 20 40 60 80 100

Noise Duration [μs]

Coherence

Figure 8.3: Decay of the DFS state (crosses) under application of B2 for up to 100 μs. The noise

is applied for a fraction of the delay time of 200 μs between encoding and decoding. Coherence

data are not normalized. The line is the best fit to a Gaussian decay curve.
174

and 2.2(0.3) × 10−3 μs−1 for the DFS state. The decoherence of the bare state was dominated

by ambient fluctuating magnetic fields with most of their power spectrum at 60 Hz and its harmonics. We treated this situation in detail in Section 7.1. Since the fields were roughly uniform

across the ion string, they induced collective dephasing. The DFS state is resistant to this type

of decoherence, as demonstrated by Figure 8.2. Since the DFS state maintained coherence much

longer than the test state, we conclude that collective dephasing from magnetic field noise was

the major ambient source of decoherence for the bare state. Though Figure 8.4 presents only

normalized contrasts, the unnormalized contrast of the DFS state is higher than that of the test

state for delay times exceeding 150 μs. Hence, encoding into the DFS materially improves the

information storage under real conditions.

The apparent decay of the DFS coherence arises from the heating of the ions. As the

delay time increases, so does nCOM, and thus the DW gate error in the decoding block increases

as well. Unfortunately, we do not have good theoretical predictions for the DW gate error when

the DW error becomes large. We found a heating rate of ∼ 20ms−1 from data taken at roughly

the same time as the data for this experiment. The error predicted by Eq. (7.33) for this heating

rate is considerably larger than that observed in Figure 8.4. In any event, Figure 8.2 conclusively demonstrates that the loss of coherence is not due to any effect of collective dephasing.
175

2

3

4

5

6

7

8

9

1

 Normalized Coherence

100 200 300 400 500

 Delay Time [μs]

Figure 8.4: Decay of DFS-encoded state (circles) and test state (crosses) under ambient decoherence. We vary the delay time between encoding and decoding to give the ambient noise a variable

time to act. Coherence data are normalized to their values for zero applied noise. The fit lines

are exponential decay curves for purposes of comparison and are not theoretical predictions.

The decay rate from the fit is 7.9(1.5) × 10−3 μs−1, for the test state and 2.2(0.3) × 10−3 μs−1

for the DFS state. Since the coherence time of the DFS-encoded state is much longer than that

of the test state, we see that the chief source of ambient decoherence is collective dephasing.
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